Solid State Theory: An Introduction

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Solid State Theory: An Introduction

U. Rössler

Solid State Theory: An Introduction Second revised and extended edition With 123 Figures and 6 Tables

Prof. Dr. h.c. Ulrich Rössler Universität Regensburg Inst. Theoretische Physik Universitätsstr. 31 93040 Regensburg Germany

ISBN 978-3-540-92761-7 e-ISBN 978-3-540-92762-4 DOI 10.1007/978-3-540-92762-4 Springer Dordrecht Heidelberg London New York Library of Congress Control Number : 2009929708 c Springer-Verlag Berlin Heidelberg 2009 This work is subject to copyright. All rights are reserved, whether the whole or part of the material is concerned, specifically the rights of translation, reprinting, reuse of illustrations, recitation, broadcasting, reproduction on microfilm or in any other way, and storage in data banks. Duplication of this publication or parts thereof is permitted only under the provisions of the German Copyright Law of September 9, 1965, in its current version, and permission for use must always be obtained from Springer. Violations are liable to prosecution under the German Copyright Law. The use of general descriptive names, registered names, trademarks, etc. in this publication does not imply, even in the absence of a specific statement, that such names are exempt from the relevant protective laws and regulations and therefore free for general use. Cover design: eStudio Calamar S.L. Printed on acid-free paper Physica-Verlag is a brand of Springer-Verlag Berlin Heidelberg Springer is part of Springer Science+Business Media (www.springer.com)

To Erika

Preface to Second Edition

Given the fact that the first edition of this book is almost sold out, the editor has asked for a second edition. I readily accepted this opportunity, because – as many users and I myself had found out in the meantime – there were quite a lot of typing and other errors in the first edition which had to be eliminated. But more important, I saw the chance to extend the book by a new chapter on light–matter interaction, which complements the canon of the book with respect to both the physics of this interaction and the theoretical methods used for its description. Moreover, during the four years since the first edition the whole field has developed as can be seen in the list of references which now contains about 30 new titles. Part of the contents of this book has been presented during a guest lecture at the National Mechnikov University in Odessa in the spring term of 2007. I would like to acknowledge the hospitality of the Department of Theoretical Physics and its head, Prof. V.M. Adamyan, and the contacts with the students who attended this one as their first lecture in English. Finally, I appreciate the support by and easy cooperation with the publisher.

Regensburg, June 2009

Ulrich R¨ ossler

Preface to First Edition

The history of my involvement in this book project starts more than 30 years ago. During the years 1969–1972, my thesis advisor, Otfried Madelung, wrote a series of three textbooks on Solid State Theory, entitled Festk¨ orpertheorie I–III, which appeared in the Springer paperback series Heidelberger Taschenb¨ ucher. My fellow graduate students and friends Manfred Lietz, Rolf Sandrock, Joachim Treusch, and I, were the first to proof-read these books. Better still, we were given the unique opportunity to provide input based on insights gained during our studies. In 1978, when Festk¨ orpertheorie I–III were partly rewritten and translated into English, Otfried Madelung again asked his former disciples, then already established in university positions, for comments and contributions based on their respective research and teaching experience in this field. The result, entitled Introduction to Solid-State Theory, became a widely used textbook, published in several further editions over the following years. Like most textbooks on Solid-State Theory currently used in university physics courses all over the world, Introduction to Solid-State Theory has meanwhile become somewhat outdated. Solid State Physics has evolved significantly and many topics, which 30 years ago were still the subject of active research or even beyond its leading edge, have now become part of our standard knowledge. The idea of accounting for this development in a textbook has been lingering in the collective mind of the Solid State Physics community for quite a while, but it took an initiative by Springer to concretize the project. When Springer editors asked Otfried Madelung to rework on Introduction to Solid-State Theory accordingly, he convinced them that it would make more sense to write a completely new book, proposing me as a potential author. This is how I got involved. Due to the formative influence of Otfried Madelung and his approach to science, my research in Solid State Theory has from the very beginning been oriented towards experimental work, often directly stimulated by concrete experimental results. This tendency was solidified during a year as postdoctoral researcher with Manuel Cardona at Brown University in Providence,

X

Preface

RI, USA. Quite commonly, my research projects were initiated by discussions with researchers renowned for their experimental work and have frequently been conducted in fruitful cooperations. It is to this continuous contact with the physics reality that I owe the down-to-earth approach which characterizes my research and teaching and which should also be noticeable in this book. At the University of Regensburg, where I became professor in 1972, Solid State Physics has been a strong research field, both in experiment and theory. Over the years the topics evolved from magnetism, phase transitions, lattice dynamics, and electronic structure of bulk material to the upcoming fields of high-Tc superconductors, correlated electron systems, surface physics, quantum wells, nano-structures, and composite materials. Quite naturally, Solid State Theory has been a standard part of the physics curriculum in Regensburg. It started as a one-semester course with four weekly lectures in the fourth year of the German diploma curriculum (corresponding to the first year of the graduate education in the Anglo-American system). Soon it was supplemented by a second course on special topics, with the purpose of guiding the students into active research fields. For more than 30 years, I taught these courses on a regular basis, taking turns with my colleagues Joachim Keller, Uwe Krey, Ulrich Schr¨ oder, and Dieter Strauch. The exchange of teaching concepts and problems with these colleagues, and also our joint research projects considerably enriched my lectures. During the last decade I benefitted much from the expertise of my senior coworker Michael Suhrke. Further, important input came from discussions with many colleagues from all over the world during conferences, visits, and sabbaticals in diﬀerent places. My lecture notes for these courses, accumulated and continuously modified over the years, constitute the backbone of this book. Clearly, the book follows a well-defined tradition. Target readers are those students in physics or material science who are interested in understanding the theoretical approach to Solid State Physics, while maintaining contact to the experimental facts. The contents are essentially comparable to those of other textbooks on the same subject, but emphasis is put on new aspects of the field that have resulted from more recent research. Extensive references to related literature in the form of textbooks, topical series, data collections, and selected original papers are provided to establish the connection with the sources of this subject and with active research fields. Each chapter contains a selection of problems and solutions, which are meant to help the reader gain practice with the concepts and the physics explained in the text. Since the number of pages is restricted, this book cannot claim completeness. Nevertheless, wherever possible, reference is given to those important topics that could be covered here only briefly or not at all. In short, this book is intended as an introduction to Solid State Theory, given from the perspective of more than 30 years of learning, teaching, and research in this field. It is a pleasure to thank all those who contributed in one way or the other to this project. I have already mentioned some friends and colleagues and would like to extend my acknowledgments to the students who attended my courses

Preface

XI

and enriched them by their constructive feedback. This applies especially to my diploma and Ph.D. students, who contributed ideas during many hours of discussion about their research projects. A highly visible contribution to this book came from Ingeburg Zirkl who prepared all the figures. A critical reading of parts of the manuscript by my friends and colleagues Joachim Keller and Dieter Strauch, and by my son Thomas has led to considerable improvements of the contents and the text. Finally, I express my gratitude to Springer, in particular to Dr. Claus Ascheron and Dr. Angela Lahee, not only for their expert help and advice but also for their patience in waiting for the final version of this book.

Regensburg, April 2004

Ulrich R¨ ossler

Contents

List of Symbols . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . XVII 1

Introduction . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.1 Aims and Outline . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.2 The Structure of Solids . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 1.3 Pair-Distribution Function and Structure Factor . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

1 2 4 9 12

2

The Solid as a Many-Particle Problem . . . . . . . . . . . . . . . . . . . 2.1 The Hamiltonian of the Solid . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.2 Separating the Motion of Electrons and Ions . . . . . . . . . . . . . . 2.3 Thermal Expectation Values . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.4 Theory of Linear Response . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2.5 Kubo’s Formulas: Response Functions . . . . . . . . . . . . . . . . . . . . 2.6 Properties of Response Functions . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

15 16 20 22 24 28 32 36

3

Lattice Dynamics: Phonons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.1 Harmonic Approximation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2 Normal Coordinates . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.3 Phonons and Occupation Number Representation . . . . . . . . . 3.4 Acoustic Phonons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.5 Optical Phonons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.6 Examples: Phonon Dispersion Curves . . . . . . . . . . . . . . . . . . . . 3.7 The Crystal Lattice at Finite Temperature . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

37 39 42 44 48 55 60 68 73

4

The Free Electron Gas . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.1 Free Electrons Without Interaction . . . . . . . . . . . . . . . . . . . . . . 4.2 Free Electrons in a Magnetic Field . . . . . . . . . . . . . . . . . . . . . . . 4.3 Occupation Number Representation for Electrons . . . . . . . . . .

75 76 84 93

XIV

Contents

4.4 Hartree–Fock Approximation . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.5 The Dielectric Function . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 4.6 Discussion of the Dielectric Function . . . . . . . . . . . . . . . . . . . . . 4.7 Electronic Correlation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

97 102 107 110 115

5

Electrons in a Periodic Potential . . . . . . . . . . . . . . . . . . . . . . . . . 5.1 Density Functional Theory . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.2 Bloch Electrons and Band Structure . . . . . . . . . . . . . . . . . . . . . 5.3 Almost Free Electrons and Pseudo-Potentials . . . . . . . . . . . . . 5.4 LCAO and Tight-Binding Approximation . . . . . . . . . . . . . . . . . 5.5 Eﬀective-Mass Approximation . . . . . . . . . . . . . . . . . . . . . . . . . . . 5.6 Subbands in Semiconductor Quantum Structures . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

119 120 127 129 139 147 152 157

6

Spin Waves: Magnons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.1 Preliminaries . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.2 The Heisenberg Hamiltonian . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.3 Spin Waves in Ferromagnets . . . . . . . . . . . . . . . . . . . . . . . . . . . . 6.4 Spin Waves in Anti-Ferromagnets . . . . . . . . . . . . . . . . . . . . . . . . 6.5 Molecular Field Approximation . . . . . . . . . . . . . . . . . . . . . . . . . 6.6 Itinerant Electron Magnetism . . . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

161 162 166 170 175 179 185 191

7

Correlated Electrons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.1 Retarded Green Function for Electrons . . . . . . . . . . . . . . . . . . . 7.2 The Hubbard Model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.3 Fermi Liquids . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.4 Luttinger Liquids . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.5 Heavy Fermion Systems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.6 Fractional Quantum Hall States . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

195 196 201 208 213 217 222 228

8

Electron–Phonon Interaction . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.1 Preliminaries . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.2 Coupling Mechanisms . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.3 Scattering Processes: Lifetime, Relaxation . . . . . . . . . . . . . . . . 8.4 The Fr¨ ohlich Polaron . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 8.5 Eﬀective Electron–Electron Interaction . . . . . . . . . . . . . . . . . . . 8.6 Cooper Pairs and the Gap . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

231 232 235 240 247 250 255 263

Contents

9

XV

Defects, Disorder, and Localization . . . . . . . . . . . . . . . . . . . . . . 9.1 Point Defects . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 9.2 Disorder . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 9.3 Approximations for Impurity Scattering . . . . . . . . . . . . . . . . . . 9.4 Electric Conductivity . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 9.5 Metal–Insulator Transition . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

265 266 273 276 280 285 290

10 Light–Matter Interaction . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.1 Preliminaries . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.2 Single-Particle Approximation . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.3 Excitons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.4 Polaritons . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.5 Light-Scattering . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.6 Coherent Interband Dynamics . . . . . . . . . . . . . . . . . . . . . . . . . . . 10.7 Semiconductor Bloch Equations . . . . . . . . . . . . . . . . . . . . . . . . . Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

291 291 294 298 307 312 316 320 323

A

Appendices . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.1 Elements of Group Theory . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.2 Fourier Series and Fourier Transforms . . . . . . . . . . . . . . . . . . . . A.3 Fermi and Bose Integrals . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.4 Sommerfeld Expansion . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . A.5 Calculation of the Exchange Energy . . . . . . . . . . . . . . . . . . . . . A.6 Operators in Fock Representation . . . . . . . . . . . . . . . . . . . . . . .

325 325 329 331 333 335 336

References . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

337

Solutions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

349

Index . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

389

List of Symbols

a a†s (q), as (q) a†i , ai ai , i = 1, 2, 3 Aˆ A bj , j = 1, 2, 3 b†k , bk

Lattice constant Creation, annihilation operator of a phonon with quantum numbers s, q Creation, annihilation operator of spin excitations at site i Primitive lattice vectors Observable Vector potential Primitive reciprocal lattice vectors Creation, annihilation operator of ferromagnetic magnons

B ˆ B B0 BS (y) † , BνQ BνQ

Magnetic induction Observable Bulk modulus Brillouin function Creation, annihilation operator of excitons

c cV cijkl (cIJ )

Cnn (r, r′ )

Velocity of light in vacuum Specific heat at constant volume Components of the compliance tensor, elastic moduli (Voigt notation) Creation, annihilation operator of Bloch electrons Density–density correlation function

d, dcv D, D1,2,3 Dτ i,τ ′ j (q)

Electric dipole vector Deformation potentials Element of the dynamical matrix

c†kσ , ckσ

XVIII List of Symbols

D(ω), D(E) D, D(q, ω)

Density of states Displacement field

e eijk es (q) E

EνQ E, E(q, ω)

Elementary charge Piezoelectric tensor Normalized phonon eigenvector Energy Young’s modulus Energy of Bloch electron Energy of ground state Fermi energy Thermal energy Energy eigenvalue of the many-body electron problem Energy of exciton Electric field

f0 fk F

Oscillator strength Fermi distribution function Free energy

g(g ∗ ) g(r) G GAB , G(α; t, t′ ) Gm , G

Land´e (eﬀective) g factor Pair–distribution function Rigidity modulus Retarded Green function Reciprocal lattice vector

H

Hjell Hso Hspin H

Hamiltonian of phonons or of electrons symmetry point of the bcc Brillouin zone, endpoint of the ∆-axis Many-body Hamiltonian of the solid Parts of H describing electrons, ions, and the electron–ion interaction Hamiltonian of the radiation field and of the light–matter interaction Hamiltonian of the jellium model Hamiltonian of spin–orbit coupling Heisenberg’s spin Hamiltonian Magnetic field

j J, JRR′ , Jij

Electric current density Exchange integrals

kB kF

Boltzmann constant Radius of Fermi sphere

En (k), ǫnkσ E0 EF E(T ) Eel

H, H0 Hel , Hion , Hel−ion Hrad , Hel−rad

List of Symbols

XIX

kFT k

Thomas–Fermi screening parameter Wave vector (of electron)

L

Symmetry point of the fcc Brillouin zone Lagragian of lattice vibrations

m m∗ m Mi , Mτ M E, M M

Free electron mass Eﬀective mass of an electron Magnetic dipole moment Mass of an ion Electric dipole moment Magnetization or magnetic dipole density

n(r) n ˆ nq ns (q, T ) N, N0

Particle (electron) density Density operator Density fluctuation Bose–Einstein distribution function Number of electrons Number of unit cells in the normalization volume Number fluctuation Number of electrons for diﬀerent spin orientation

Nq N± , N↑↓

P nτ Ps (q)

Momentum of an electron Dielectric polarization or electric dipole density Momentum of an ion Conjugate momentum to Qs (q)

q Qi , Qext Qs (q)

Wave vector (of phonon) (Point) charge Normal coordinate of lattice vibration

rs

Density parameter (radius of Wigner–Seitz sphere) Position vector of an electron Lattice vector, equilibrium position of an ion in a Bravais lattice Actual position of an ion in a Bravais lattice Actual position of an ion in a lattice with basis

p, pl P

r, r l R0n Rn Rnτ

sijkl S(q)

Components of the elastic or stiﬀness tensor Static structure factor

XX

List of Symbols

S(q, ω) SR, Si Si, S±

Dynamic structure factor Vector operator for spin at R or at site i Components of vector spin operator

t, tij , tRR′ T Tc (TC ) TN TR0n T1 , T2

Transfer integral Absolute temperature Critical (Curie) temperature N´eel temperature Translation operator Longitudinal, transverse relaxation time

unτ (t) U

Ion displacement from equilibrium position Exchange interaction, correlation energy (Hubbard) Bloch vector Adiabatic potential

U (k, t) Uα ({Rn }) vq vF vL , vT V, Vc VWSC Veﬀ (r) Vxc (r) Vext Vs (q)

Fourier transform of (Coulomb) potential Fermi velocity Longitudinal, transverse sound velocity Crystal or normalization volume Volume of Wigner–Seitz cell Eﬀective single particle potential Exchange-correlation potential External perturbation Matrix element of electron–phonon interaction

Z, ZG

Partition function of canonical and grand-canonical ensemble

α α†k , αk ; βk† , βk

Thermal expansion coeﬃcient Creation, annihilation operators of anti-ferromagnetic magnons Froehlich couling constant Gr¨ uneisen parameter, Sommerfeld coeﬃcient center of the Brillouin zone, k = (0, 0, 0) damping parameter Symmetry line in the fcc (bcc) Brillouin zone connecting Γ and X(H) Gap parameter in superconductors Response function, susceptibility Dielectric, magnetic susceptibility Tensor components of χE , χM Spin susceptibility

αF γ Γ ∆

χAB (ω) χE , χM M χE ij , χij M χ+− (q, ω)

List of Symbols

ǫjl (εJ ) ǫ0 ǫc (ǫxc ) ε0 ε(0) ε∞ ε(q, ω) ε 1 , ε2 η λ Λ μ

μB μ0 ν ωc ωD ωs (q) ωL(O) , ωT(O) ωk ωp ωR Ω ψ(r) ψnk (r) ˜ Ψ, Ψ φν (r − R0n ) φν (r) Φ Φ0 ρ, ρˆ ρ0 ρM ρAB ρ(r) ρ

XXI

Components of the strain tensor (Voigt notation) Average ground state energy per electron Average correlation (exchange–correlation) energy per electron Vacuum dielectric constant Static or low-frequency dielectric constant High-frequency dielectric constant Dielectric function Real, imaginary part of dielectric function (Transverse) Eﬀective charge Localization length Symmetry line in the Brillouin zone connecting Γ and L Reduced mass chemical potential mobility Bohr’s magneton Magnetic field constant Filling factor of Landau levels Cyclotron frequency Debye frequency Frequency of phonon with quantum numbers s, q longitudinal, transverse (optical) phonon frequency at q = 0 Magnon frequency Plasma frequency Rabi frequency Grand-canonical potential Single-electron wave function Bloch function Many-electron wave function (stationary, time-dependent) Atomic orbital centered at R0n Exciton envelope function Force constant Magnetic flux Elementary magnetic flux quantum Statistical operator Statistical operator of thermal equilibrium Mass density Spectral function Charge density Resistivity tensor

XXII

List of Symbols

ρk (t) ρcc (k, t), ρvv (k, t), ρcv (k, t) σik (σI ) σ Σ Σ(α, E), Σ(k, σ) ΘD τk τtr τ ζ ζn (z)

Time-dependent density matrix Components of the density matrix Components of the stress tensor (Voigt notation) Conductivity tensor Symmetry line in the Brillouin zone connecting Γ and K Electron self-energy Debye temperature Single-particle lifetime Transport relaxation time Ion position relative to the lattice vector in a lattice with basis Degree of spin polarization Subband envelope function

1 Introduction

The advent of quantum mechanics in the early twentieth century has fundamentally improved our understanding of the physics of matter in general and of the solid state in particular. Consisting of a very large number of atoms, solids exhibit a rich variety of material properties, whose understanding represents a challenge to the curious scientist. These properties are at the same time a rich source for technical applications. Consequently, in the course of the last century our increasing knowledge about the relationship between the chemical composition and the structure of solids on one side and their particular properties – according to which we identify metals, semiconductors, superconductors, and magnetic materials – on the other side, has led to the invention of an enormous variety of solid state devices. Whole industries have been created based on products that make use of solid state properties. Transistors, sensors, solid state lasers, light-emitting diodes (LED), superconducting quantum interference devices (SQUID), dynamic and magnetic random access memories (DRAM and MRAM) have become essential parts of electronic appliances such as computers, mobile phones, compact disc (CD) and digital video disc (DVD) players, which have revolutionized our daily life and continue to do so. The impressive development in the technology of data storage and handling, symbolized by Moore’s law, demonstrates an unprecedented technological progress, which, though driven by the market, would be unthinkable without the ingenious investigations of generations of physicists. In the past decades, solid state research has uncovered interconnections between the structural, elastic, electronic, magnetic, and optical properties of solids and led to unpredicted fundamental discoveries like heavy fermions, the quantum Hall eﬀects, high-Tc superconductors, the giant magnetoresistance eﬀect, and solid state lasers. The progress in material growth and manipulation, which enables the tailoring of solid state properties by properly choosing the chemicals and the structure, together with the invention of pioneering experimental techniques, has paved the road to the world of low-dimensional systems and nanostructures with new physical and technological perspectives. Quantum dots, single electron transistors, and Carbon nanotubes are the U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 1,

2

1 Introduction

buzzwords in this research field, which has attracted the attention of many solid state physicists. Most recently, the electron spin has moved into the focus of interest with the perspective of applications in spintronics and quantum computation. Accompanying and enabling these developments, powerful theoretical concepts, complemented by numerical tools and computational physics, have been developed for qualitative and quantitative modeling of the solid state and its properties. Thus, solid state physics presents itself as a dynamic and rich research field whose results and progress are well documented not only in the physics journals but also in series publications and data collections devoted to this subject [1]. Recently even an Encyclopedic Dictionary of Condensed Matter Physics appeared on the market [2]. A thorough understanding of the theoretical foundations of the field and how the properties of solids derive from the chemical composition and structure in a quantum mechanical description are an indispensable part of a university physics curriculum.

1.1 Aims and Outline As stated in the preface, this book follows the tradition of a series of textbooks on Solid State Theory that have served generations of graduate students in physics [3–14]; but there are also a number of more recently published books in the field [15–28]. Courses on Solid State Theory, which are found in physics curricula all over the world, are based on theoretical concepts developed in quantum mechanics and statistical physics, but they also require some basic knowledge about solid state phenomena from an introductory course in Solid State Physics for which [12, 29–31] are standard references. The aim of this book is to provide the methods required to describe a many-particle system with about 1023 atoms per cm3 and its material specific properties. This can only be done in an approximate way and for special quantum mechanical states of the system. We focus here on the ground state, which defines the structure of the solid, and on low-lying excited states, which determine the response of the solid to a small external perturbation. This response characterizes material properties like dielectric or magnetic susceptibility and electrical conductivity, whose meaning is already known from more elementary courses. They establish the connection between theoretical description and the real world, explored by experiments and relevant for applications. Therefore, in the individual chapters, the relation to material properties, be it elastic, electronic, optical, or magnetic, will be emphasized in view also of recent results in the field. The aims of the book become more transparent when one looks at the outline. In this introduction (Chap. 1), we survey the diﬀerent forms of condensed matter, which can be distinguished from each other by their pair-distribution function or its Fourier transform, the structure factor. The crystalline solids,

1.1 Aims and Outline

3

which are the focus of this book, represent a specific form of condensed matter with a pronounced structure due to a characteristic long-range order. At the beginning of Chap. 2, the Hamiltonian of a solid composed of ions (nuclei and closed electron shells) and valence electrons will be introduced. These two kinds of constituents, with masses diﬀering by orders of magnitude, can be treated separately as independent subsystems after applying the adiabatic or Born–Oppenheimer approximation. Before doing so, the linear response theory is introduced as the basic concept to describe the material properties of a solid. In Chap. 3, the dynamics of the ions, the heavy constituents of a solid, will be described as theory of lattice vibrations. This will first be done in a classical approach using the model of massive spheres connected by springs; but in a second step, we turn to the quantum-mechanical concept of phonons as elementary excitations of the lattice. Acoustic phonons will be discussed in the context of heat capacity, elastic properties, and sound propagation; optical phonons will be related to optical properties of solids in the far-infrared spectral range. Examples of phonon dispersion curves for quite diﬀerent solids will be presented to illustrate the influence of structure and chemical composition. The next chapters (Chaps. 4–7) are devoted to electrons and their properties. The basic concept of the Fermi surface and the fundamentals of the many-particle theory, like Fock representation, Hartree–Fock approximation, dielectric screening, and electronic correlation, for free electrons in the jellium model (Chap. 4) will be introduced. The influence of the periodic lattice structure on the electron states will be treated in the singleparticle approximation justified by the density–functional theory (Chap. 5). In this chapter, we also present methods for calculating the band structure, which are important to understand the material specific aspects of energy bands, and discuss the properties of two-dimensional electron systems. As a particular outcome of electron–electron interaction, the Heisenberg Hamiltonian will be the starting point in Chap. 6 to discuss spin waves as excitations out of a ground state with ferromagnetic or anti-ferromagnetic ordering. This Hamiltonian will also be used to demonstrate the molecular field approximation and the ferromagnetic phase transition. Finally, the theory of itinerant electron magnetism will be presented in this chapter. Electron–electron interaction is the focus also in Chap. 7 which is devoted to correlated electrons. For the treatment of some aspects in this field, we take advantage of using Green functions, which have to be introduced for this purpose. They will be used to deal with the Hubbard model, leading to the Mott–Hubbard metal-insulator transition. We discuss the phenomenological concept of Fermi liquids and its modification for one-dimensional electron systems. Finally, heavy fermions and the fractional quantum Hall states, both dominated by correlation, are also introduced. In Chap. 8, we go beyond the adiabatic approximation and study the electron–phonon interaction as a prototype of coupling between fermions and bosons. It is relevant in scattering processes, which are essential for the electric conductivity, for relaxation and lifetime eﬀects of free carriers, but can also mediate an attractive electron–electron interaction that gives rise

4

1 Introduction

to the formation of Cooper pairs, a basic concept of superconductivity. While the ideal crystalline order of the ions has been assumed in all these chapters, Chap. 9 will be devoted to disorder caused by impurities and other deviations from lattice periodicity, and their consequences for electron states and electric conductivity. This chapter includes an excursion to weak and strong localization and the disorder-induced metal-insulator transition. The interaction between light and matter will be the subject of Chap. 10. Besides the quantum mechanical description of the optical properties of semiconductors and insulators we present the concepts of polaritons and of semiconductor Bloch equations.

1.2 The Structure of Solids Atoms or groups of atoms, when brought into close contact, stick together because of chemical binding. The resulting stable spatial configuration is governed by electrostatics (due to the charged ions) and quantum mechanics (for the electrons). Under given thermodynamic conditions, it represents the state of condensed matter with the lowest total energy, the ground state. For molecules consisting of a small number of atoms, the variety in structure and functionality results from the many possible combinations of diﬀerent atoms from the periodic table. A larger number of identical atoms (or identical groups of a small number of atoms) can form a cluster [32–34] or macromolecule with identical building blocks (e.g., polymer chains). With an increasing number of atoms, this eventually results in macroscopic structures whose physical properties are determined by the chemical nature of the constituents and their configuration in space. These macroscopic structures are understood in a more restricted sense as condensed matter, which comprises matter in the liquid or solid phase. Crystalline structures result from a space-filling periodic repetition of the same building blocks (an atom and its nearest and next-nearest neighbors). These configurations are characterized by a long-range order, which causes sharp Bragg1 peaks in diﬀraction experiments. These Bragg peaks are used to identify crystalline solids (or solids in the crystalline state) by their crystal classes. Crystalline solids will be the primary subject of this book. Some elementary concepts for describing their structure shall be polished up by solving Problem 1.1. Before discussing crystalline structures in more detail, we should mention other forms of condensed matter, namely quasi-crystals [35–37], amorphous solids [38, 39], liquid crystals [40], and soft matter [41–46]. Quasi-crystals exhibit long-range order, but they are not simple periodic structures. Instead, they result from superimposed incommensurate periodicities and are selfsimilar. Examples are the one-dimensional Fibonacci chain (Problem 1.2) 1

Sir William Henry Bragg 1862–1942, Sir William Lawrence Bragg 1890–1971, shared the Nobel prize in physics 1915.

1.2 The Structure of Solids

5

and the Penrose tilings. The latter allows a local fivefold symmetry, which is prohibited in crystalline solids. In contrast, this long-range order is missing completely in amorphous solids which are characterized by disorder in the spatial configuration (structural disorder). The X-ray diﬀraction patterns of quasi-crystals show sharp peaks owing to the long-range order, while those of amorphous solids are diﬀuse. Liquid crystals have long-range order but not in all spatial directions. The building blocks – usually large rod-like or cyclic molecules – are arranged such that a long-range order exists with respect to the orientation of these molecules in at least one direction, whereas in other directions, a liquid structure prevails. Because of their order–disorder phase transitions at room temperature which can be triggered by applied voltages, liquid crystals have seen widespread use in displays and large scale television screens. Finally, we refer to soft matter, a class of materials that comprises foams, polymer melts, biological membranes, and colloid systems. Their particular material properties result from structuring on a mesoscopic scale on which normal liquids and solids are homogeneous. Let us return to the periodic structures. Their systematic description can be found at the beginning of most Solid State Physics books and shall be repeated here only briefly (see Problem 1.1). An infinite periodic structure can be characterized by a point lattice, which in three dimensions is defined by the set of lattice vectors Rn = n1 a1 + n2 a2 + n3 a3

(1.1)

with linearly independent vectors ai , (i = 1, 2, 3), the primitive lattice vectors, and integers ni , (i = 1, 2, 3) combined to n = (n1 , n2 , n3 ). Point lattices in one and two dimensions are defined analogously. While in one dimension there is only one point lattice, there are 5 in two (Problem 1.3) and 14 in three dimensions [29]. The point lattice is used to define the crystal unit cell or its particular choice, the Wigner–Seitz2 cell, which by repetition fills the whole space. Clearly, the lattice structure is mapped onto itself under a translation by a lattice vector. Mathematically, these operations form the translation group of the point lattice (see Appendix). Lattice translations commute with the system Hamiltonian and allow one to characterize the quantum states of the solid by a wave vector k. It corresponds to the linear momentum. However, as the crystalline solid is only invariant under the discrete lattice (and not under infinitesimal) translations, the meaning of this momentum is modified as will be explained below. Therefore, it is called crystal momentum. A crystal structure is obtained by assigning an atom or a group of atoms to each lattice point. The former case corresponds to the Bravais3 lattices. For the latter case, called lattice with basis, the position of the atoms can be 2

3

Eugene Paul Wigner 1902–1995, Nobel prize in physics 1963; Frederick Seitz 1911–2008. Auguste Bravais 1811–1863.

6

1 Introduction

described by Rnτ = Rn + τ

(1.2)

where τ denotes the position of an atom of the basis relative to the lattice point. For each point lattice, a reciprocal lattice is defined by the vectors Gm = m1 b1 + m2 b2 + m3 b3

(1.3)

with integers mj , (j = 1, 2, 3), m = (m1 , m2 , m3 ), and primitive reciprocal translations bj defined by ai · bj = 2πδij . (1.4)

The Wigner–Seitz cell of the reciprocal lattice is the Brillouin4 zone. Thus, within the Brillouin zone, there are no two wave vectors which diﬀer by a reciprocal lattice vector, and k is unique. These k vectors – or the corresponding crystal momentum – characterize the eigenstates of the crystal Hamiltonian, which at the same time are eigenstates of the translation operator (Problem 1.4). Their wavefunctions, which according to Bloch’s theorem are modulated plane waves, yield equal probabilities in each Wigner–Seitz cell. Therefore, they represent extended states. Among the diﬀerent Bravais lattices, the face-centered cubic (fcc) and the body-centered cubic (bcc) are most frequently realized in normal, noble, and transition metals. One can easily check (see Problem 1.1) that these two lattices are reciprocal to each other. Their Brillouin zones, which will be used in later chapters, are shown in Figs. 1.1 and 1.2. Owing to the point lattice, the shape of the Brillouin zone exhibits a high symmetry under operations such as rotations, reflections, and inversions, which form the point group (see Appendix). Diﬀerent points and lines of high symmetry in the Brillouin zone, denoted by letters, e.g., Γ, ∆, Λ, X, W, H, are invariant under subgroups of

kz

P Λ

kx

ΓΣ

N

∆

H

ky

Fig. 1.1. Brillouin zone of the body-centered cubic (bcc) lattice 4

L´eon Brillouin 1889–1969.

1.2 The Structure of Solids

7

kz

L

Λ ∆

Γ

kx

Σ K

W

X

ky

Fig. 1.2. Brillouin zone of the face-centered cubic (fcc) lattice

the point group, consisting of all operations under which k is not changed, the group of the wave vector. For the Γ point, the center of the Brillouin zone (k = 0), this group is identical with the point group of the crystal. The energies of quantized lattice vibrations, of electrons or spin waves will be drawn in later chapters as a function of the crystal momentum along these lines. The fcc lattice is the point lattice of the diamond and zinc blende structures, in which the technologically most important semiconductors such as Si, Ge, and GaAs crystallize: both are lattices with a basis of two atoms (at τ = ±a(1, 1, 1)/8, where a is the cubic lattice constant), which are identical (from the fourth group of the periodic table) for diamond, but diﬀerent (e.g., from the third and the fifth group in A3 B5 compounds) for zinc blende. Also, the rocksalt structure, in which most alkali halides crystallize, derives from the fcc lattice, but with the two diﬀerent ions placed at τ = ±a(1, 1, 1)/4. As demonstrated in Problem 1.5, these crystal structures lead to quite diﬀerent space filling and coordination numbers, which signalize distinct types of chemical binding. More complex crystal structures with a basis of many atoms are realized, for instance, in molecular crystals and high-Tc superconductors. The crystal structure is invariant under the operations of the space group, which is composed of translations and operations of the point group. Crystal eigenstates can be classified by exploiting these symmetries within the concepts of group theory [47–50]. A short introduction will be given in the Appendix A.1. But even within the simple crystal structures, a variety of solids can be realized in mixed crystals when the positions Rnτ (1.2) are randomly occupied, e.g., by diﬀerent kinds of atoms with the same number of valence electrons but from diﬀerent rows of the periodic table. In this case, the periodicity of the point lattice still exists, but there is disorder in the occupation of the sites (compositional disorder). It is to be noted that already the

8

1 Introduction

occupation of the lattice positions by diﬀerent isotopes of the same element represents a case of compositional disorder. Compositional disorder is to be distinguished from that of amorphous material, where atoms (or groups of atoms) of the same kind are placed without long-range order to create a solid with structural disorder. Disorder will become important in Chap. 9 with the discussion of localized states whose wave functions yield a probability which decays exponentially away from its maximum. More recently, periodic structures have become an issue in connection with photonic crystals [51–54]. Just as the lattice periodicity influences the electrons by Bragg reflections and causes the electronic band structure (Chap. 5), a periodic arrangement of dielectric matter or a periodicity in the refractive index gives rise to a photonic band structure. This modification of the dispersion relation of electromagnetic waves results in a wavelength dependent reflectivity which is not only observed in man-made structures, but can also be the reason for color phenomena in the natural world [55]. Any real solid has a finite size. Its surface is the border that separates the interior of the solid from the investigator. All experimental knowledge about a solid is based on information inter-penetrating this border and should be checked with respect to possible artifacts caused by the surface. On the other hand, the surfaces of solids – their structure, dynamics, and functionality – comprise an important part of solid state physics with a high potential for applications [56–60]. The surfaces of crystalline solids have a two-dimensional periodic structure and can be understood as a solid with reduced dimensionality. They are defined by those terminating layers of the solid whose atoms experience a surrounding that diﬀers from that of the bulk atoms. Usually, these are the outermost two or three layers. With the changing surrounding and chemical binding, the surface atoms can take positions diﬀerent from those of the bulk material given by (1.2): The lattice spacing or even the lateral periodicity can change. These changes are called relaxation and reconstruction, respectively. Some experimental techniques are more sensitive to surface eﬀects than others and some are especially designed to study the surface of solids. Besides surface sensitive optical methods and particle scattering processes, scanning tunneling microscopy (STM), invented by Binnig and Rohrer5 [61], has become a diversified tool in surface physics within a short period of time. These methods allow one not only to investigate the surface of solids but also to manipulate individual atoms on the surface. The eye-catching pictures of company logos constructed by arranging individual atoms or molecules on a clean surface or of a quantum corral have gained some publicity [58, 62]. In the world of nanophysics [63], STM and its variants (like atomic force microscopy) play an important role.

5

Gerd Binnig *1947, Heinrich Rohrer *1933, shared the Nobel prize in physics 1986.

1.3 Pair-Distribution Function and Structure Factor

9

1.3 Pair-Distribution Function and Structure Factor In order to characterize the structure of a system that consists of a very large number of constituents, it is not necessary to know the position of each individual particle. Instead, one can use a quantity which describes the probability of finding pairs of particles with given relative positions: the pair-distribution function. As we will see, this suﬃces to distinguish between a gas, a liquid or amorphous solid, or a crystalline solid. Let us consider a system of N particles in a volume V (in the context of this chapter, it will be the atoms or ions of a solid, but it could also be the electrons in a plasma or the galaxies in the universe) at positions r i , with i = 1, ...N . The particle density is given by n(r) =

N i=1

δ(r − ri ).

(1.5)

The correlations of positions or of the density are quantified by the density– density correlation function Cnn (r, r ′ ) = n(r + r ′ )n(r ′ ) , (1.6) where ... denotes the thermal average or quantum-mechanical expectation value, depending on the situation. By integration of (1.6) and normalizing we obtain the function 1 d3 r′ p(r) = δ(r + r ′ − ri )δ(r ′ − rj ) N ij 1 δ(r + r i − rj ) = N ij 1 δ(r + ri − r j ) , (1.7) = N δ(r) + N ij i=j

where the last term contains the information about the distribution of pairs i = j with relative position ri − r j . Replacing the δ-function by its Fourier transform (see Appendix A.2) leads to 1 iq·r 1 iq·(ri −r j ) . (1.8) e e p(r) = V q N i,j For q = 0, the double sum over the sites gives N 2 irrespective of the configuration. One separates this term to introduce the static structure factor [64] 1 iq·(ri −rj ) − N δq,0 . S(q) = e (1.9) N i,j

10

1 Introduction

We may also write ⎛ ⎞ 1 ⎝ eiq·(ri −rj ) ⎠ 1+ N δq,0 + S(q) = N i j =i ′ N d3 r′ g(r ′ )eiq·r , = 1+ V V

(1.10)

where g(r) is the pair-distribution function [16, 64]. The last equation can be solved for g(r) and yields g(r) = 1 +

1 iq·r S(q) − 1 , e N q

(1.11)

The pair-distribution function gives the probability of finding a particle at r if there is a particle at r = 0. In general g(r = 0) = 0 because of strong shortrange repulsive forces, which prevent two particles from occupying the same position. For a crystalline solid with atoms at the fixed positions Rnτ , i.e., without thermal motion, g(r) has sharp peaks for r = r i − rj = Rnτ . This feature is a consequence of the long-range order or density–density correlation in the crystal lattice. When this order is relaxed, the peaks will become diﬀuse. This happens because of thermal motion and is described by the Debye–Waller factor (see Sect. 3.7), but is the case also for amorphous solids or liquids due to structural disorder. If long-range and short-range order is absent, as in a gas, g(r) is constant for |r| > d, where d is a characteristic parameter of the particles, e.g., the hard-core diameter. Let us have a look also at the static structure factor written as S(q) = d3 r p(r)e−iq·r 1 = d3 r δ(r + ri − r j )e−iq·r N i,j 1 iq·ri −iq·rj 1 iq·(ri −rj ) . (1.12) = e e e = N i,j N j i For a crystalline solid with the ions at rest, one has ri − r j = Rnτ with Rnτ from (1.2) giving 1 iq·τ iq·Rn S(q) = . (1.13) e e N τ n

Because of the periodic structure, the phases of the last sum cancel each other except for q = G, where G is a vector of the reciprocal lattice, and we have n exp (iq · Rn ) = N G δq,G . Thus, the structure factor S(q) = δq,G eiq·τ (1.14) G

τ

1.3 Pair-Distribution Function and Structure Factor

11

is characterized by sharp peaks at the reciprocal lattice vectors G. They specify sets of parallel crystal planes: Their orientation is given by the direction of G and their spacing by 2π/|G|. The static structure factor determines the cross section for X-ray scattering, which is the standard method to identify the crystal structure of a solid (Problem 1.6). The structure factor can also be expressed in terms of the Fourier transform (see Appendix) of the density 1 −iq·rj 1 nq = d3 r n(r)e−iq·r = , (1.15) e V V j which, for q = 0, describes the deviations from the average particle density nq=0 = n = N/V called density fluctuations. In terms of the corresponding number fluctuations Nq = nq V , the static structure factor can be written as S(q) =

1 Nq N−q . N

(1.16)

For describing correlations between time-dependent positions of particles, e.g., due to thermal motion of the ions, one uses the dynamic structure factor 1 iq·r i (t) −iq·r j (0) −iωt S(q, ω) = e dt, (1.17) e e 2π i,j which, in terms of number fluctuations, can also be written as 1 e−iωt Nq (t)N−q (0)dt. S(q, ω) = 2π

(1.18)

As we shall see later, thermal motion reduces the intensity of the Bragg peaks and leads, if it gets suﬃciently strong, to their complete suppression as an indication of loss of structural correlation. This is the transition to the liquid phase. The functions, S(q) and g(r), containing the full information about the 3-dimensional structure, are diﬃcult to plot due to their dependence on vectors. For practical reasons, this information is reduced to the radial distribution function g(r), which determines the number of particles in a spherical shell of radius r and thickness dr: ng(r)4πr2 dr. Plots of the radial distribution function and the corresponding structure factor are shown in Fig. 1.3 for a gas, a liquid or an amorphous solid, and a crystalline solid. For a gas with random positions of the atoms, there is equal probability in the radial distribution for r > d, where d is the minimum distance between pairs of particles due to repulsive forces (e.g., the diameter of hard spheres). In contrast, the crystalline solid with long-range order due to the periodic arrangement is characterized by a radial distribution that exhibits peaks corresponding to the coordination shells of nearest, next-nearest etc. neighbors. Amorphous solids

12

1 Introduction

S(q)

g(r)

gas

1 d

amorphous

1

solid

crystal

r

q

Fig. 1.3. The radial pair-distribution function (left) and static structure factor (right) of a gas, an amorphous solid, and a crystalline solid

(and liquids) still have short-range order and the nearest coordination shells are visible, while others are washed out with increasing r because of missing long-range order. The structure factor exhibits the same features as can be understood from (1.11). The more pronounced structures in these functions indicate the strength of correlation in the density or relative positions. In this section, we have used the pair-distribution function and the structure factor to characterize and distinguish diﬀerent forms of condensed matter with respect to correlations in their ion configuration. The concept of analyzing the pair-distribution function of a particle system is, however, quite general and will be applied in Chaps. 4 and 7 to identify and quantify correlation in electron systems.

Problems 1.1 The following concepts are frequently used in solid state physics: point (Bravais) lattice, lattice vector, reciprocal lattice, Wigner–Seitz cell, Brillouin zone. What is the meaning of these quantities? Give explicit descriptions (by formulas or drawings) for a two-dimensional square lattice, a simple cubic (sc), a face-centered cubic (fcc), and a body-centered cubic (bcc) lattice. 1.2 Consider the one-dimensional model system constructed from two building blocks of diﬀerent lengths, S (for short) and L (for long), starting from the configuration LS by the following rule: replace S by L and L by LS. Construct the chain by repeated application of the rule to obtain what is

Problems

1.3 1.4 1.5

1.6

13

called a Fibonacci chain. Realize, that it is not periodic and that it has the same configuration when L = S ′ is taken as the short and LS = L′ as the long building block! This property is called self-similarity or fractality. Although not being periodic, the structure has long-range order as can be seen by taking the Fourier transform. Find the five diﬀerent possible point lattices in two dimensions. Explain the meaning of Bloch’s theorem. On which symmetry is it based, and to which constant of motion is it related? Give the coordination numbers and calculate the relative spatial filling in a simple cubic, a face centered cubic, a body centered cubic, and a diamond lattice, by considering non-overlapping spheres of maximal diameter at each lattice point for fixed lattice constant of the elementary cube. The structure of solids can be investigated by diﬀraction experiments with photons, electrons or neutrons. The cross section for elastic scattering is determined by the scattering amplitude ′ F (k, k′ ) = n(r)ei(k−k )·r d3 r, (1.19) where n(r) is the mass density and k(k′ ) is the wave vector of the incident (scattered) wave. Make use of the periodic mass density in a crystalline solid to find the values k − k′ for which the scattering amplitude does not vanish. Give a geometrical interpretation of the result with respect to the meaning of the direction and length of reciprocal lattice vectors.

2 The Solid as a Many-Particle Problem

As mentioned in Sect. 1.1, we understand the solid as being composed of ions (nuclei and closed electron shells) and valence electrons. A more rigorous approach would start from nuclei and electrons, but a simple consideration of the spatial extension of electrons in diﬀerent shells of the isolated atoms shows immediately that this is not necessary. The wave functions of electrons in inner shells (the core electrons) with binding energies of hundreds or thousands of eV extend over a distance much smaller than the lattice spacing in a solid, as visualized in Fig. 2.1. In fact, when the atoms are assembled into the configuration of a crystal lattice (or likewise of a molecule, cluster, liquid. . .) it will be the outermost, weakly bound valence electrons which first experience the presence of their nearest neighbors. They will rearrange from their states in the isolated atoms into those which establish the chemical binding. Together with the electrostatic energy of the ion configuration, this defines the stable structure. Some textbooks on Solid State Theory start with a detailed description of this structure of crystalline solids (e.g., [4, 7, 9, 11]) which is only briefly repeated here. Instead, we follow the approach of [5, 14, 21] with a presentation of the basic Hamiltonian, which defines the solid as a quantum-mechanical many-body problem. The eﬀectiveness of chemical binding depends on the overlap of the electronic wave functions at neighboring lattice sites and on their coordination number. Thus, metals prefer a close-packed structure, namely the bodycentered cubic (bcc) and face-centered cubic (fcc) lattices, with delocalized electrons acting as glue between the positively charged ions (metallic binding), while in (binary) ionic crystals, electrons are transferred from the cation to the anion to complete their outer shells (ionic or heteropolar binding) and form lattices dominated by electrostatic interaction (like the rocksalt structure). Rare gases with closed shell configurations as well as larger molecules form crystalline solids due to the weak van der Waals1 forces and are stable only at low temperatures. Elements of the fourth group of the periodic table 1

Johannes Diderik van der Waals 1837–1923, Noble prize in physics 1910.

U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 2,

16

2 The Solid as a Many-Particle Problem

V(r), Ψ(r) r valence electrons

core electrons

Fig. 2.1. Schematic view of a solid: periodic potential (dashed line) and wave functions of core and valence electrons (solid lines) drawn at their respective energy levels

share each of their four valence electrons with the four nearest neighbors in directed covalent bonds (covalent or homopolar binding), which results in the diamond structure. A mixture of covalent and ionic binding, whereby the contribution of the latter increases with the polarity of the material, is typical for the zinc blende structure realized in A3 B5 , A2 B6 , and A1 B7 compounds. A dominant covalent binding is typical for semiconductors. In most cases, the distinction between valence and closed shell electrons is justified by the large energy separation, which can be detected by photo– electron spectroscopy (PES [65, 66]). With reference to the heavier elements (including transition-metals and rare-earths) and their compounds, for which d states are in the same energy range as the s like valence electrons, one should keep in mind, however, that even complete d shells can participate in the chemical binding and influence the electronic structure. In Chap. 5, we shall come back to this case but will rely here on a clear separation between closed-shell ions and valence electrons.

2.1 The Hamiltonian of the Solid These introductory considerations about the composition of a solid out of valence electrons and ions justify writing the Hamiltonian as H0 = Hion + Hel + Hel−ion .

(2.1)

The first term depends only on the coordinates of the ions and reads Hion =

NI NI P 2i 1 + V (Ri − Rj ) 2Mi 2 i,j=1 i=1

(2.2)

i=j

where P i , Ri , and Mi denote the momentum, position, and mass of the ith ion, respectively, and NI is the number of ions in the crystal volume Vc .

2.1 The Hamiltonian of the Solid

17

V (Ri −Rj ) can be a quite general ion–ion interaction potential, but for pointlike ions with charge Qi = sign(Qi )Zi e and positive integer Zi , it will be the Coulomb interaction V (Ri − Rj ) =

Qi Qj . 4πε0 |Ri − Rj |

(2.3)

Here, ε0 is the vacuum dielectric constant. The assumption of point-like ions is justified for closed shells with tightly bound electrons. This situation is visualized in Fig. 2.1. The electron part of H0 is given by Hel =

Ne Ne p2l e2 1 + 2m 2 k,l=1 4πε0 |r l − rk | l=1

(2.4)

k=l

where pl and r l are the momentum and position of the lth electron, m is the free electron mass, and Ne is the number of electrons in the crystal volume. The charge neutrality of the solid requires that NI i=1

Qi − eNe = 0 .

(2.5)

The interaction between electrons and ions is described by Hel−ion =

NI Ne l=1 i=1

v(r l − Ri ) ≃

NI Ne l=1 i=1

Qi e 4πε0 |r l − Ri |

(2.6)

where the last expression is the approximation for point-like ions. It is conceivable that the physical properties of the system described by H0 do not depend on NI,e as long as these numbers are suﬃciently large. These properties will be considered in the thermodynamic limit NI,e → ∞, with the corresponding densities NI,e /Vc kept constant. This situation, valid for macroscopic solids, is to be distinguished from that of mesoscopic systems for which clusters are a representative example. The number of their surface ions is comparable with that of the bulk ions with the consequence that their properties depend on the shape defined by the surface. The physics of clusters has become a research field in its own right [32–34]. Nevertheless, it is interesting to note that the properties of macroscopic solids can be thought of as evolving with increasing size from those of clusters as the ratio of the number of surface ions divided by the number of bulk ions goes to zero. So far the spin of electrons or ions did not appear because neither the kinetic energy nor the Coulomb interaction depends on spin. The expression for the kinetic energy is that of the non-relativistic formulation. In principle, one could have started, e.g., for the electrons, from the relativistic Dirac operator (which is in fact used in band structure theory and becomes relevant for

18

2 The Solid as a Many-Particle Problem

solids composed of atoms with large Zi , see Chap. 5). However, in most cases, it is suﬃcient to consider just the spin-orbit coupling in the electron part of the Hamiltonian Hso =

¯ h (∇l Veﬀ (r l ) × pl ) · σ l 4m2 c2

(2.7)

l

where Veﬀ (r l ) is the eﬀective single-particle potential, which will be introduced in Chap. 5, and σ l is the vector of the Pauli spin matrices of the lth electron. A similar term can be added to the ion part of the Hamiltonian, in order to account for the nuclear spins, if their influence is addressed. But even without spin–orbit coupling, the spin becomes important in the presence of a magnetic field, when the Zeeman term causes removal of spin degeneracy, and, because of the fermion character of the electrons, by considering the Pauli principle, when dealing with the many-particle aspect of the problem (see Chap. 4). Moreover, spin alignment and spin excitation are essential for magnetic properties (see Chap. 6). As mentioned already in Sect. 1.2, we consider a solid primarily for the case of the crystalline periodic order of the ions. Only in Chap. 9, we shall remove this restriction by allowing structural disorder and its eﬀect on the electronic eigenstates and transport. The perfect crystalline configuration – which, in order to simplify the notation, is assumed here to be a Bravais lattice – is characterized by ion positions Ri forming a point lattice (see Sect. 1.2) R0n = n1 a1 + n2 a2 + n3 a3

(2.8)

where the upper index 0 indicates that these vectors now mark the equilibrium positions, while the actual position of an ion is given by Rn = R0n + un

(2.9)

with a displacement un accounting for the motion of the ion about its equilibrium position. These displacements will be assumed to be small in comparison with the lattice constant to motivate a separation of the ion–ion potential according to V (Rn − Rm ) = V (R0n − R0m + un − um ) = V (R0n − R0m ) + δV (Rn − Rm ),

(2.10)

where V (R0n − R0m ) is determined by the equilibrium configuration and δV can be explicitly expressed as a Taylor series in the displacements. In the same way we may proceed with the electron–ion potential v(r l − Rn ) = v(r l − R0n ) + δv(r l − Rn ).

(2.11)

2.1 The Hamiltonian of the Solid

19

The Hamiltonian of the solid may now be written as

H0 =

P2 1 n δV (Rn − Rm ) + 2Mn 2 n,m n n=m

Ne Ne e2 p2l 1 + + v(r l − R0n ) + 2m 2 k,l=1 4πε0 |rl − r k | l=1

n,l

k=l

1 + V (R0n − R0m ) + δv(r l − Rn ). 2 n,m

(2.12)

n,l

n=m

The first line of (2.12), depending only on the positions and momenta of the heavy constituents of the solid, will be the subject of Chap. 3, the lattice dynamics. In the quantum mechanical version, it is characterized by phonons, the quanta of lattice vibrations, which determine the elastic properties, the specific heat, and the optical properties in the far-infrared spectral range. The second line, describing the electrons in a static periodic lattice potential and their mutual interaction, will be the starting point of Chaps. 4–7. In Chap. 4, we will present the properties of the homogeneous electron gas (for which the periodic potential will be smeared out into a homogeneous positive charge background defining the jellium model) and learn how to treat the electron–electron interaction. In Chap. 5, the single-particle concepts of band structure, which allow one to distinguish metals from insulators will be introduced. Also, spin dynamics and magnetic properties (Chap. 6) and correlated electron systems (Chap. 7) will be treated, essentially by starting from the second line of (2.12). The third line of (2.12) contains two contributions. The first one is the electrostatic interaction energy of the ions in their equilibrium configuration, representing a constant energy which becomes important in total energy calculations. For a binary cubic lattice, with a basis consisting of two point-like ions with opposite charge ±Ze (e.g., the zinc blende and rocksalt structure), this term can be expressed using 1 1 (Ze)2 (Ze)2 = αM , sign(Qi )sign(Qj ) 2 4πε0 |Ri − Rj | 2 4πε0 R

(2.13)

j =i

where |Ri − Rj | = Rpij , with the nearest neighbor distance R, and αM =

j =i

sign(Qi )sign(Qj )

1 pij

(2.14)

is the Madelung2 constant. It is independent of the reference lattice point i [29]. The second contribution depends on the actual positions of the ions 2

Erwin Madelung 1881–1972.

20

2 The Solid as a Many-Particle Problem

and electrons and couples the dynamics of both subsystems, thus giving rise to electron–phonon interaction (Chap. 8). It is important for electron transport properties, relaxation phenomena, and superconductivity. The structuring of H0 as given in (2.12) is by purpose quite suggestive: Without the third line, the electron and ion subsystems are separated and can be considered as independent. We will discuss in the next section the conditions under which such an approximation is justified.

2.2 Separating the Motion of Electrons and Ions In order to describe a solid and its properties, we have to solve the quantum mechanical problem: ˜ = i¯ ˜˙ with hΨ H0 Ψ

˜ =Ψ ˜ ({rl }, {Rn (t)}) . Ψ

(2.15)

Here, {rl }, {Rn (t)} denote the configurations of the electrons and ions of the system, of which the latter is considered time-dependent because of the thermal motion of the ions. Employing thermodynamic concepts, electrons and ions can be considered as two systems which are in contact as mediated by the last term of (2.12). Let us assume that the thermal equilibrium of the whole system, i.e., electrons and ions have the same temperature and the same average thermal energy per degree of freedom. This can be quantified by the mean values of the kinetic energy and we may write 2 2 pl Pn = . (2.16) 2m 2M

Because of the mass ratio m/M ≃ 10−4 , this equation indicates that the electrons move much faster than the ions. Thus, the electrons will experience any actual configuration of the ions (caused by thermal fluctuation) as a stationary potential to which they adjust instantaneously by adopting the state of lowest energy. This fact can be used to formulate the conditions for separating the motions of ions and electrons. Consider the stationary Schr¨ odinger3 equation of the electrons moving in the electrostatic potential, defined by the actual configuration of the ions, (Hel + Hel−ion ) Ψα = Eel,α Ψα

(2.17)

as being solved. The eigensolutions Ψα ({rl }, {Rn }) and energy eigenvalues Eel,α ({Rn }) are characterized by a complete set of quantum numbers α and depend parametrically on the ion configuration {Rn }. For a given configuration, the Ψα form a complete set and can be used to expand the solution of the time-dependent problem (2.15) in the form 3

Erwin Schr¨ odinger 1887–1961, shared in 1933 the Nobel prize in physics with P.A.M. Dirac.

2.2 Separating the Motion of Electrons and Ions

˜ ({rl }, {Rn (t)}) = Ψ

α

Φα ({Rn }, t)Ψα ({rl }, {Rn })

21

(2.18)

with the expansion coeﬃcients Φα depending on the ion configuration and on t. Then the full problem (2.15) reads ˜ = ˙ α . (2.19) H0 Ψ (Hel + Hel−ion )Ψα Φα + Hion Ψα Φα = i¯ h Ψα Φ α α =Eel,α Ψα

Using the completeness of the Ψα , we eliminate the electron coordinates by multiplying from left with Ψ∗β and integrating over all rl to obtain ˙β . Eel,β Φβ + . . . Ψ∗β Hion Ψα d3 {rl }Φα = i¯ hΦ (2.20) α

The operator Hion contains the momentum operators P n which act on the ion coordinates of both Ψα and Φα . Thus the integration over electron coordinates in the second term of (2.20) can be performed by writing (as operator equation applied to Φα ) 1 2(P n )βα · P n + (P 2n )βα , (2.21) . . . Ψ∗β Hion Ψα d3 {rl } = 2M n

where the matrix elements (for κ = 1, 2) κ (P n )βα = . . . Ψ∗β ({rl }, {Rn })P κn Ψα ({r l }, {Rn })d3 {rl }

(2.22)

couple between diﬀerent electron states Ψα , Ψβ . When writing the momentum in atomic units (Bohr radius aB = 4πε0 ¯ h2 /me2 and Rydberg energy 1 Ry = 2 me4 /(4πε0 )2 2¯h ), P n = P ′n ¯ h/aB , this coupling, expressed in terms of

m ′2 P2 = P Ry, 2M M is seen to carry a factor m/M ≪ 1, thus representing a small term. If this coupling is neglected, one arrives at the Born–Oppenheimer4 approximation. It is sometimes also called adiabatic approximation because the electrons follow the (slow) motion of the ions adiabatically, i.e., without changing their eigenstate. Within this approximation, the motion of the electrons is separated from that of the ions, and we may write for these two systems, the equations (Hion + Eel,α )Φα = i¯ hΦ˙ α

(2.23)

with Hion + Eel,α =

P2 1 n + Eel,α ({Rn }) + V (Rn − Rm ) 2M 2 n,m n

(2.24)

n=m

4

Max Born 1882–1970, Nobel prize in physics 1954; J. Robert Oppenheimer 1904– 1967.

22

2 The Solid as a Many-Particle Problem

and (see (2.17)) (Hel + Hel−ion ) Ψα = Eel,α Ψα .

(2.25)

Equation (2.23) is the time-dependent Schr¨ odinger equation for the ions which move in the adiabatic potential 1 V (Rn − Rm ) (2.26) Uα ({Rn }) = Eel,α ({Rn }) + 2 n,m n=m

defined by the ion configuration and the energy of the eigenstate α of the electron system. This problem will be studied in more detail in Chap. 3. Equation (2.25) is the stationary eigenvalue problem of the electron system for a given ion configuration. We will discuss its solution in Chaps. 4–7 with the simplification of always replacing {Rn } by the equilibrium configuration {R0n }. It should be mentioned that the Born–Oppenheimer approximation does not apply only to solids but also to molecules.

2.3 Thermal Expectation Values One of the important aims of a theory is to explain experimental results and possibly arrive at a quantitative description of measurements. Experiments yield information about physical observables, which in quantum formulation are Hermitian operators Aˆ = Aˆ† . Thus, our theory has to aim at the descripˆ of such operators. For solids, which are tion of the expectation values A macroscopic thermodynamic systems, they have to be understood as thermal expectation values. Therefore, a brief repetition of the related concepts of statistical mechanics is required. Simple quantum mechanical systems (e.g., an atom, an oscillator) can be prepared in a state |i and the expectation (or mean) value is given by ˆ = i|A|i. ˆ A

(2.27)

The state |i can be an eigenstate or a mixture of eigenstates and thus, in repeated measurements, the measured values will be sharp (an eigenvalue ˆ or fluctuate around the mean value, respectively. A solid, consisting of A) of many particles, is not a simple system and cannot be prepared in a well defined quantum state. Instead, being a thermodynamic system, the solid can be prepared in a state characterized by a set of thermodynamic variables like temperature T , particle number N , volume V . These thermodynamic quantities specify the state of the solid as a statistical ensemble of quantum mechanical micro-states {|i, i = 1, . . . I; T, N, . . . fixed}, and the experimental values for the observable Aˆ are described by the thermal expectation value (or ensemble mean value) ˆ = A

I 1 ˆ i|A|i. I i=1

(2.28)

2.3 Thermal Expectation Values

23

Making use of a complete orthonormal set of states |n and of the property that factors under the trace operation Tr(. . . ) = n n| . . . |n can be interchanged, we may formulate the expectation or thermal mean value as I 1 ˆ = ˆ A |ii| A|n (2.29) n| I i=1 n or in short

where

ˆ = Tr ρˆAˆ , A

(2.30)

I

1 ρˆ = |ii| I i=1

(2.31)

is the statistical operator. Let us repeat the important properties of the statistical operator: 1. It is Hermitian: ρˆ = ρˆ† 2 ≥ 0, for arbitrary |ψ 2. It is positive semi-definite: ψ|ˆ ρ|ψ = 1I i |ψ|i| 1 1 3. It is normalized: Trρˆ = I n,i n|ii|n = I i i| ( n |nn|) |i = 1 in any representation. These properties qualify ρˆ as the operator of probability distributions. It satisfies the eigenvalue equation ρˆ|n = pn |n

(2.32)

with the eigenvalue pn = n|ˆ ρ|n =

1 |i|n|2 , I i

(2.33)

which quantifies the probability of finding the state |n in the ensemble of micro-states {|i . . . .}. In thermal equilibrium, one has 1 ˆ ρˆ˙ = [H, ρˆ] = 0, i¯ h

(2.34)

ˆ is the Hamiltonian of the system in which the states |i are realized where H ˆ the (in our case that of the solid). In the representation of eigenstates of H, ˆ statistical operator, which commutes with H, is diagonal (Problem 2.1). The following statistical ensembles are of importance in solid state physics and will be used throughout the book:

24

2 The Solid as a Many-Particle Problem

1. The canonical ensemble with fixed temperature T and particle number N (β = 1/kB T where kB is the Boltzmann5 constant): ρˆ =

1 −β Hˆ e Z

(2.35) ˆ

with the canonical partition function Z = Tr(e−β H ). In the energy representation ρˆ becomes pn = Z −1 exp (−βEn ) . 2. The grand-canonical ensemble with fixed temperature T and variable particle number N : ρˆG =

1 −β(H−µ ˆ ˆ) N e ZG

(2.36) ˆ

ˆ

with the grand-canonical partition function ZG = Tr{e−β(H−µN ) }, where ˆ is the particle number operator and μ the chemical potential. In the N −1 exp (−β(En − μNn )), where energy representation, ρˆG becomes pn = ZG Nn is the particle number in the state |n. In the next section, the thermal expectation values of observables will be evaluated, especially for thermal equilibrium formulated in terms of eigenstates of the Hamiltonian H0 (see (2.12)). A particular situation is obtained for very low temperatures, at which the thermal expectation value becomes the ground state expectation value (Problem 2.2). When considering the number operator as observable for a system in thermal equilibrium, one obtains the well-known distribution functions of the Fermi–Dirac6 or Bose–Einstein7 statistics (depending on the system), and of their high-temperature limit, the Maxwell distribution (Problem 2.3).

2.4 Theory of Linear Response Any experiment constitutes a perturbation of the system under investigation: By scattering light or particles, we obtain information on the structure of the solid or of its characteristic excitations; by applying an electric or magnetic field, we study the transport or magnetic properties; by probing with light, we detect the optical properties. If we do this to characterize the material, these perturbations have to be weak and must not change the system properties. This defines the regime of linear response. The system (in our case the solid) plus the external perturbation applied to investigate its properties, is described by the Hamiltonian H = H0 + Vext . 5 6

7

(2.37)

Ludwig Boltzmann 1844–1906. Enrico Fermi 1901–1954, Nobel prize in physics 1938; Paul Adrien Maurice Dirac 1902–1984, Nobel prize in physics 1933. Satyendra Nath Bose 1894–1974; Albert Einstein 1879–1955, Nobel prize in physics 1922.

2.4 Theory of Linear Response

25

Table 2.1. Examples of observables used in the response formalism ˆ B ˆ A,

Vext

F

Response function

Electric current density Dielectric polarization Magnetic polarization Heat current density

−j · A −P · E −m · H −v · ∇T

Electric field Electric field Magnetic field Temp. gradient

Electric conductivity Dielectric function Magnetic susceptibility Heat conductivity

H0 is the Hamiltonian (2.12) of the unperturbed solid and the perturbation, (assumed for simplicity as being independent of space variables) can be written as. ˆ (t). Vext (t) = −BF (2.38)

ˆ is an observable and F (t), a (in general, time-dependent) scalar funcHere, B tion. We may distinguish between dynamic (time-dependent) and static (timeˆ independent) perturbations. Consider the measurement of an observable A. The measured values can be described by ˆ t = Tr ρˆAˆ = dt′ R(t, t′ )F (t′ ) . A (2.39)

They are ruled by a linear response function R(t, t′ ), which is expected to ˆ It will turn out to be a correlation function of these depend on Aˆ and B. observables as shown later in this section. But before doing this, let us look at the examples of experimental situations listed in Table 2.1 together with their translation into the response formalism. To determine the electric conductivity as a material property of a solid one has to design a measurement of the electric current density j by exposing the sample to an electric field, which can be formulated as the time-derivative of the vector potential A. The perturbation results from the minimal coupling according to which the particle momenta in H0 are replaced by p + eA which, to lowest order in A, leads to Vext (t) = −j · A(t) with the electric current density j = −e l pl /m. Thus the electric conductivity is a current–current (or velocity–velocity) correlation function. Due to the vector character of j and A, the response function is a second rank tensor. Likewise, the dielectric function, characterizing the optical and dielectric properties, follows from measuring the dielectric polarization by probing with an electric field as perturbation. The dielectric function will turn out to be a correlation function between polarizations (or electric dipole moments). A similar situation leads to the magnetic susceptibility. A heat current is caused by a temperature gradient; its measurement provides the heat conductivity. In generalizing (2.39) to also include dependence on space variables, we may write ′ ˆ A(r)t = dt d3 r′ R(r, t; r ′ , t′ )F (r ′ , t′ ). (2.40)

26

2 The Solid as a Many-Particle Problem

For homogeneous systems, the response function depends only on r − r′ and, therefore, a spatially harmonic external perturbation of the form F (r′ , t′ ) = Fq (t′ ) exp (iq · r ′ ) leads to the same spatial dependence of the observable, which in brief means ˆ Aq t = dt′ R(q; t, t′ )Fq (t′ ) (2.41) with R(q; t, t′ ) being the Fourier transform of R(r − r ′ ; t, t′ ). Let us return to the task of finding an expression for the thermal expecˆ which is linear in the perturbing field F (t) for the system tation value of A, described by the Hamiltonian H of (2.37). In view of (2.39), the dependence on the field enters through the statistical operator, which contains the system Hamiltonian including the external perturbation. Without this perturbation, the system is in equilibrium, and the statistical operator satisfies the equation of motion (in the following theˆ-sign, indicating operators, is dropped) [H0 , ρ0 ] = i¯ hρ˙ 0 = 0.

(2.42)

If the perturbation is switched on, the system is driven out of the thermal equilibrium described by ρ0 , and the statistical operator will become timedependent: ρ0 → ρ(t) = ρ0 + ∆ρ(t). (2.43) In order to find the deviation ∆ρ(t) from equilibrium, caused by the perturbation, we have to solve the equation of motion for ρ(t): [H, ρ(t)] = i¯ hρ(t). ˙

(2.44)

Looking for the first order perturbation correction to the equilibrium distribution, we rewrite this equation as [H0 + Vext , ρ0 + ∆ρ(t)] = [H0 , ρ0 ] + [H0 , ∆ρ(t)] hρ(t) ˙ +[Vext , ρ0 ] + [Vext , ∆ρ(t)] = i¯

(2.45)

and keep only those terms, which are of first order in the perturbation (i.e., we neglect [Vext , ∆ρ]) to write 1 ∂ ρ(t) − [H0 , ∆ρ(t)] . [Vext , ρ0 ] ≃ i¯ h (2.46) ∂t i¯ h By multiplying from left and right with the proper exponentials we change from the Schr¨ odinger into the interaction picture (indicated by an overbar), which for the deviation from the equilibrium distribution reads i

i

∆¯ ρ(t) = e h¯ H0 t ∆ρ(t)e− h¯ H0 t .

(2.47)

Note, that the time dependence of ∆ρ(t) on the rhs (in the Schr¨ odinger picture) is that of the external field, while that on the lhs also includes the time

2.4 Theory of Linear Response

27

evolution due to H0 . In the interaction picture, the equation of motion for the first order correction ∆¯ ρ to the equilibrium distribution reads i¯ h

∂ i i ∆¯ ρ(t) = e h¯ H0 t [Vext (t), ρ0 ]e− h¯ H0 t . ∂t

(2.48)

It is an inhomogeneous linear diﬀerential equation indicating a linear relation between ∆¯ ρ(t) and the external perturbation Vext . By direct integration one finds ′ ′ i i 1 t ∆¯ ρ(t) = (2.49) e h¯ H0 t [Vext (t′ ), ρ0 ]e− h¯ H0 t dt′ , i¯ h −∞ and in the Schr¨ odinger picture ′ i 1 t − i H0 (t−t′ ) [Vext (t′ ), ρ0 ]e h¯ H0 (t−t ) dt′ , e h¯ ∆ρ(t) = i¯ h −∞

(2.50)

which is explicitly a linear expression in the external perturbation. Let us now calculate the thermal expectation value of the observable A (the index t indicates the possible time dependence) At = Tr (ρ0 + ∆ρ(t))A = Tr(ρ0 A) + Tr ∆ρ(t)A . (2.51)

The first term is the thermal expectation value of A of the unperturbed system (i.e., in equilibrium), which may be written A0 . The second term, depending on the perturbation (2.38), can be transformed in several steps by exploiting the meaning of the trace operation: i ′ ′ i t i Tr e− h¯ H0 (t−t ) [BF (t′ ), ρ0 ]e h¯ H0 (t−t ) A dt′ At = A0 + h −∞ ¯ i t Tr A(t − t′ ) {BF (t′ )ρ0 − ρ0 BF (t′ )} dt′ = A0 + h −∞ ¯ i t Tr ρ0 {A(t − t′ )B − BA(t − t′ )} F (t′ )dt′ . (2.52) = A0 + h −∞ ¯

In the first step, we applied a cyclic permutation of operators under the trace operation, in the second step we introduced the interaction picture for the observable A written as A(t − t′ ), and in the last step, we extracted the scalar factor F (t′ ) (note, that it is not an operator) and performed a cyclic permutation in the second term. We may now write B = B(0) because, for vanishing time argument, the Schr¨ odinger picture coincides with the interaction picture. Moreover, the upper limit of the integral can be shifted to infinity when the integrand is multiplied by the unit step function θ(t − t′ ). Thus we find i ∞ At = A0 + θ(t − t′ )[A(t − t′ ), B(0)]0 F (t′ )dt′ (2.53) h −∞ ¯

28

2 The Solid as a Many-Particle Problem

which has the form of (2.39). The expectation value . . . 0 under the integral is to be taken with the statistical operator ρ0 in the equilibrium. The usual writing is by taking the Fourier transform +∞ eiωt (At − A0 ) dt ∆A(ω) = −∞

i = h ¯

+∞

−∞

′

′

dtdt′ eiω(t−t ) θ(t − t′ )[A(t − t′ ), B(0)]0 eiωt F (t′ ).

Finally, by changing the two time integrations with the substitution t − t′ = τ we have +∞ ′ i +∞ iωτ ∆A(ω) = dt′ eiωt F (t′ ) (2.54) dτ e θ(τ )[A(τ ), B(0)]0 h −∞ ¯ −∞ and ∆A(ω) = χAB (ω)F (ω).

(2.55)

This is the linear response of the system when measuring the observable Aˆ (now using the ˆ-sign again) by applying a (time-dependent) perturbation ˆ (ω). It is expressed as the Fourier transform of the diﬀerence between −BF the thermal expectation value of Aˆ and its equilibrium value and written as a product of the perturbing field and a response function or susceptibility i +∞ ˆ ), B(0)] ˆ χAB (ω) = dτ eiωτ θ(τ )[A(τ (2.56) 0. h −∞ ¯ The response function is completely defined by the unperturbed system (in our case the solid in thermal equilibrium) and describes a material property of the solid. It is a correlation function between the measured observable Aˆ ˆ appearing in the perturbation. As will be seen later, and the observable B ˆ ˆ are Hermitian adjoints of each other. The time the observables A and B dependent integrand with the step function warrants the causality between perturbation and response. This characteristic structure will we identified later (Sect. 7.1) as that of a retarded Green function. Having anticipated a periodic time dependence of the perturbation with the frequency ω, we obtain the response depending on ω (dynamical response). The special case of ω = 0 refers to the static response.

2.5 Kubo’s Formulas: Response Functions The general concept of linear response has been developed in an earlier paper by Ryogo Kubo[67] and can be found in some text books (e.g., [9, 13, 18]). In this section, we restrict ourselves to the response functions of importance in solid state physics, some of which have already been mentioned in Table 2.1. 1. Dielectric Susceptibility: When exposing a piece of matter to an electric field, the matter will be polarized, i.e., it responds by shifting positive and

2.5 Kubo’s Formulas: Response Functions

29

negative charges against each other or by aligning existing electric dipoles which, without the field, are randomly oriented. The polarization or electric dipole density P = M E /V (here V is the volume) is given by the relation P = χE · E,

(2.57)

which has the form of (2.55). The dielectric susceptibility χE is a symmetric second rank tensor. Its principal values depend on the crystal structure of the solid as will be seen later. This tensor property allows the two vectors P and E to have diﬀerent directions and to describe birefringence. This is the case in solids with axial anisotropy, while in cubic crystals the principal values of χE are equal to each other. In order to formulate the dielectric susceptibility as a correlation function ˆ The perturbation as in (2.56), we have to identify the two observables Aˆ and B. is the potential energy of the electric dipole moment M E in the applied electric field E(t) Vext (t) = −M E · E(t) = − MjE Ej (t). (2.58) j

ˆ to be one of the components of the electric dipole We find the observable B moment and the scalar field F (t) to be the corresponding component of the applied electric field. The observable to be measured for investigating the dielectric susceptibility (or one of its tensor components) is a component of the electric dipole density, say Pi , and we may write (2.53) in the form Pi t − Pi0 =

i +∞ θ(t − t′ )[MiE (t − t′ ), MjE (0)]0 Ej (t′ )dt′ (2.59) h ¯ V −∞ j

and obtain after Fourier transformation, the response function +∞ i E χij (ω) = eiωτ θ(τ )[MiE (τ ), MjE (0)]0 dτ. hV −∞ ¯

(2.60)

This is the Kubo formula for the dielectric susceptibility, which is a correlation function of components of the electric dipole moment. The contribution of the crystal lattice to the dielectric susceptibility will be described in more detail in Sect. 3.5. We mention in passing that systems with Pi0 = 0 are called ferroelectric; their configuration of electrons and ions allows for a spontaneous dielectric polarization. 2. Magnetic Susceptibility: Magnetic properties of matter can be characterized by the response of the system to an applied magnetic field B. The measured quantity will be the magnetization M = m/V , which in complete analogy to the dielectric case is the magnetic dipole density. Making use of this analogy, we may write the perturbation Vext (t) = −m · B(t) and identify the observˆ with a able Aˆ with a component of the magnetization and the observable B

30

2 The Solid as a Many-Particle Problem

component of the magnetic dipole moment to write the Kubo formula for the magnetic susceptibility +∞ i eiωτ θ(τ )[mi (τ ), mj (0)]0 dτ. (2.61) χM (ω) = ij hV −∞ ¯ It is a correlation function between the components of the magnetic dipole moment. In Chap. 6, we shall express the magnetic dipole moment in terms of the electron spin operators S l as Sl, (2.62) m = gμB l

with the Land´e8 g factor and the Bohr9 magneton μB . For this case, the magnetic susceptibility i 2 2 +∞ iωτ M g μB e θ(τ )[Sl,i (τ ), Sl′ ,j (0)]0 dτ. (2.63) χij (ω) = hV ¯ −∞ ′ ll

turns out to be a correlation function between the components of the electron spin operator or in short, a spin–spin correlation function. 3. Dielectric Function: The dielectric function (or frequency dependent dielectric constant) is known from electrodynamics; it connects the displacement field with the electric field according to D(q, ω) = ε0 ε(q, ω)E(q, ω).

(2.64)

The space and time dependence of the quantities are specified by the wave vector q and the frequency ω, respectively. Depending on the experimental situation, one distinguishes between the response to a longitudinal or transverse perturbation. Let us consider here, the case of probing the dielectric system with an external test charge, which exerts a longitudinal electric field. This scenario is typical for a scattering experiment. It is intuitively clear that the charges of the dielectric will rearrange in the presence of the test charge, and this rearrangement is quantified by an induced charge density enind . The fields are determined by their respective charges via Poisson’s equation. For the displacement field, these are the external charges enext (q, ω) (the Fourier transform of the external charge density enext (r, t)) iq · D(q, ω) = enext (q, ω)

(2.65)

and for the electric field, in addition the induced charges (due to polarization in the presence of the external charge) enind (q, ω) 8 9

Alfred Land´e 1888–1975. Niels Bohr 1885–1962, Nobel prize in physics 1922.

2.5 Kubo’s Formulas: Response Functions

iq · E(q, ω) =

e next (q, ω) + nind (q, ω) . ε0

31

(2.66)

It is straightforward to eliminate the fields from (2.64) to (2.66) and to write ntot (q, ω) = nind (q, ω) + next (q, ω) =

next (q, ω) . ε(q, ω)

(2.67)

The last relation is the standard expression for the dielectric screening of a test charge in dielectric matter. In general, for time-dependent processes, it depends on ω, and describes dynamical screening. We may read this relation also from the viewpoint of linear response: The system responds to the external or test charge by building up the total charge density entot (q, ω) with the inverse dielectric function as the response function. In order to cast it into the form of (2.56), we identify the observable Aˆ as the charge density operator n ˆ q eiq·r (2.68) eˆ n(r) = e q

or its Fourier transform, the operator of density fluctuations n ˆ q . The time dependent induced density fluctuations are given by nind (q, t) = ˆ nq t − n0 .

(2.69)

The perturbation is caused by the charge density of the external (or test) charge ′ enext (r, t) = e next (q ′ , t)eiq ·r (2.70) q′

with the interaction energy −e2 n ˆ (r)next (r ′ , t) Vext (t) = d3 rd3 r ′ 4πε0 |r − r′ | ′ ′ ei(q·r+q ·r ) −e2 . n ˆ q next (q ′ , t) d3 rd3 r′ = 4πε0 ′ |r − r ′ |

(2.71)

qq

The double integral can be expressed in terms of the Fourier transform, and one obtains ˆ−q Next (q, t), Vext (t) = − (2.72) vq N q

where vq = e2 /V ε0 q 2 is the Fourier transform of the Coulomb potential (see ˆ−q as ˆ−q = V n ˆ −q and Next (q, t) = V next (q, t). By identifying N Appendix), N ˆ the observable B and next (q, t) as the scalar time dependent field F (t) of the general response formalism, we may now write

32

2 The Solid as a Many-Particle Problem

nind (q, ω) = vq

i h ¯

+∞

ˆ−q (0)]0 next (q, ω)dτ. eiωτ θ(τ )[Nˆq (τ ), N

(2.73)

−∞

In comparison with (2.67), we find as the Kubo formula for the inverse dielectric function i +∞ iωτ 1 ˆ−q (0)]0 dτ. e θ(τ )[Nˆq (τ ), N (2.74) = 1 + vq ε(q, ω) h −∞ ¯ It is a correlation function of the number (or density) fluctuations and will be discussed in more detail in Chap. 4 as energy–loss function. As further examples, one may consider the tensor components of the electric conductivity σµν (ω) (Problem 2.4) or the thermal conductivity mentioned in Table 2.1. The response concept can be generalized beyond the linear regime by considering higher order terms in the external perturbation (Problem 2.5). The corresponding response functions describe the nonlinear properties of the material, which become important, for instance, in the response of a solid to intense laser light (nonlinear optics) or strong electric fields (nonlinear transport) [68].

2.6 Properties of Response Functions Because response functions will be used throughout the book, some of their properties have to be compiled at the beginning [64, 69, 70]. Response functions χAB (ω) of the general form (2.56) are complex valued functions for real frequencies. If they are considered in the complex frequency plane with z = ω + iδ, they represent analytic functions in the upper half-plane. The analyticity is a consequence of the causal connection between the perturbation and its eﬀect on the system as expressed by the unit step function under the integral. Due to this property, Cauchy’s theorem holds, which reads χAB (z ′ ) ′ 1 χAB (z) = dz , (2.75) 2πi z′ − z where the contour integral is along the real axis and closes along the great semicircle in the upper half-plane. Assuming that χAB (z ′ ) vanishes suﬃciently rapidly at infinity, the semicircle does not contribute to the integral, and we have +∞ χAB (ω ′ ) ′ 1 χAB (z) = dω . (2.76) 2πi −∞ ω ′ − z This relation can be evaluated for real z with ! 1 1 =P + iπδ(ω ′ − ω), lim ′ Γ→0 ω − ω − iΓ ω′ − ω

(2.77)

where P denotes the principal part. Using (2.77) in (2.76) and rearranging terms gives

2.6 Properties of Response Functions

+∞

33

χAB (ω ′ ) ′ dω . ω′ − ω

(2.78)

χAB (ω) = ReχAB (ω) + iImχAB (ω),

(2.79)

χAB (ω) =

1 P iπ

−∞

By separating this relation into real and imaginary part,

one finds the dispersion or Kramers–Kronig10 relations

Re χAB (ω) =

1 P π

1 Im χAB (ω) = − P π

+∞

Im χAB (ω ′ ) ′ dω ω′ − ω

−∞ +∞ −∞

Re χAB (ω ′ ) ′ dω ω′ − ω

(2.80) (2.81)

which are a consequence of causality. The real and imaginary parts of the susceptibility are also called the dissipative and absorptive parts, respectively. The meaning of these relations can be illustrated by considering the imaginary part of a susceptibility of the form Imχ(ω ′ ) ∼ (δ(ω0 − ω ′ ) − δ(ω0 + ω ′ )) (as will be shown in Chap. 3, this describes the absorption (and emission) of a harmonic oscillator with eigenfrequency ω0 ). For this case, the real part of χ(ω) takes the form 1/(ω02 − ω 2 ), characteristic for the anomalous dispersion of an oscillator (Problem 2.6). According to the structure of (2.56), the susceptibility is the Fourier transform of an object of the more general form i ˆ ), B(0)] ˆ GAB (τ ) = − θ(τ )[A(τ ± h ¯

(2.82)

ˆ B ˆ are fermion which is a retarded Green11 function. Depending on whether A, ˆ ˆ ˆ +B ˆ Aˆ or boson operators, one has to take the anti-commutator [A, B]+ = AˆB ˆ ˆ ˆ ˆ ˆ ˆ or the commutator [A, B]− = AB − B A, respectively. Note, that a response function χAB is always written in terms of the commutator. The thermal expectation value is evaluated for the system Hamiltonian (in case of a response function it is the one without external perturbation). The retarded Green function GAB (τ ) can be decomposed into the corre> < ˆ )B(0) ˆ ˆ A(τ ˆ ) lation functions CAB (τ ) = A(τ and CAB (τ ) = B(0) > i < GAB (τ ) = − θ(τ ) CAB (τ ) ± CAB (τ ) . h ¯

(2.83)

Its Laplace transform, to which the Fourier transform reduces because of the step function in GAB (τ ),

10 11

Hendrik Anton Kramers 1894–1952; Ralph de Laer Kronig 1904–1995. George Green 1793–1841.

34

2 The Solid as a Many-Particle Problem

∞

eizτ GAB (τ )dτ +∞ dω −iωτ > i ∞ izτ < e dτ e CAB (ω) ± CAB (ω) (2.84) =− h 0 ¯ −∞ 2π

GAB (z) =

0

< > can be expressed in terms of the Fourier transforms CAB (ω) of (ω) and CAB the correlation functions. After performing the time integration, one finds the spectral representation of the Green function 1 +∞ ρAB (ω ′ ) ′ GAB (z) = dω (2.85) π −∞ ω ′ − z

with the spectral function

ρAB (ω) = −

1 > < CAB (ω) ± CAB (ω) . 2¯h

(2.86)

The upper(lower) sign refers to expressions formed with fermion(boson) operators. The spectral representation (2.85) can be used to express ρAB (ω) in terms of the Green function ρAB (ω) =

1 lim GAB (ω + iΓ) − GAB (ω − iΓ) 2i Γ→0

(2.87)

which, under the conditions given below, is real valued. Some properties of the spectral function ρAB (ω) can be obtained by writing the time-dependent correlation functions in the energy or spectral representation (using eigenstates |n, |m of the system Hamiltonian H) i > ˆ − h¯i Hτ B ˆ CAB (τ ) = e h¯ Hτ Ae i = pn Anm Bmn e h¯ (En −Em )τ .

(2.88)

nm

Here pn is the statistical factor, En , Em are the eigenvalues of H and ˆ The Fourier transform of C > (τ ) Anm , Bmn the matrix elements of Aˆ and B. AB is given by > (ω) = pn Anm Bmn 2πδ(¯hω + En − Em ). (2.89) CAB nm

Similarly we find < (ω) = CAB

nm

pm Bmn Anm 2πδ(¯hω + En − Em ).

(2.90)

The following relations hold for the correlation functions: > < CAB (ω) = CBA (−ω) < −β¯ hω > CAB (ω). CAB (ω) = e

(2.91) (2.92)

2.6 Properties of Response Functions

35

The first of these relations is obtained by interchanging n and m under the sum in (2.89) and (2.90). The second is found by using the ratio pm /pn = exp (−β¯ hω) of the statistical factors. (This relation holds not only for the canonical ensemble, but also for the grand-canonical ensemble if the operators ˆ do not change the particle numbers.) Because of (2.91), the spectral Aˆ and B function has the property ρAB (−ω) = ± ρBA (ω),

(2.93)

which in connection with (2.85), implies GAB (−z) = ∓ GBA (z), where again, the upper(lower) sign refers to the quantity defined for fermion(boson) operˆ = Aˆ† , and using (2.89), (2.90), and (2.92) we may write ρBA (ω) ators. For B in the form π ρAA† = − 1 ± e−β¯hω pn Anm Amn δ(¯hω + En − Em ), (2.94) h ¯ m,n

to see with Amn = A∗nm that it is real, and with (2.87) that ρAA† (ω) = ˆ = A, ˆ one Im{GAA† (ω)}. In the special case of Hermitian operators, B has Im{GAA (−ω)} = −Im{GAA (ω)}. Using (2.92), the correlation function > CAB (ω) can be expressed also as > CAB (ω) = −2¯h

and its Fourier transform is ¯ h A(τ )B(0) = − π

+∞

−∞

ρAB (ω) 1 ± e−β¯hω e−iωτ

ρAB (ω) dω. 1 ± e−β¯hω

(2.95)

(2.96)

Let us finally specialize to the case of GAB being a response function with ˆ = Aˆ† and consider the case τ = 0. Then with ρAA† (ω) = −Im χ ˆ ˆ† (ω) the B AA last relation takes the form h +∞ ImχAˆAˆ† (ω) ¯ AˆAˆ† = dω, (2.97) π −∞ 1 − e−β¯hω which connects the imaginary (or dissipative) part of the response function with a quantity that describes the fluctuations in the observable Aˆ and is therefore known as the dissipation–fluctuation theorem. ˆ and the appearance Due to the hermiticity of the operators Aˆ and B of their commutator in the response function, its imaginary part fulfills the relations Im χAB (−τ ) = −Im χBA (τ )

(2.98)

Im χAB (−ω) = −Im χBA (ω)

(2.99)

χ∗AB (ω) = χAB (−ω),

(2.100)

and with the response ∆A(ω) in (2.55) is real as required for an observable.

36

2 The Solid as a Many-Particle Problem

Problems 2.1 Formulate the matrix representation of the statistical operator with the eigenstates of the system Hamiltonian and show that it is diagonal. 2.2 Show that the thermal expectation value of an observable reduces for T = 0 K to the expectation value of this observable in the ground state of the system. 2.3 Calculate the thermal expectation value of the particle number operator for a system of (free) fermions in thermal equilibrium to find the Fermi–Dirac distribution. Do the same for a system of bosons to obtain the Bose–Einstein distribution. Consider the high temperature limit. 2.4 In order to derive the Kubo formula for the electric conductivity σµν (ω), ˆ in (2.56). Keep in mind which quanidentify the observables Aˆ and B tity is measured in transport experiments, and how the coupling between electrons and the electromagnetic field is described. 2.5 Develop the concept of nonlinear response by considering (as in perturbation theory) corrections to the equilibrium distribution ∆ρ = ∆ρ1 + ∆ρ2 + . . . of increasing order in the external perturbation. Find the structure of the lowest order nonlinear response function for the observable Aˆ due to the perturbation Vext (2.38). 2.6 Let the imaginary part of a response function be of the form ImχAB (ω) = χ0 δ(ω0 − ω) − δ(ω0 + ω) . Calculate the real part of the response function using (2.80).

3 Lattice Dynamics: Phonons

The motion of the heavy constituents of the solid and the solid state properties, which are essentially determined by this motion, are the subject of this chapter. Lattice dynamics is a standard topic in textbooks on Solid State Physics, but there is also a variety of monographs on this subject, e.g., [71–83], of which the one by Born and Huang [71], entitled “Dynamical Theory of Crystal Lattices” is the pioneering textbook in this field. The progress in Lattice Dynamics, especially the strong mutual influence between experimental and theoretical investigations, is well documented in the three volumes “Phonons: Theory and Experiment,” edited by Br¨ uesch [77–79]. In simplest terms, the dynamics of a lattice can be described by employing a classical model with massive spheres (representing the atoms or ions) connected by springs (representing the chemical bonds) in a periodic array (see Fig. 3.1). For a more rigorous description, the starting point is the adiabatic approximation of Sect. 2.2, according to which the motion of the ions is governed by the time-dependent Schr¨ odinger equation (see 2.23) (Hion + Eel,α )Φα = i¯ hΦ˙ α ,

(3.1)

where (see 2.24) Hion + Eel,α =

P2 1 n + Eel,α ({Rn }) + V (Rn − Rm ) . 2M 2 n,m n

(3.2)

n=m

The last two terms on the rhs represent the adiabatic potential energy Uα ({Rn }) of the ion configuration {Rn }, with the electron system being in the state described by the wave function Ψα ({r l }, {Rn }) as introduced in Sect. 2.2. Thus, the motion of the ions, described by the wave function Φα ({Rn }), is not only determined by the potential energy of the ion configuration but also by the energy of the electrons Eel,α ({Rn }) in this configuration. Instead of considering the dependence of this motion on α, let us assume, U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 3,

38

3 Lattice Dynamics: Phonons

Fig. 3.1. Classical model for lattice dynamics: spheres and springs

un(t)

Rn0

M

Rn

Fig. 3.2. Equilibrium position and displacement of an ion

for simplicity, the electron system to be in its ground state (see Chaps. 4 and 5) and drop the index α. Hence, in this chapter, we are dealing with the Hamiltonian P2 nτ H = Hion + Eel = + U ({Rnτ }) (3.3) 2M τ nτ

to describe the dynamics of the crystal lattice, which, in general, can be a lattice with basis as indicated by the vector index τ . In a crystalline solid, the ion configuration is characterized by the actual positions of the atoms Rnτ (t) = R0nτ + unτ (t),

(3.4)

with equilibrium positions R0nτ at the points of a lattice with basis (see Sect. 1.2) and time-dependent displacements unτ (t) around these positions (see Fig. 3.2). The vector indices n = (n1 , n2 , n3 ) for the lattice cell and τ for the atoms of the basis have been introduced in Sect. 1.2. The aim of lattice dynamics is to set up and solve equations of motion for the displacement vectors unτ (t).

3.1 Harmonic Approximation

39

3.1 Harmonic Approximation It is reasonable and required for the Born–Oppenheimer approximation to start with displacements |unτ (t)| which are small compared to the lattice spacings |aj |, i.e., the springs are only slightly distorted and the ions remain close to their equilibrium positions. (The other extreme, when |unτ (t)| becomes comparable with |aj |, would lead to a destruction of the lattice, e.g., in the case of melting, which is not the subject here.) Under this condition the adiabatic potential can be expanded in a power series with respect to the components of unτ (t) around the equilibrium positions. Let us adopt here a frequently used notation [84] ⎛ ⎞ n ← lattice cell (3.5) unτ i =: u ⎝ τ ⎠ ← atom of basis i ← vector component to write the expansion

⎞ ⎛ ⎞ n n Φ⎝τ ⎠u⎝τ ⎠ U({Rnτ }) = U({R0nτ }) + nτ i i i ⎛ ⎞ ⎛ ⎞ ⎛ ⎞ m nm n 1 ⎝ + Φ τ τ ′ ⎠ u ⎝ τ ⎠ u ⎝ τ ′ ⎠ + O(u3 ), 2 nτ i i j i j mτ ′ j

⎛

(3.6)

with ⎞ ⎛ ⎞ " " nm n " " ∂U ∂2U ′⎠ " " ⎠ ⎝ ⎝ and Φ τ τ = Φ τ = " " 0 . ′j ∂u ∂u ∂u 0 } nτ i nτ i mτ {R {Rnτ } i i j nτ ⎛

(3.7)

For the equilibrium configuration, the potential energy U({Rnτ }) is at its minimum and all first derivatives vanish; those of second order become the leading terms and are the only ones if higher order terms are neglected because of their smallness. This defines the harmonic approximation with the Hamiltonian ⎛ ⎞ ⎛ ⎞ ⎛ ⎞ m nm n P2 1 nτ Φ ⎝ τ τ′ ⎠ u ⎝ τ ⎠u ⎝ τ′ ⎠ + H= (3.8) 2Mτ 2 nτ i nτ i j i j mτ ′ j

formulated in terms of the displacements unτ and their conjugate momenta P nτ . It corresponds to that of a set of coupled harmonic oscillators for which the canonical equations of motion ∂H P˙nτ i = − , ∂unτ i

u˙ nτ i =

∂H Pnτ i = ∂Pnτ i Mτ

(3.9)

40

3 Lattice Dynamics: Phonons

or Newton’s equations of motion ⎛ ⎛ ⎞ ⎞ ⎛ ⎞ nm n m Mτ u Φ ⎝ τ τ′ ⎠u ⎝ τ′ ⎠ ¨⎝τ ⎠ = − i j i j mτ ′ j

(3.10)

are easily obtained. These equations generalize the simple pendulum model, a mass attached to a spring, to a network like that of Fig. 3.1: The rhs of (3.10) represents the force exerted onto the mass Mτ at Rnτ in direction i if the masses at Rmτ ′ are moved in directions j. The Φ are the force constants symbolized in Fig. 3.1 by the springs, but their microscopic meaning derives from the adiabatic potential by means of (3.7). Simple examples are the subject of Problem 3.1 (the linear chain with two diﬀerent atoms in the unit cell) and Problem 3.2 (the two-dimensional square lattice). Supplement: Reducing the number of force constants The number of force constants, which is the squared number of degrees of freedom, can drastically be reduced in the following way (see also Problem 3.2): 1. According to Newton’s third axiom (action = reaction), we can write ⎛ ⎛ ⎞ ⎞ nm mn ′ ′ Φ⎝ τ τ ⎠ = Φ⎝ τ τ ⎠. i j j i

(3.11)

2. If the displacements um τ ′ are the same for all mτ ′ , the solid experiences a translation by u, which (in the absence of external forces) does not change the potential energy. Formally, this means ⎞ ⎛ nm ′ (3.12) Φ ⎝ τ τ ⎠ uj = 0 for arbitrary uj , j = 1, 2, 3 mτ ′ j i j or

⎞ nm Φ⎝ τ τ′ ⎠ = 0 mτ ′ i j

⎛

for

i, j = 1, 2, 3.

(3.13)

3. Due to the invariance of the lattice under discrete translations or by replacing R0n → R0n + R 0n ′ , we have ⎛ ⎞ ⎛ ⎞ ⎛ ⎞ nm n−m 0 n + n′ m + n′ τ′ ⎠ = Φ⎝ τ Φ⎝ τ τ′ ⎠ = Φ⎝ τ (3.14) τ′ ⎠ , i j i j i j

where the last equation follows for the choice n′ = −m. 4. Making use of the symmetry of the lattice under operations of the point group (rotations, mirror reflections, inversion), the number of independent force constants is further reduced depending on the actual lattice structure. 5. Finally, the forces reduce with increasing distance between the masses and usually only those between nearest and next nearest neighbors need to be considered (see Problems 3.1 and 3.2).

3.1 Harmonic Approximation

41

The equations of motion (3.10), being a homogeneous set of coupled linear diﬀerential equations, can be solved by the standard procedure assuming ⎛ ⎞ n 1 u ¯ ⎝ τ ⎠ e−iωt , (3.15) unτ i (t) = √ Mτ i where, by convention, a factor with the square-root of the mass is extracted. It yields a homogeneous system of coupled linear equations for the displacements ⎛ ⎛ ⎞ ⎞ ⎛ ⎞ nm n m 1 ⎝ τ τ′ ⎠u ⎝ τ′ ⎠. # Φ ¯⎝τ ⎠ = (3.16) ¯ ω2u Mτ Mτ′ i j i j mτ ′ j

We make use of the Bloch theorem (see Problem 1.4), according to which the displacements on diﬀerent lattice sites diﬀer only by phase factors ⎛ ⎞ ⎞ ⎛ ⎞ ⎛ n 0 n+m 0 0 0 (3.17) u ¯ ⎝ τ ⎠ = u¯ ⎝ τ ⎠ eiq·Rm = u¯ ⎝ τ ⎠ eiq·(Rm +Rn ) , i i i and apply the periodic boundary conditions (Problem 3.3), which restrict the components of the wave vector to the discrete values qi =

2π νi , N i ai

νi = 0, . . . , Ni − 1,

i = 1, 2, 3.

(3.18)

Here N = N1 N2 N3 ≫ 1 is the number of unit cells in the periodicity (or crystal) volume V . Thus, (3.10) can be written in the compact form ¯τ i (q) = Dτ i,τ ′ j (q)¯ uτ ′ j (q), (3.19) ω2u τ ′j

with the dynamical matrix Dτ i,τ ′ j (q) =

m

⎛

⎞ n−m 0 0 0 √ τ ′ ⎠ eiq·(Rm −Rn ) . Φ⎝ τ Mτ Mτ ′ i j 1

(3.20)

It does not depend on the index n and on the sign of q (as can be seen by rearranging the sum over m), thus Dτ i,τ ′ j (q) = Dτ i,τ ′ j (−q). The number of solutions of (3.19) is the same as the number 3rN of degrees of freedom of the system, where r is the number of atoms in the basis and N the number of lattice cells (equal to the number of diﬀerent q in the first Brillouin zone). For each q, the secular problem $ $ $Dτ i,τ ′ j (q) − ω 2 δτ τ ′ δij $ = 0 (3.21)

42

3 Lattice Dynamics: Phonons

yields 3r eigenfrequencies ωs (q) = ωs (−q), s = 1, . . . , 3r, with corresponding normalized eigenvectors esτ (q). It is important to note that the solutions of (3.19) describe collective modes or excitations , for which all ions of the lattice move with the same time dependence but phase shifted with respect to each other according to Bloch’s theorem. For the collective mode sq, the motion of the individual ion (or mass) at Rnτ is described by the displacement 0 1 esτ (q) ei(q·Rn −ωs (q)t) . usnτ (q, t) ∼ √ Mτ

(3.22)

As usual, the eigenmodes form a complete set of solutions that can be used as the basis for representing an arbitrary motion of the lattice or of the individual ion. Essential aspects of lattice dynamics in the harmonic approximation, such as the dependence of the frequencies on the force constants and masses, the distinction between acoustic and optical branches (see Sects. 3.4 and 3.5), and the anisotropy of the dispersion curves, can be studied already in simplified models as those of Problems 3.1 and 3.2. Moreover, even the reduced dimension assumed in these problems is not hypothetical but corresponds to physical reality: Take the atoms at the surfaces of solids, they move diﬀerently from those in the bulk and give rise to investigate surface phonons [85–87]. They represent a (quasi) two-dimensional dynamical system with phonon amplitudes that decay over a few lattice constants away from the surface.

3.2 Normal Coordinates Having found the eigensolutions of the Hamiltonian (3.8) in the previous section, we now aim at formulating this Hamiltonian in terms of these eigensolutions or normal coordinates. We expect that in this representation the Hamiltonian will be that of a set of uncoupled harmonic oscillators, each corresponding to a collective mode. The displacement of an ion (or atom) can be expressed in terms of the complete set of eigensolutions (3.22) 0 1 fs (q) e−iωs (q)t esτ i (q) eiq·Rn , unτ i (t) = √ N Mτ sq

(3.23)

with expansion coeﬃcients fs (q). The normal coordinate for the collective lattice mode sq is defined by Qs (q, t) = fs (q) e−iωs (q)t .

(3.24)

The following scheme demonstrates the intended transformation from a system with coupled localized motions of the individual masses (or ions) around their equilibrium position to the uncoupled delocalized collective motions with all ions moving in phase:

3.2 Normal Coordinates

coupled motion localized unτ i (t) Dτ i,τ ′ j (q) H(unτ , P nτ )

⇒ → → →

43

uncoupled collective motion delocalized Qs (q, t) diagonal form with ωs2 (q) H (Qs (q), Ps (q)) .

Starting from the original Hamiltonian H(unτ , P nτ ) (3.8) and making use of (3.23) and % 0 Mτ ˙ (3.25) Qs (q, t)esτ i (q) eiq·Rn , Pnτ i = Mτ u˙ nτ i = N sq we find the Lagrangian corresponding to H (Qs (q), Ps (q)), which can be written as ′ 0 ′ 1 ˙ L= Qs (q)Q˙ s′ (q ′ ) ei(q+q )·Rn esτ i (q)esτ i (q ′ ) 2N nτ i sq s′ q ′

−

nτ i sq τ ′ j s′ q ′

′ ′ 0 Dτ i,τ ′ j (q ′ )Qs (q)Qs′ (q ′ )esτ i (q)esτ ′ j (q ′ ) ei(q+q )·Rn ,

(3.26) where, in the second term, the force constants are expressed by the dynamical matrix (3.20). The sum over the lattice points n can be performed with 1 i(q+q′ )·R0n e = δq′ ,−q+G , N n

(3.27)

G

where G is a vector of the reciprocal lattice. On the rhs, only the term with G = 0 contributes, because q and q ′ are vectors in the first Brillouin zone. The displacements unτ are real and, therefore, ∗ esτ i (−q)Qs (−q, t) = es∗ τ i (q)Qs (q, t).

(3.28)

As this relation holds for arbitrary Qs and all esτ i , we have esτ i (−q) = es∗ τ i (q)

and Qs (−q, t) = Q∗s (q, t),

(3.29)

and can write the Lagrangian as ′ 1 & ˙ Qs (q)Q˙ ∗s′ (q)esτ i (q)eτs i∗ (q) L= 2 τ i ss′ q

− Qs (q)Q∗s′ (q)esτ i

τ ′j

' ′ Dτ i,τ ′ j (−q)esτ ′∗j (q) .

(3.30) ′

The sum in the last term on the rhs of (3.30) simplifies, because esτ ′∗ (q) = ′ esτ ′ (−q) is an eigenvector of the dynamical matrix (see (3.19)). The

44

3 Lattice Dynamics: Phonons

Hamiltonian (3.8) corresponding to L, formulated in terms of the normal coordinates Qs (q, t) and their conjugate momenta Ps (q, t) =

∂L = Q˙ ∗s (q, t) ∂ Q˙ s (q, t)

(3.31)

takes the form H (Qs (q), Ps (q)) =

) 1 ( ∗ Ps (q)Ps (q) + ωs2 (q)Q∗s (q)Qs (q) . 2 sq

(3.32)

As expected, it describes 3rN uncoupled harmonic oscillators, each of which corresponds to a collective mode or elementary excitation of the lattice (characterized by s and q). For the individual mode, the equation of motion follows from (3.32) P˙s∗ (q, t) = −

∂H ∂Q∗s (q, t)

or

¨ s (q, t) = −ω 2 (q)Qs (q, t), Q s

(3.33)

and is easily identified as that of a harmonic oscillator. Note, that the decoupling of the ion motions is possible only in the harmonic approximation. If in the expansion of the adiabatic potential (3.6) higher order terms in the displacements are taken into account, one arrives at nonlinear lattice dynamics, which has to be invoked to describe thermal lattice expansion or the temperature dependence of the frequencies ωs (q) [84]. We come back to these properties in Sect. 3.7.

3.3 Phonons and Occupation Number Representation The lattice dynamics, so far formulated in terms of classical mechanics, can be cast into the language of quantum mechanics by using Q∗s (q) = Qs (−q), Ps∗ (q) = Ps (−q), and converting the conjugate variables Qs (q) and ˆ s (q) and Pˆs (q), for which we postulate Ps (q) of the system into operators Q the commutation relations ˆ s (q), Pˆs′ (q ′ )] = i¯ [Q hδss′ δqq′

(3.34)

ˆ s (q), Q ˆ s′ (q ′ )] = [Pˆs (q), Pˆs′ (q ′ )] = 0. [Q (3.35) ˆ s (q), and Pˆs (q) it is more convenient to introduce for Instead of using Q each mode s, q creation and annihilation operators known from the harmonic oscillator of elementary quantum mechanics −1/2 ˆ s (−q) − iPˆs (q) a†s (q) = 2¯hωs (q) ωs (q)Q (3.36) −1/2 ˆ s (q) + iPˆs (−q) . ωs (q)Q as (q) = 2¯hωs (q) (3.37)

3.3 Phonons and Occupation Number Representation

45

They obey the commutation relations (Problem 3.4) [as (q), a†s′ (q ′ )] = δss′ δqq′ , [as (q), as′ (q ′ )] = [a†s (q), a†s′ (q ′ )] = 0.

(3.38)

With the inverted relations of (3.36) and (3.37), Qs (q) = Ps (q) = i

¯ h 2ωs (q)

!1/2

hωs (q) ¯ 2

!1/2

a†s (−q) + as (q)

a†s (q) − as (−q) ,

(3.39) (3.40)

and with ωs (q) = ωs (−q), the Hamiltonian of uncoupled harmonic oscillators (3.32) can be written as ! 1 † ˆ . (3.41) hωs (q) as (q)as (q) + ¯ H= 2 sq Here, ¯hωs (q) is the energy quantum of the collective lattice mode. This excitation out of the ground state is called phonon and h ¯ ωs (q) is the phonon energy. Let us contemplate briefly on how we arrived at this Hamiltonian: We started from a classical formulation of the field of lattice displacements in terms of the normal coordinates, imposed the quantization conditions (3.34, 3.35), introduced the creation and annihilation operators (3.36) and (3.37), and achieved a representation in terms of phonons as the quanta of the moving lattice. This procedure is an example of the general concept of field quantization [10]. As a reminder of quantum mechanics, we briefly present the properties of annihilation and creation operators for a harmonic oscillator: If |Ψ is an ˆ (3.41) with energy E, then a† (q)|Ψ (or as (q)|Ψ) is also an eigenstate of H s eigenstate but with energy E + h ¯ ωs (q) (or E − ¯hωs (q)), i.e., with the energy increased (decreased) by one quantum. In other words, applying a†s (q) (or ˆ creates a new eigenstate with one phonon added as (q)) to an eigenstate of H (removed). As the spectrum of the harmonic oscillator is positive definite, there must be a state |Ψ0 with the lowest non-negative energy, such that as (q)|Ψ0 = 0, indicating that there is no phonon in this state. |Ψ0 = |{0} is, therefore, called the ground state of the lattice or phonon vacuum. Its energy 1 E0 = hωs (q) ¯ (3.42) 2 sq

is that of the zero point motion of the ions about their equilibrium positions, due to the uncertainty principle. The notation {0} is a short writing for entries 0 for all quantum numbers s, q : |{0} = |0 . . . .0, all s, q. Excited states can be generated from the phonon vacuum by applying creation operators that turn the zeros into numbers ns (q) according to the number of a†s (q) applied

46

3 Lattice Dynamics: Phonons

to |{0}. Because of the commutation relations (3.38) (typical for boson operators), the result is independent of the order of the applied operators. Thus, ˆ have the form |{ns (q)}, all s, q indicating the number the eigenstates of H of phonons present in the state: this is the occupation number representation. In general, we can write these states as * † ns (q) 1 # |{ns (q)} = |{0}, (3.43) as (q) ns (q)! sq with a factor introduced for normalization. Their energy eigenvalues are E({ns (q)}) = E0 + ns (q)¯hωs (q). (3.44) sq

By comparing with the Hamiltonian (3.41), ns (q) is identified as the eigenvalue of the occupation number operator n ˆ s (q) = a†s (q)as (q). In general, a solid cannot be prepared in a certain eigenstate of the Hamiltonian (3.41), but it can be prepared in a thermodynamic state defined by the temperature. The energy of this state will be given by the thermal expectation value ! 1 ˆ H = , (3.45) hωs (q) ˆ ¯ ns (q) + 2 sq which, up to the energy of the zero point motion, is the thermal expectation value of the occupation number operator n ˆ α = a†α aα , α = s, q. As outlined in Sect. 2.3, it is given by 1 ˆ ˆ (3.46) ˆ nα = Tr n ˆ α e−β(H−µN ) , ZG ˆ = with β = 1/kB T, N ˆ α , the grand-canonical partition function ZG , αn and the chemical potential μ, which for massless particles (as for phonons) is ˆ = E0 + ′ n zero. We may write H hωα′ and obtain in occupation number ˆ α′ ¯ α representation ˆ nα =

∞ e−βE0 nα e−β α′ nα′ h¯ ωα′ . ZG

(3.47)

N =0 {nα′ }N

Here {nα′ }N stands for sets of occupation numbers nα′ with α′ nα′ = N . Instead of summing over N and sets of occupation numbers with total occupancy N , we can sum over all sets of occupation numbers without restriction. Moreover, the sum over the occupation number nα for the mode α may be separated as a factor to yield ˆ nα =

∞ e−βE0 e−β α′ =α nα′ h¯ ωα′ . nα e−β α nα h¯ ωα ZG n =0 {n } α

α′ α′ =α

(3.48)

3.3 Phonons and Occupation Number Representation

Given the grand-canonical partition function e−β α′ nα′ h¯ ωα′ , ZG = e−βE0

47

(3.49)

{nα′ }

this reduces to nxn n e−βn¯hωα ˆ nα = n −βn¯hωα = n n , ne nx

(3.50)

where x = e−β¯hωα , which can be evaluated with ˆ nα =

x(1 − x) x(1 + 2x + 3x2 + . . .) x = . = 1 + x + x2 + . . . (1 − x)2 1−x

(3.51)

The final result for the thermal expectation value ˆ ns (q) of the number of phonons with s, q written as ns (q, T ) =

1 eβ¯hωs (q)

−1

(3.52)

is the Bose–Einstein distribution function. Here, the dependence of the phonon occupancy on the temperature (remember: β = 1/kB T ) is explicitly considered in the notation. More complex thermal expectation values of phonon operators like those of Problem 3.5 can be calculated and will be needed later. A quantity frequently used to characterize a phonon spectrum and to interpret experimental data is the density of states (or phonon density of states, to distinguish from the electron density of states that will be introduced in Chaps. 4 and 5), defined by δ(ω − ωs (q)). (3.53) D(ω) = sq

It counts the number of phonon modes (or states) at a frequency ω by summing over all branches in the first Brillouin zone. Instead of summing over the discrete q values (see (3.18)), we may assign to each of them the volume (2π)3 /V in q space (see Problem 3.3), to write D(ω) as an integral over the Brillouin zone (here V is the crystal volume) V D(ω) = δ(ω − ωs (q)) d3 q. (3.54) (2π)3 s BZ The three-dimensional integration over the δ function reduces to a twodimensional integral over surfaces Sq in q space, for which ω = ωs (q) and we may write V dSq D(ω) = . (3.55) 3 (2π) s Sq |∇q ωs (q)|

48

3 Lattice Dynamics: Phonons

The integrand increases with the flatness of the dispersion curve and we expect singular behavior if |∇q ωs (q)| = 0, which defines a van Hove singularity [88] and the corresponding q as critical point (for a classification of the critical points see [21, 89]). An example of D(ω) is shown in Fig. 3.13 for GaAs. It is clearly seen that the critical points cause pronounced structures in the density of states. Especially the rather flat dispersion curves of the optical phonons result in characteristic peaks.

3.4 Acoustic Phonons Common to the phonon dispersions of all solids is the group of lowest branches with frequencies starting from the center of the Brillouin zone, with linear dependence on the wave vector (see the solutions of Problems 3.1 and 3.2). This particular property being connected with sound propagation, as will be discussed later in this section, has led to the term acoustic phonons. The other important aspect of these phonons with the smallest quanta of energy is that they dominate the low-temperature specific heat of the crystal lattice. If the solid is heated, starting from the ground state at T = 0 K, the macroscopic change of the temperature is connected microscopically with the creation, first of all, of acoustic phonons. This addition of energy in a quantized form is responsible for a peculiar behavior of the specific heat at low temperatures to which we turn our attention at the beginning of this section. The specific heat is the change of the thermal energy E(T ) with the temperature T . The thermal energy of the crystal lattice, connected with the thermal motion of the ions (or atoms), can be calculated as the thermal ˆ (3.41) expectation value of the Hamiltonian H ns (q, T )¯hωs (q). (3.56) E(T ) = E0 + sq

Let us first consider the classical limit of this general expression, which is valid for suﬃciently large temperatures h ¯ ωs (q) ≪ kB T , i.e., for a phonon energy much smaller than the average thermal energy per degree-of-freedom. In this case, the phonon occupation simplifies according to −1 kB T ns (q, T ) = eβ¯hωs (q) − 1 . ≃ ¯hωs (q)

(3.57)

As we can also neglect the ground state energy E0 (as compared to kB T ), the thermal energy of the lattice is given by E(T ) ≃ kB T = 3rN kB T. (3.58) sq

Taking the derivative with respect to T (at constant volume)

3.4 Acoustic Phonons

" dE(T ) "" cV = dT "

= 3rN kB ,

49

(3.59)

V =const

we find the classical result of the Dulong–Petit law, according to which at high temperatures each degree of freedom contributes kB to the specific heat. When approaching the low temperature regime, the phonon energy h ¯ ωs (q) is not much smaller than kB T , and its discreteness has to be taken into account in evaluating the thermal energy. This can be done within what is known as the Debye1 model. It is based on the assumption that at low temperatures only acoustic phonons are excited and that their dispersion is isotropic and follows ωs (q) = vq with the same sound velocity v for all three acoustic branches. Under these assumptions the sum over q in (3.56) can be evaluated as an integral hωs (q) ¯ V E(T ) = E0 + d3 q. (3.60) (2π)3 s eβ¯hωs (q) − 1 The limitation of the sum over q to values within the first Brillouin zone has to be considered in the integral over q by a cut-oﬀ radius qD , thus the Brillouin zone is replaced by a sphere of the same size, containing N states qD V N= q 2 dq, (3.61) 4π (2π)3 0 which yields qD = (6π 2 N/V )1/3 . This cut-oﬀ in q space can be converted into a cut-oﬀ frequency, the Debye frequency ωD = v 6π

2N

V

!1/3

.

(3.62)

The energy quantum h ¯ ωD is frequently expressed as a temperature ΘD = ¯ ωD /kB , the Debye temperature. A collection of Debye temperatures is given h in Table 3.1. Note that ωD and ΘD are proportional to the sound velocity v. For C(diamond), it is with 10,000–20,000 m s−1 (depending on the acoustic branch and the direction of propagation), the highest known for solids, while for Ne, which can be solidified only at low temperatures, it is significantly smaller. Within the solids consisting of group-IV elements (C, Si, Ge, all in the diamond structure) the Debye temperature decreases with increasing mass Table 3.1. Debye temperatures in K of some solids

ΘD 1

C

Si

Ge

Al

Cu

Ne

1,860

625

360

394

315

63

Peter Joseph Debye 1884–1966, Nobel prize in chemistry 1936

50

3 Lattice Dynamics: Phonons

of the ions. The fcc metals Al and Cu, in spite of their diﬀerent ion masses, have quite similar Debye temperatures. Substituting q = ω/v, the thermal energy of the lattice (3.60) can be written as ωD ¯hω 3 3V dω, (3.63) E(T ) = E0 + 2 3 β¯ h 2π v 0 e ω − 1 where we may identify D(ω) = V ω 2 /2π 2 v 3 with the density of states for each branch of the Debye model. It is a smooth function and does not exhibit critical points, because of the assumed linear dispersion and the neglect of Bragg reflection at Brillouin zone boundaries. By introducing the dimensionless variable x = β¯ hω (xD = β¯ hωD ), one has xD x3 3 dx. (3.64) E(T ) = E0 + 3N kB T 3 x xD 0 e − 1 The integral belongs to a class of similar expressions with diﬀerent powers of x (see Appendix A.3). For T → 0 (and xD → ∞) it approaches π 4 /15, giving for the thermal energy E(T ) → E0 + 3N kB T

π4 5

T ΘD

!3

.

(3.65)

Thus, in the low-temperature limit the specific heat exhibits a characteristic T 3 dependence " !3 dE(T ) "" 4π 4 T (3.66) = 3N k cV (T ) = B dT "V =const 5 ΘD

which is known as the Debye law. The observed vanishing of the lattice specific heat at low temperatures had puzzled physicists for quite some time. In 1905, Einstein came up with a first explanation based on the assumption of dispersionless oscillators, as in the derivation of the law for black-body radiation by Max Planck2 in 1900, but did not yield the experimental T 3 -law. Therefore, the correct derivation of this law within the Debye model in 1912 was the breakthrough in demonstrating the quantum nature of lattice vibrations. In Chap. 4, the specific heat of free electrons in metals will be discussed, which adds a contribution depending linearly on T at low temperature. Another eﬀect of the thermal motion of the lattice can be detected in scattering experiments as a reduction of the scattering amplitude. It is described by the Debye–Waller factor (see Sect. 3.7), which is calculated in Problem 3.6. Let us now turn to the elastic properties of solids by studying acoustic phonons in the long-wavelength limit q → 0, for which all atoms in a unit cell 2

Max Karl Ernst Ludwig Planck 1858–1947, Noble prize in physics 1918

3.4 Acoustic Phonons

51

u(r,t) a

r

λ/2 Fig. 3.3. Transverse ion displacements of an acoustic mode in the long-wavelength limit

move with the same phase. In this case, the equation of motion (3.10) can be written in the form ! ! ! m n nm , (3.67) =− u Mu ¨ Φ j i i j mj

where M = τ Mτ . If the lattice moves in an acoustic mode, the displacements at nearby lattice points diﬀer only little from each other, thus, making up a wave like distortion with a wavelength λ much larger than the lattice constant a. The situation is depicted in Fig. 3.3 for a transverse mode of the linear chain. The limit q → 0, called the continuum limit, allows us to consider instead of the displacements of the discrete ions, a continuous displacement field u(r, t) with the aim to derive the equation of motion. In view of this continuous displacement field, let us write ! n = ui (R0n , t) u (3.68) i and expand the displacement uj (R0m , t) at a nearby lattice point R0m around R0n " ∂uj " " R0 uj (R0m , t) = uj (R0n , t) + ∂rk "R0 mk n k " 2 " ∂ uj " 1 + R0 R0 + . . . (3.69) 2 ∂rk ∂rl "R0 mk ml kl

n

with derivatives of the continuous function uj (r, t), with respect to the components of r. When being used in the equation of motion (3.67), the first and second term of this expansion do not contribute because of the properties of the force constants discussed in Sect. 3.1, and we find " ! ! 1 ∂ 2 uj "" n nm 0 0 =− Rmk Rml Mu ¨ Φ + ... (3.70) i i j 2 ∂rk ∂rl "R0 m jkl

n

still expressed in terms of the discrete lattice points. The continuum limit is now completed by replacing the discrete masses M with the mass density

52

3 Lattice Dynamics: Phonons

per unit (or Wigner–Seitz) cell, ρM = M/Vcell and the displacements at the discrete lattice points by the continuous displacement field ui (r, t): ρM u ¨i =

cijkl

jkl

∂ 2 uj . ∂rk ∂rl

(3.71)

!

(3.72)

Here, the coeﬃcients cijkl = −

1 Φ 2Vcell m

nm i j

R0mk R0ml

are the elastic or stiﬀness constants. Equation (3.71), which is the wave equation for the elastic continuum, can be simplified by making the reasonable assumption that the force constants derived from the adiabatic potential U({Rn }) are due to central forces between the ions, thus, ! nm Φ = g(|Rm − Rn |)R0mi R0mj (3.73) i j and cijkl = −

1

Vcell

m

g(|Rm − Rn |)R0mi R0mj R0mk R0ml .

(3.74)

In this form, i.e., due to the assumed central forces, the components of the elasticity tensor are invariant under exchange of any two of the indices, and the rhs of (3.71) can be rewritten as

cijkl

jkl

∂ 2 uj ∂ǫjl = , cikjl ∂rk ∂rl ∂rk j

(3.75)

kl

where we have introduced the components of the strain tensor ! 1 ∂uj ∂ul . + ǫjl = 2 ∂rl ∂rj

(3.76)

The long-wave length limit or continuum approximation, considered here, implies that the strain tensor field is locally homogeneous on a length-scale, limited by the wave length λ of the acoustic phonon mode. This homogeneous strain is related to a homogeneous stress, σik , by Hooke’s3 law σik = cikjl ǫjl , (3.77) jl

which belongs to the family of linear-response relations (see Chap. 1), with the fourth rank tensor of elastic constants as the response function. As the 3

Robert Hooke 1635–1703

3.4 Acoustic Phonons

53

stress and strain tensors are symmetric, one has cikjl = ckijl = ciklj = ckilj and the number of independent elastic constants cikjl is reduced to 21. If the strain is understood as the response of the elastic continuum to an applied stress, the inverted relation to (3.77) reads ǫij = sijkl σkl , (3.78) kl

with the elastic moduli or compliance constants sijkl . With Hooke’s law, the wave equation for the elastic continuum takes the final form ρM u ¨i =

∂ σik . ∂rk

(3.79)

k

By definition, the strain and stress tensors are symmetric, e.g., ⎞ ⎛ ǫ11 ǫ12 ǫ13 ǫ = ⎝ ǫ12 ǫ22 ǫ23 ⎠ , ǫ13 ǫ23 ǫ33

(3.80)

thus having in general six independent components, which frequently are cast into a vector or Voigt4 notation according to the scheme ik I

11 1

22 2

33 3

23 4

13 5

12 6

matrix notation Voigt notation

in which Hooke’s law (3.77) takes the form σI =

6

cIJ ǫJ .

(3.81)

J=1

The number of independent elastic constants is further decreased by crystal symmetry (as was the case for the force constants) and can be looked up in the literature [90, 91]. For cubic crystal structures, there are only three independent elastic constants, which in Voigt notation are c11 = c22 = c33 , c23 = c13 = c12 , and c44 = c55 = c66 (Problem 3.7). For this case, the wave equation (10.6) can be turned into eigenvalue equations for the components of the normalized eigenvector e, using u(r, t) ∼ e ei(q·r−ωt) . For the first component, it reads ω 2 ρM e1 = c11 q12 + c44 (q22 + q32 ) e1 + (c12 + c44 )q1 (q2 e2 + q3 e3 )

(3.82)

(3.83)

and similar equations for the other components are obtained by cyclic permutation of the indices. The eigenfrequencies depend on the direction of 4

Woldemar Voigt 1850–1919

54

3 Lattice Dynamics: Phonons

propagation and the surfaces of constant frequency are not spherical but warped. As a consequence, the group velocity is anisotropic and phonon wave packets propagate preferentially in the direction of zero curvature, giving rise to caustics. This phonon focusing is connected with the energy flux and can be measured by propagating heat pulses [74, 92, 93]. A further simplification is possible by assuming, instead of cubic symmetry, an isotropic solid (conceivable in the continuum limit), for which c11 = c12 + 2c44 and ω 2 ρM e = (c12 + c44 )q(q · e) + c44 q 2 e.

(3.84)

The solutions of this equation can clearly be distinguished as longitudinal waves, with eq and ρM ωL2 = (c12 + 2c44 )q 2 , and transverse waves, with 2 e ⊥ q and ρM ωT = c44 q 2 . The corresponding sound velocities v = ω/q are % c12 + 2c44 longitudinal sound velocity : vL = ρM % c44 transverse sound velocity : vT = < vL . (3.85) ρM Measuring the sound velocities in solids provides information about the elastic properties and allows one to determine the elastic constants (see Problem 3.7). In general, the velocities for longitudinal sound propagation (and the corresponding frequencies) are larger than those for transverse sound propagation in the same direction. Within the simple mechanical model this can be understood by the stronger distortions of the springs, if the masses are displaced parallel (and not perpendicular) to the propagation direction. This fact can also be expressed in terms of the compliance constants (or elastic moduli) sIJ . The isotropic solid is characterized by only two independent moduli, s11 and s12 , which are used to define [94] Young’s modulus E = 1/s11 rigidity modulus G = 1/ 2(s11 − s12 ) bulk modulus B0 = 1/ 3(s11 + 2s12 ) Poisson’s ratio ν = −s12 /s11

(3.86) (3.87) (3.88) (3.89)

corresponding to longitudinal (E) and transverse (G) distortions and volume changes (B0 ). The rigidity modulus also is called shear modulus. The bulk modulus is the inverse of the volume compressibility. The elastic properties of isotropic homogeneous systems can be characterized by two parameters, the Lam´e constants, which can be expressed by Poisson’s ratio and Young’s modulus according to νE (1 + ν)(1 − 2ν) E μ= . 2(1 + ν) λ=

(3.90) (3.91)

3.5 Optical Phonons

55

3.5 Optical Phonons Phonon dispersions of solids with a crystal structure, containing more than one atom in the unit cell exhibit, besides the acoustic phonons, additional branches starting with finite frequencies at q = 0 in the center of the Brillouin zone. An example has been dealt with in Problem 3.1. We know from its solution, that in this phonon mode the atoms of the unit cell move π out of phase with each other as demonstrated in Fig. 3.4. Some of these phonons are dipole active, i.e., they can be excited optically by infra-red light, thus, determining the response of the solid to the external electric field of an electromagnetic wave in this spectral range. They are, therefore, called optical phonons. Their properties will be the subject of this section. Let us consider the situation for a crystal unit cell with two ions carrying charges ητ , where τ = ± is the index for the basis, as depicted in Fig. 3.4. The electric dipole moment of the solid (here denoted as M ) can be expressed as M= (3.92) ητ R0nτ + unτ . n,τ

For the static lattice in equilibrium, M 0 = n,τ ητ R0nτ defines the spontaneous electric dipole moment. Solids with M 0 = 0 are called ferroelectric. We consider here the case M 0 = 0, which results for the considered binary solid with two ions with opposite charge in the unit cell. An electric field drives the two charged ions into opposite directions, thus, inducing a dipole moment. In a time-dependent electric field E(t), this system can be resonantly excited at the frequency of the optical phonon. In the language of the response formalism, the system Hamiltonian (3.41) has to be extended by the time-dependent perturbation (see (2.58)) Vext (t) = −M · E(t).

(3.93)

It causes a (dielectric) polarization with components Pi =

Mi = χij Ej , V j

i, j = 1, 2, 3

(3.94)

with the dielectric susceptibility χ as a response function, which is a symmetric second rank tensor. The vector components Pi and Mj can be identified with un+

un _

_

+ 0 Rn+

E(t)

R0n _

Fig. 3.4. Characteristic ion displacements in a unit cell for an optical phonon

56

3 Lattice Dynamics: Phonons

ˆ respectively, of the response formalism outlined in the observables Aˆ and B, Chap. 2. After Fourier transformation, we write i +∞ ˆ i (t), M ˆ j (0)]0 dt Ej (ω) Pi ω = eiωt θ(t)[M (3.95) hV −∞ ¯ j χij (ω)Ej (ω), (3.96) = j

with the dipole operator given by ητ 0 ˆ (t) = √ ˆ nτ (t) = ητ u esτ (q) eiq·Rn Qs (q, t). M N Mτ nτ sq nτ

(3.97)

The normal coordinate Qs (q, t) can be expressed in terms of phonon operators to write ˆ i (t) = M Mis (q) a†s (−q, t) + as (q, t) , (3.98) sq

with the component Mis (q) =

ητ

nτ

¯ h 2N Mτ ωs (q)

!1/2

0

esτ i (q) eiq·Rn

(3.99)

of the dipole moment connected with a phonon in the mode s, q. The dipole ˆ i (t), M ˆ j (0)]0 can be evaluated using Problem 3.5, correlation function [M together with ns (q, T ) = ns (−q, T ) to write [a†s (−q, t), as′ (q ′ , 0)]0 = −δss′ δ−qq′ eiωs (−q)t and yields, with ωs (q) = ωs (−q), ' & ˆ j (0)]0 = ˆ i (t), M Mis (q)Mjs (−q) −eiωs (q)t + e−iωs (q)t . [M

(3.100)

(3.101)

sq

Let us now consider the limit of long wavelengths λ ≫ a (where |q| = 2π/λ and a is the lattice constant), which holds for far-infrared light, to simplify Mi (q) with the dipole approximation 0

eiq·Rn = 1 + iq · R0n + . . . ≃ 1

(3.102)

and distinguish between longitudinal (s = L) and transverse phonons (s = T) with the properties M L eLτ (q)q

and M T eT τ (q) ⊥ q,

respectively.

(3.103)

In the first case and for an electromagnetic wave propagating in the direction of q, one has M L ·E = 0, i.e., the transverse light cannot excite a longitudinal

3.5 Optical Phonons

57

phonon mode. In the second case, there is a contribution to P ω and thus, to the susceptibility. In the long-wavelength limit, with MiT (q

h ¯ 2ωT (0)

→ 0) =

!1/2

ητ

τ

%

N T e (0) =: MiT , Mτ τ i

(3.104)

one finds for the dielectric susceptibility

iMiT MjT∗ Γ→0 hV ¯

χij (ω) = lim

∞

0

&

' ei(ωT (0)+ω+iΓ)t − e−i(ωT (0)−ω−iΓ)t dt.

(3.105)

Here, the exponential with the parameter Γ has been introduced for the adiabatic switching on of the perturbation Vext of (3.93), and the lower integration limit is a consequence of the step function in (3.95). After integration, one has MiT MjT∗ Γ→0 hV ¯

χij (ω) = lim

1 1 + ωT (0) + ω + iΓ ωT (0) − ω − iΓ

.

(3.106)

The susceptibility exhibits a characteristic pole structure in the complex ωplane, with poles at ω = ±ωT (0) − iΓ (see Fig. 3.5a) and frequencies ωT (0) = ωT in the infra-red spectral range. For solids with a cubic lattice, χij (ω) simplifies to a scalar χij (ω) = χ(ω)δij , which determines the complex frequency-dependent dielectric function ε(ω) = ε∞ +

1 χ(ω) = ε1 (ω) + iε2 (ω), ε0

(3.107)

where ε0 is the vacuum dielectric constant, while ε∞ accounts for the dielectric background, which is caused by contributions of oscillators in the electronic

Imω

-ωT(0)

Imε(ω)

ωT(0)

-ω T(0) ω T(0)

Reω

(a)

ω

(b)

Fig. 3.5. (a) Complex ω plane with poles of the dielectric function; (b) Real part of the dielectric function without (solid) and with (dashed) damping vs. frequency

58

3 Lattice Dynamics: Phonons

system of the solid (see Chaps. 4 and 5), with resonances at much higher frequencies than those of the optical phonons. Therefore, it is called highfrequency dielectric constant. The imaginary part of the dielectric function is related to the absorption coeﬃcient (see Chap. 10) and takes the form ε2 (ω) = Im

χ(ω) ε0

|M T |2 = lim Γ→0 ¯ hV ε 0

+

Γ 2

(ωT − ω) + Γ2

−

Γ 2

(ωT + ω) + Γ2

,

. (3.108)

Each term in the last bracket is a Lorentzian, which for Γ → 0 becomes a δ function peaked at ω = −ωT and ω = ωT (see Fig. 3.5b), corresponding to emission and absorption respectively of an optical phonon at q = 0. Note, that this final result with the δ functions would also result by applying (2.77) directly to (3.106) for i = j. Combining both terms, the real and imaginary part of the dielectric function can be written as

ε1 (ω) = ε∞ + lim

Γ→0

2 − ω2) |M T |2 2ωT (ωT , 2 hV ε0 (ωT − ω 2 )2 + ω 2 Γ2 ¯

|M T |2 2ωT ωΓ . 2 Γ→0 ¯ hV ε0 (ωT − ω 2 )2 + ω 2 Γ2

ε2 (ω) = lim

(3.109) (3.110)

We note in passing that ε1 (ω) and ε2 (ω) are connected with each other by the Kramers–Kronig relations (see Sect. 2.6). The parameter Γ is introduced as a mathematical trick to regularize the integral in (3.105). But in a more general view, it can be understood as also being caused by interactions not explicitly included in the present model. By taking into account electron–phonon interaction (see Chap. 7) or anharmonic eﬀects due to higher order terms in the expansion (3.6) it would be finite, thus, indicating a finite phonon lifetime or a damping of the phonon oscillator. In fact, an optical phonon can decay into two acoustic phonons. Microscopically, such processes mean dissipation of energy from the optical phonon to other degrees of freedom. We go into a more detailed discussion of the optical properties by looking at ε1 (ω) and the reflection coeﬃcient R(ω), which are plotted in Fig. 3.6. For ω = 0 (and Γ = 0), the real part of the dielectric function can be written as 2 ε(0) = ε∞ +S/ωT , known as the static dielectric constant, while for frequencies ω >> ωT (or ω → ∞ in (3.109)) it approaches the high frequency value ε∞ . In-between, it passes through the resonance at ωT and is negative up to a frequency identified with that of the longitudinal phonon ωL at which ε1 (ωL ) = 0. Without damping, this relation yields 2 ωL2 = ωT +

S 2 > ωT . ε∞

(3.111)

3.5 Optical Phonons

59

ε1(ω)

ε(0) ε∝ 1 0

ω

1

R(ω) 1

0

ωT

ωL

ω

Fig. 3.6. Real part of the dielectric function (upper part) and reflection coeﬃcient (lower part) without (solid) and with (dashed) damping vs. frequency

By including Γ, one obtains the dashed curve for ε1 (ω) in Fig. 3.6, which is characteristic for a damped oscillator with frequency ωT . The reflection coeﬃcient is defined by the complex index of refraction N = n1 + in2 (all quantities depending on ω): " " " 1 − N (ω) " "= R(ω) = "" 1 + N (ω) "

(1 − n1 (ω))2 + n22 (ω) (1 + n1 (ω))2 + n22 (ω)

!1/2

.

(3.112)

Without damping, using ε1 = n21 − n22 , and ε2 = 2n1 n2 , we find in the interval ωT < ω < ωL that ε1 is negative while ε2 vanishes, which leads to n1 = 0 and n2 = 0 and consequently to R(ω) = 1. In other words, the solid cannot propagate light in this frequency range, i.e., all light is perfectly reflected from the surface. It is called Reststrahlen band because the reflected light is dominated by these frequencies. Taking into account damping gives the more realistic (dashed) curve shown in Fig. 3.6. It is in accordance with experimental spectra, which can be picked up from textbooks like [89] or from data collections like [94]. 2 From (3.111), we have S = ε∞ (ωL2 − ωT ) and can express the (real part of the) dielectric function (see (3.109) for Γ = 0) as ε(ω) = ε∞

ωL2 − ω 2 2 − ω2 ωT

(3.113)

in terms of the characteristic optical phonon frequencies and the material constant ε∞ . For ω = 0, this expression reduces to

60

3 Lattice Dynamics: Phonons

ε(0) ω2 = L2 ε∞ ωT

(3.114)

known as the Lyddane–Sachs–Teller5 relation. Making use of this relation, we also may write S = ωT2 (ε(0) − ε∞ ) in terms of the macroscopic quantities ωT , ε(0), and ε∞ . On the other hand, S = 2ωT |M T |2 /¯hV ε0 is determined by the microscopic parameters of the system (see (3.104)), the charge η± = ±η and the masses M± of the ions. In the long-wavelength limit of# the optical mode the latter move against each other with eigenvectors eT (M∓ /M ), = ± ± where M = M+ + M− . This allows us to express η=

!1/2 μV ε0 (ε(0) − ε∞ ) ωT N

(3.115)

in terms of the macroscopic material parameters (here the reduced mass μ = (1/M+ + 1/M− )−1 appears because of the relative motion of the two ions in the unit cell, see Problem 3.1). As this expression for η contains also the transverse phonon frequency ωT , it is called the transverse charge [89]. It is related to the strength of the phonon oscillator and can be determined from the measured spectrum by a line-shape fit. In crystalline solids with more complex unit cells than the one with two oppositely charged ions assumed here, there are several triples of optical phonon branches with diﬀerent longitudinal-transverse splittings, giving rise to diﬀerent transverse eﬀective charges [95].

3.6 Examples: Phonon Dispersion Curves Phonon dispersion curves, showing the phonon frequencies ωs (q) for diﬀerent branches s, are usually plotted versus q along diﬀerent high symmetry directions in the Brillouin zone. For the examples to be discussed in this section, which all have fcc or bcc point lattices, we refer to the Brillouin zones depicted in Figs. 1.1 and 1.2. Phonon dispersion curves are obtained either experimentally from inelastic scattering preferentially with neutrons (for a more recent introduction and examples see [96, 97]) and also with photons and atoms, or from model calculations of diﬀerent sophistication. Both kinds of investigation have influenced and stimulated each other and are well documented [77–79], thus, at present the phonon dispersion curves of solids are well known. For collections of phonon dispersion curves, together with a compilation of the original references, we refer to [76, 81, 94]. A selection will be presented and discussed in this section to provide the knowledge how to read phonon dispersion curves and understand their principal material specific features. 5

Russell Hancock Lyddane 1913–2001, Robert Green Sachs 1916–1990, Edward Teller 1908–2003

3.6 Examples: Phonon Dispersion Curves

61

In all the examples presented below, experimental data are shown together with calculated dispersion curves. In order to judge the quality of the theoretical data, it is necessary to briefly characterize the diﬀerent models used. Historically (and this has been in the first place a matter of available computer power), a variety of phenomenological models has been developed. Depending on their complexity, these models are characterized by a number of parameters, which have to be determined by fits to the experimental data. We have already learned about the simplest of these phenomenological models with massive spheres (representing rigid ions) and springs (representing the interionic forces due to chemical binding) in Sects. 3.1 and 3.2 (including Problems 3.1 and 3.2). For given masses of the ions, this rigid-ion model contains the force constants as adjustable parameters (their number depending on the symmetry of the lattice and the considered number of neighbors coupled by springs, see Problem 3.2). As demonstrated by comparing measured with fitted dispersion curves for Al in Fig. 3.7 and for Fe in Fig. 3.8, this model works well for metals. However, this rigid-ion model does not account for a polarization connected with a motion of the ions in ionic crystals. For these systems, so-called shell models of diﬀerent complexity have been developed. By decomposing the ion into the nucleus with the closed shells of core electrons and the shell of valence electrons, responsible for the chemical binding, it becomes possible to introduce additional parameters or force constants representing springs between each two movable parts of the model. Especially, a spring between the closed-shell ion and the shell of valence electrons can be used to account for the polarization of the atom. With this increased number of parameters, shell models are quite flexible and have been successfully applied to ionic crystals with heteropolar binding like KI (Fig. 3.9). KI crystallizes in the rocksalt structure with two ions in the Wigner–Seitz cell and, therefore, its phonon dispersion curves have acoustic (TA, LA) and optical branches (TO, LO).

ω [1013 rad s–1]

Al T2

6 L

L

4

0

T1

T

2

Χ

∆

Γ

Σ

Κ

Χ

Fig. 3.7. Phonon dispersion curves for Al (fcc structure). Symbols are experimental data from inelastic neutron scattering, dispersion curves calculated with the rigid ion model using three force constants fitted to elastic constants (after [77])

62

3 Lattice Dynamics: Phonons 10

Fe

υ [THz]

8 L

T

6 4 2 Γ

H

∆

F

P

Λ

Γ

Σ

N

Fig. 3.8. Phonon dispersion curves for α−Fe (bcc structure). Symbols are experimental data from inelastic neutron scattering, dispersion curves calculated with the rigid ion model using force constants for up to five nearest neighbors (after [81])

υ [THz]

KI 4

L0

3

T0 LA

2

TA

1

Γ

∆

Χ

Σ

Γ

Λ

L

Fig. 3.9. Phonon dispersion curves for KI (rocksalt structure). Symbols (open for L, full for T modes) are experimental data from inelastic neutron scattering, dispersion curves calculated with an 11-parameter shell model (after [76])

In solids with dominating covalent binding, like those made of the group-IV elements, which crystallize in the diamond structure (two interpenetrating fcc lattices with four nearest neighbors to each atom), the valence electron density exhibits a strong maximum halfway between each pair of nearest neighbors (see Chap. 5). This bond charge is characteristic for covalent binding. It turns out to be important not only for the stability of the diamond structure but also for the lattice dynamics. Because the rigid-ion models as well as the shell models fall short of describing the bond charge and its motion, a bond charge model has been invented, which introduces springs between neighboring bond charges. This model has been used to calculate the phonon dispersion curves of Si in Fig. 3.10. Of particular nature are solid rare gases, which, due to the weak van der Waals forces, exist only at low temperatures. Their characteristic phonon frequencies are much smaller than those of other solids, therefore, already very

3.6 Examples: Phonon Dispersion Curves

63

15 Si υ [THz]

10

L

L

L T2

5

T

T2

T 0 Γ

∆

Χ Κ

Σ

Γ

L

Λ

Fig. 3.10. Phonon dispersion curves for Si (diamond structure). Symbols (open for L, full for T modes) are experimental data from inelastic neutron scattering, dispersion curves calculated with the bond-charge model (after [76])

4

hω [meV]

3.0

He(bcc)

L

2.0

L

L 1.0

T2

T

T1 Γ

∆

Η

Λ

Γ

Σ

Ν

4

Fig. 3.11. Phonon dispersion curves for solid He in the bcc structure. Symbols (open for L and full for T modes) are experimental data from inelastic neutron scattering, dispersion curves calculated with the self-consistent harmonic approximation (after [76])

little thermal energy creates lattice displacements that require the consideration of anharmonic corrections. This is done in the self-consistent harmonic approximation (SCHA) used in Figs. 3.11 and 3.12 (see also Problem 3.8). These diﬀerent models, although successful in many cases, turned out to be in conflict with experimental data of increasing accuracy. Discrepancies were found in particular for phonon eigenvectors, which have been measured besides the frequencies [98]. At the same time, theoretical concepts and computer power became available for microscopic lattice dynamics based on ab-initio methods [99, 100]. These methods aim at the self-consistent solution of the time-dependent Schr¨ odinger equation for the motion of the ions in the adiabatic potential (3.1) P2 n ˙ + U({Rn }) Φ = i¯hΦ 2M n

(3.116)

64

3 Lattice Dynamics: Phonons 4

υ [THz]

3

He(fcc)

2 1

Γ

∆

Χ

Σ

Γ

Λ

L

Fig. 3.12. Phonon dispersion curves for solid 4 He under hydrostatic pressure in the fcc structure. For symbols (experiment) and lines (model calculations) see caption of Fig. 3.11 (after [76])

wave number [cm–1]

400 GaAs 300 200 100 0

Γ

Κ

Χ

Γ

L

DOS

Fig. 3.13. Phonon dispersion curves and density of states (DOS) for GaAs (zinc blende structure). Symbols are experimental data from inelastic neutron scattering, dispersion curves are ab-initio results from DFPT (after [101])

and of the time-independent Schr¨ odinger equation for the electrons in the equilibrium configuration of the ions p2 e2 1 l + + v(r l − Rn ) Ψ0 = Eel Ψ0 . 2m 8πε0 k,l |rk − rl | l

k=l

(3.117)

n,l

This becomes possible within the density–functional theory (DFT), which will be introduced in Chap. 5, as an extension to the density–functional perturbation theory (DFPT) [98–100]. As an example of such calculations we show phonon dispersion curves of GaAs in Fig. 3.13, together with the phonon density of states. When looking at the phonon dispersion curves, one finds diﬀerent units for the phonon frequencies ν, wave numbers ν¯, or energies ¯hω, following from the relations E = h ¯ ω = hν = hc¯ ν . The scales are typically THz or 1012 s−1

3.6 Examples: Phonon Dispersion Curves

65

for ν, 100 cm−1 for ν¯, and a few meV for h ¯ ω. In order to convert these units for comparison of the diﬀerent dispersion curves, the following scheme can be used (e.g., 1 THz = 1012 s−1 ≡ 4.136 meV):

THz 100 cm−1 10 meV

THz

100 cm−1

10 meV

1 2.9979 2.4181

0.3336 1 0.8066

0.4136 1.2398 1

Let us first discuss the phonon dispersion curves of solids with Bravais lattices (with only one atom per unit cell) shown for Al (Fig. 3.7), Fe (Fig. 3.8) and solid 4 He (Figs. 3.11 and 3.12). The dispersion is depicted for the directions [100], [111], and [110] in the Brillouin zone which correspond, respectively, to the lines ∆, Λ, and Σ (see Figs. 1.1 and 1.2). In each case, one has only acoustic branches, one longitudinal (L) and two transverse ones (T or T1 , T2 ). The transverse branches are degenerate for the [100] and [111] directions, whose group of the wave vector contains fourfold and threefold rotations, respectively, but split into two separate branches for the [110] direction (which is invariant only under twofold rotation). By comparison of bcc 4 He and Fe on one side and those of fcc 4 He and Al we recognize a remarkable correlation between lattice structure and topology of the phonon dispersion curves. This is most striking for the dispersion along the Λ line (or [111] direction), for which the longitudinal branch from Γ to H of bcc 4 He and Fe exhibits a pronounced dip, while it increases monotonously from Γ to L for the fcc solids. For a given crystal structure, the diﬀerence in the material and its specific chemical binding determines the force constants and therefore, the phonon frequencies. Thus, the phonon spectrum of the metals Al and Fe spreads over a much larger interval than those of both modifications of solid 4 He. But already the comparison of bcc 4 He and fcc 4 He (obtained by applying hydrostatic pressure) shows an increase of the phonon frequencies as the strength of the chemical binding (here van der Waals forces) is increased by compression. The next group of examples consists of solids, all based on the fcc point lattice but with diﬀerent basis: diamond structure for Si (Fig. 3.10), zinc blende structure for GaAs (Fig. 3.13), and rocksalt structure for KI (Fig. 3.9). With two atoms per unit cell, these solids have phonon dispersion curves exhibiting three optical branches with similarities as well as pronounced diﬀerences inside the acoustic branches. For the acoustic branches, with dispersion curves similar to those of Al and fcc He4 , we mention only the diﬀerences in the slopes. They are largest for Si, which, similar to C(diamond), is a rather rigid material due to the covalent bonds, and its Debye temperature is about 50% larger than that of Al. Note the degeneracy of the transverse phonons in KI along the Σ axis. It is a consequence of the higher symmetry of the rocksalt lattice compared to that of the zinc blende lattice. The optical branches of Si diﬀer

66

3 Lattice Dynamics: Phonons

from those of GaAs and KI by convergence toward a degeneracy with the acoustic branches, e.g., at the X point, while these two groups of branches do not interpenetrate for GaAs and are even well separated by a gap for KI. This behavior can be understood from the solution for Problem 3.1, the linear chain with two atoms in the unit cell: If the two atoms in the unit cell have diﬀerent masses, the dispersion curves show a gap at the boundary of the Brillouin zone (as for GaAs and KI); this gap closes, if the masses are equal (as for Si). Even the diﬀerence between these gaps of GaAs and KI can be explained within this model, it increases with increasing mass diﬀerence. The other striking diﬀerence is the splitting of the optical phonon branches at the Γ point for GaAs and KI – the longitudinal–transverse (or LT) splitting – while these branches are threefold degenerate in Si. The splitting is a consequence of the macroscopic polarization inherent with a longitudinal optical mode in a binary compound solid (such as GaAs and KI) that gives rise to a stronger restoring force than for the transverse mode, i.e., ωL > ωT , as discussed in Sect. 3.5, and is connected with the Reststrahlen band. The rather flat optical phonon branches lead to a pronounced peak in the density of states (DOS), shown for GaAs in Fig. 3.13. Nevertheless, the phonon dispersion curves for Si and GaAs resemble each other much more than those of GaAs and KI. This is due to the similarity of the diamond and zinc blende structure and the small mass diﬀerence between Ga and As. On the other hand, the characteristic frequencies are higher in Si than in GaAs, which can be ascribed to a weakening of the covalent binding in GaAs (which becomes partially ionic) and to the larger masses of Ga and As as compared to the mass of Si. With these aspects of selected phonon curves in mind, it is not diﬃcult to make an excursion to other materials maintaining the same crystal structure but replacing the atoms. As an example, we may consider AlAs and GaP, both in the zinc blende structure. Compared to GaAs, we expect a more pronounced separation of acoustic and optical branches in the phonon spectrum of both materials, because the mass diﬀerence of the two atoms in the unit cell, taken from diﬀerent rows of the periodic table, has increased. Moreover, due to the lighter masses of Al and P, compared to those of Ga and As, respectively, the characteristic phonon frequencies are higher than those of GaAs. These features are found in the phonon dispersion curves of AlAs and GaP (see [94]). The lattice dynamics presented in this chapter is designed for the extended solid without regarding its surface. Surface atoms experience a diﬀerent structural surrounding and forces, which diﬀer from those acting on the bulk atoms. Consequently, they have their own dynamics, which, in a simplified two-dimensional model, has been treated already in Problem 3.2. For a more detailed description we refer to [85–87]. A simple example is the phonons of a Cu(100) surface as shown in Fig. 3.14. Calculations of surface phonons are usually performed for slab configurations where the outermost atomic layers experience the modified environment of the surface, while the central layers reproduce the bulk situation. The results are plotted for wave vectors in the first BZ of the 2D periodic surface structure. The shaded area in Fig. 3.14

3.6 Examples: Phonon Dispersion Curves

67

Cu (100) 8

υ [THz]

6 4 2

Γ

Χ

Fig. 3.14. Projected bulk phonons and surface phonons for a Cu(100) surface. Experimental data (circles) are from HREELS (after [102])

results from the bulk modes that have the same component of the wave vector in the surface along Γ − X, while the solid lines represent the collective modes of the surface layers, whose frequency is below the continuum of the projected bulk phonon frequencies. The surface phonons have been measured by high resolution electron energy loss spectroscopy (HREELS). A more complex spectrum results for the clean GaAs(110) surface shown in Fig. 3.15. This surface represents a rectangular 2D lattice, whose BZ is depicted in the inset. Results from DFPT calculations [103] show again the projected bulk phonon spectrum (shaded area), but in addition solid lines from surface phonons, which partially overlap with the bulk continuum. Experimental data are from inelastic scattering with He atoms and from HREELS. Adsorbate layers modify the surface phonon spectrum. Modern epitaxial material growth has created new solids with artificial periodicity by periodically changing the kind of atoms supplied during the growth process [104, 105]. These new materials are called superlattices. Their period is larger than that of the underlying crystal structure and thus, the Brillouin zone and the phonon branches for the growth direction are backfolded. If the masses of the atoms in the diﬀerent layers diﬀer strongly from each other, the optical phonons of one layer may fall into a frequency range outside of the optical phonon branch of the other layer. This means that these optical modes cannot propagate along the growth direction but only along their respective layers. They are called confined phonon modes (see Sect. 9.3 in [89]). As already mentioned in Sect. 1.2, solids are, in general, compositionally disordered due to the natural abundance of isotopes for each element. As a consequence, not all oscillators in the model of Fig. 3.1 have the same resonance frequency. This becomes relevant especially for the optical phonons

68

3 Lattice Dynamics: Phonons

υ [THz]

energy [meV]

GaAs (110)

Fig. 3.15. Projected bulk phonons and surface phonons for a clean GaAs(110) surface. Experimental data are from inelastic He scattering (triangles) and HREELS (squares) (after [103])

and results in an inhomogenously broadened infrared absorption line. This broadening can be (and has been) overcome by growing isotopically clean crystals [106]. The mass diﬀerence of isotopes also has been exploited in growing isotope superlattices, e.g., of 70 Ge and 74 Ge and to investigate their phonon spectra by inelastic light scattering [107].

3.7 The Crystal Lattice at Finite Temperature The experimental phonon energies presented in the last section were all obtained from inelastic scattering. Besides the most frequently employed and versatile neutron scattering, inelastic scattering with photons and atoms is also in use [96, 97]. A quantitative analysis has to ask for the scattering intensity including its dependence on the temperature. The key quantity is the diﬀerential cross section. It describes the probability that an incoming particle (neutron, photon, light atom) with momentum h ¯ k is scattered by a solid (or its surface) into an outgoing particle with momentum ¯hk′ . In the scattering process, the energy h ¯ ω (the diﬀerence of the kinetic energies of the particle before and after the encounter with the solid) and the momentum hq = h ¯ ¯ (k − k′ ) are exchanged with the solid. Under the assumption of weak scattering eﬃciency, which justifies the Born approximation, an excitation with just this energy and momentum is created (or destroyed) in the solid. For this situation, the scattering cross section can be written as [4, 83, 96, 97] d2 σ = Aq S(q, ω), dωdΩ

(3.118)

3.7 The Crystal Lattice at Finite Temperature

69

where Ω refers to the direction of the scattered particle relative to the direction of the incoming particle, Aq accounts for the details of the interaction mechanism giving rise to the scattering, and S(q, ω) is the dynamic structure factor (see Sect. 1.3). It is the Fourier transform of the density–density correlation function ′ 1 d3 r d3 r′ eiq·(r−r ) e−iωt n(r, t)n(r ′ , 0) dt. S(q, ω) = (3.119) 2π Its static counterpart has been introduced already in Sect. 1.3. The time dependence of the density n(r, t) = δ(r − Rn (t)), Rn (t) = R0n + un (t) (3.120) n

has its origin in the time-dependent positions of the ions, which can be expressed by the departures from their equilibrium positions R0n (here we assume a Bravais lattice). Thus, the dynamical structure factor can be written as 0 0 1 S(q, ω) = dt e−iωt e−iq·(Rn +un (t)) eiq·(Rn′ +un′ (0)) . (3.121) 2π ′ nn

When evaluating the thermal expectation value under the integral, the displacements un have to be understood as quantum-mechanical operators (in actual calculations they will be expressed by phonon operators), while the equilibrium positions are parameters. Therefore, we can write 0 0 . . . = e−iq·(Rn −Rn′ ) e−iq·un (t) eiq·un′ (0) 0 0 = e−iq·(Rn −Rn′ ) e−iq·(un (t)−un′ (0)) e[q·un (t),q·un′ (0)]/2 , (3.122) where the last expression was obtained by using the operator relation ˆ ˆ

ˆ

ˆ

ˆ ˆ

eA eB = eA+B e[A,B]/2 .

(3.123)

If the lattice sum of the prefactor with the diﬀerence R0n − R0n′ could be performed irrespective of the second factor, it would give a δq,G , indicating the Bragg peaks, which correspond to the set of parallel lattice planes defined by the reciprocal lattice vector G (see Problem 1.6). In fact, the second factor simplifies for elastic scattering, which corresponds to the limit t → ∞, where all correlations are lost. Thus, the commutator in the exponential vanishes and we may write "2 " lim e−iq·un (0) eiq·un′ (t) = " e−iq·un " = e−2W . (3.124) t→∞

This is the Debye–Waller factor, with W = (q·u)2 /2. It is independent of the lattice site (because of the translational symmetry) and describes the reduction of the intensity of the Bragg peaks due to the thermal motion of the ions.

70

3 Lattice Dynamics: Phonons

For a more detailed evaluation of . . . (3.122), we first restrict to the longitudinal modes by writing q · un = un and express the displacement by phonon operators (see (3.23) together with (3.39)) 0 0 1 # un (t) = a(q) ei(q·Rn −ω(qt)) + a† (−q) e−i(q·Rn −ω(q)t) . 2N M ω(q) q

(3.125)

We find that the commutator 1 1 aq , a†q eiαn′ n (t) + a†q , aq e−iαn′ n (t) [un′ (t), un (0)] = 2N M q ω(q) i 1 sin (αn′ n (t)) , N M q ω(q)

=

(3.126)

with αn′ n (t) = ω(q)t + q · (R0n′ − R0n ), is a scalar and the exponential with this argument becomes a factor in front of the thermal expectation value of the first exponential. For a harmonic oscillator, it can be shown [4] that ! iq(u ′ (t)−un (0)) 2 q 2 un′ (t) − un (0) . (3.127) e n = exp − 2 We evaluate the argument of the exponential with the substitution (3.125) and obtain 2 1 1 un′ (t) − un (0) = (1 − cos αn′ n (t)) 2a†q aq + 1 (3.128) N M q ω(q)

up to terms containing two creation or two annihilation operators (which would vanish when taking the thermal expectation value). Thus, the thermal expectation value in (3.122) can be expressed as

& q2 1 ) q2 1 ′ 2n(q , T ) + 1 exp 2N M ′ ω(q ′ ) 2N M ′ ω(q ′ ) q q ' ′ × 2n(q , T ) + 1 cos αn′ n (t) + i sin αn′ n (t)

( . . . = exp −

where n(q, T ) is the phonon occupation function. The first exponential, which is time-independent, can be identified as the Debye–Waller factor (see Problem 3.6), while the second exponential can be expanded with the leading terms q 2 1 ≃ 1+ 2n(q ′ , T ) + 1 cos αn′ n (t) + i sin αn′ n (t) ′ 2N M ′ ω(q ) q

2

= 1+

q 2N M

q′

) 1 ( n(q ′ , T ) + 1 eiαn′ n (t) + n(q ′ , T ) e−iαn′ n (t) ′ ω(q )

3.7 The Crystal Lattice at Finite Temperature

71

and we obtain, after time-integration and performing the lattice sums for the structure factor, S(q, ω) = δ(ω) e−2W

δq,G

G 2

( q n(q ′ , T ) + 1 δ ω − ω(q ′ ) δq,q′ +G ′) 2N M ω(q q′ ) (3.129) + n(q′ , T )δ ω + ω(q ′ ) δq,q′ −G .

+

The first term is the elastic peak that corresponds to the static structure of the lattice. It is modified by the thermal motion as quantified by the Debye–Waller factor. The second and third term, are the lowest order inelastic contributions describing absorption and emission, respectively, of a single phonon, with their particular energy and momentum conservation. Note that phonon emission is possible only if phonons are thermally excited, i.e., if n(q, T ) = 0. Morephonon processes would follow from the higher order terms in the expansion of the exponential. One of the most prominent eﬀects that heat causes in matter is thermal expansion. It is quantified by the linear thermal expansion coeﬃcient at constant pressure P ! 1 ∂l α= . (3.130) l ∂T P Here, l is the linear extension of a piece of matter. With the substitution ∆V /V = 3∆l/l, this can also be written as ! ! ! (∂P/∂T )V 1 1 1 ∂V ∂P − = = , (3.131) α= 3V ∂T P 3V (∂P/∂V )T 3B ∂T V with the bulk modulus B0 = −V (∂P/∂V )T .

(3.132)

The thermal expansion can be expressed in terms of phonons by considering the latter as a gas and using the statistical expression for the free energy F = U − T S, from which the pressure is obtained by ! ∂ ∂F (U − T S) . (3.133) =− P =− ∂V T ∂V Using S(T ), obtained by integrating the diﬀerential equation ! ! ∂U ∂S =T , ∂T V ∂T V

(3.134)

72

3 Lattice Dynamics: Phonons

one can write (3.133) in the form T 1 ∂U (T ′ , V ) ∂ ′ U (T, V ) − T P (T, V ) = − dT . ′ ∂V ∂T ′ 0 T

(3.135)

The inner energy U of the phonon gas is given by U = Tr(ρ0 H) =

hωs (q) ns (q, T ) + ¯

s,q

1 2

!

,

(3.136)

with the Hamiltonian H from (3.41), and we find ! ∂ 1 ∂ hωs (q) + ¯ ¯hωs (q) . P (T, V ) = − ns (q, T ) − ∂V s,q 2 ∂V s,q Taking the derivative with respect to the temperature we arrive at ! 1 ∂ ∂ hωs (q) ¯ ns (q, T ). − α= 3B0 s,q ∂V ∂T

(3.137)

(3.138)

This expression allows one to recognize two mechanisms responsible for the thermal expansion in terms of the phonons: One is the change of the thermal occupation with the temperature, the other is the change of the phonon energy with the volume. The latter is specified by the mode Gr¨ uneisen parameter γsq = −

∂(ln ωs (q)) V ∂ωs (q) =− . ωs (q) ∂V ∂ ln V

(3.139)

In order to characterize the temperature dependence of the whole spectrum with a single parameter, the Gr¨ uneisen parameter γ(T ) =

hωs (q) ∂ ¯ 1 ns (q, T ) γsq cV s,q V ∂T

(3.140)

is introduced. It is the average of the mode Gr¨ uneisen parameter weighted with the contribution of each mode to the lattice specific heat cV =

¯ hωs (q) ∂ ns (q, T ). V ∂T s,q

(3.141)

uneisen The quantities α, B0 , γ, and cV are connected with each other by the Gr¨ relation 3B0 (T )α(T ) = γ(T )cV (T ).

(3.142)

For reasons of technical applications material research has been conducted very early on solids that do not show thermal expansion, at least in some relevant temperature range. The prominent first outcome has been the alloy

3.7 The Crystal Lattice at Finite Temperature

73

Fe0.65 Ni0.35 (called INVAR) found by Guillaume.6 A more recently investigated material to show this property is YbGaGe [108]. Although this eﬀect is not yet understood microscopically, there are indications that it is related to the electronic properties of these materials.

Problems 3.1 Consider a linear chain with two diﬀerent masses (M1 , M2 ) per unit cell (lattice constant a) connected by springs (spring constant f ) as shown in Fig. 3.16. Give the possible diﬀerent force constants of the model and verify the symmetry properties of the matrix of force constants presented in Sect. 3.1! Formulate the dynamical matrix and solve the eigenvalue problem! Discuss the dispersion ωs (q) close to the center and the boundary of the Brillouin zone and visualize the corresponding motion of the masses! What happens for M1 = M2 = M ? 3.2 In a two-dimensional square lattice, let elastic forces (described by harmonic force constants Φ1 , Φ2 ) act between the nearest and next nearest neighbors (see Fig. 3.17). Calculate the frequencies of the lattice vibrations ω(q) for wave vectors q along the symmetry lines Γ − X, Γ − M , and X − M in the Brillouin zone! Discuss the eigenvectors (normal coordinates) and visualize the vibrations! 3.3 Explain the concept of periodic (or Born–von K´ arm´ an) boundary conditions! What are they good for? Show explicitly the relation between the

n M1

n+1 M2

r0 a

Fig. 3.16. Linear chain model with two diﬀerent masses per unit cell

M Φ2

Φ1

a

Γ

Χ

Fig. 3.17. Model for a quadratic lattice (left) and corresponding Brillouin zone with symmetry points (right) 6

Charles–Edouard Guillaume 1861–1938, Noble prize in physics 1920

74

3 Lattice Dynamics: Phonons

crystal volume and the possible values of the wave vector q and give their number! How can a sum over all q be converted into an integral? 3.4 Verify the commutation relations (3.38) for the phonon creation and annihilation operators a†s (q) and as′ (q ′ ) by making use of those introduced for the normal coordinates and their conjugate momenta. 3.5 Show the validity of the following thermal expectation values: as (q, t)a†s′ (q ′ , 0) = δss′ δqq′ (ns (q, T ) + 1) e−iωs (q)t a†s (q, t)as′ (q ′ , 0) = δss′ δqq′ ns (q, T )e+iωs (q)t as (q, t)as′ (q ′ , 0) = 0 !

3.6 The thermal motion of the lattice reduces the scattering amplitude (see Problem 1.6) of the rigid lattice by the Debye–Waller factor exp(−2W ) with W = (q · u(0))2 /2. Here, u(0) is the ion displacement in the unit cell which is taken as the origin. Show, that (q · u(t))2 does not depend on t! Calculate W by making use of the thermal expectation values given in Problem 3.5! Discuss the temperature dependence of W in the high and low temperature limits! Use the Debye model for the latter case! 3.7 (a) Show that due to spatial symmetry the number of independent elastic constants for a cubic crystal is reduced to 3 (make use of the transformation properties of a tensor!) (b) Solve for this case the equation of motion (3.71) for the elastic continuum with ui (r, t) = ui exp i(q · r − ωt) for q along Γ − X and Γ − K! How do the sound velocities depend on the elastic constants? Take (measured or calculated) phonon dispersion curves (e.g., those of GaAs, Fig. 3.13) and determine the nonvanishing elastic constants via the sound velocities! To compare with values from the literature see [94]! 3.8 Calculate the correction to the frequency of an oscillator in its ground state due to a cubic anharmonicity. It arises by expanding the adiabatic potential beyond the harmonic terms. This corresponds to an oscillator problem with the Hamiltonian H =h ¯ ω0 a† a + ∆(a† + a)3 . Treat the anharmonicity by bringing the third order terms in the phonon operators first into normal order and reduce them by replacing the number operator whenever possible by the thermal expectation value n(T ). The self-consistent harmonic phonon approximation consists in considering the anharmonic correction in Brillouin–Wigner perturbation theory. Find the lowest eigenvalue and discuss its dependence on T to understand the meaning of a soft mode.

4 The Free Electron Gas

The structure of a solid, its composition of atoms, as well as the electronic, optical, and lattice properties are essentially determined by the electrons. However, not all electrons of the atoms constituting the solid are involved in the same way. At the beginning of Chap. 2 we have distinguished between the core electrons and the valence electrons. The former are tightly bound to the nuclei and extend over a distance (much) smaller than the lattice constant, while the wave functions of the latter overlap with those of the neighboring atoms, thus giving rise to the chemical binding. Therefore, we describe the solid as being composed of ions (nuclei plus closed shell electrons) in equilibrium positions at R0n and valence electrons, which are responsible for the stability of the solid and its electronic properties. They are ruled by the Hamiltonian (see discussion of (2.12))

H=

N N p2l e2 1 + 2m 2 k,l=1 4πε0 |r k − r l | l=1

+

n,l

k=l

v(r l − R0n ) +

1 V (R0n − R0m ), 2 mn

(4.1)

where N denotes the number of electrons. The first and second term describe the kinetic energy of the electrons and the electron–electron interaction, respectively. The third term, the interaction between valence electrons and ions, and the fourth term, the ion–ion interaction, depend on the equilibrium positions R0n of the ions (for simplicity a Bravais lattice is assumed) and are responsible for the structural aspects. The structure of the solid is defined by the configuration {R0n , all n} with the lowest eigenvalue of H. This operator comprises all terms of the Hamiltonian of the solid (2.12), if the ions are fixed at their equilibrium positions. It has to be augmented by spin-dependent terms if related eﬀects (e.g., spin-orbit coupling, Zeeman eﬀect) shall be considered (see Sects. 4.2, 4.3, and Chap. 5). U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 4,

76

4 The Free Electron Gas

Instead of this complex problem, for which only approximate solutions can be found with numerical methods, we want to consider in this chapter the much simpler problem of the homogeneous electron gas. It is characterized by neglecting the structural aspects and by replacing the configuration of pointlike positive ions by a homogeneous positive background charge to ensure charge neutrality of the system (Problem 4.1). This so-called jellium model (remember the diﬀerence between jelly and confiture) represents a many-body system of free charged fermions, which – replacing the Coulomb interaction by a more general two-particle interaction – applies also to physical systems beyond solid state physics. Thus, the treatment of the homogeneous electron gas, the calculation of its ground state energy per particle (which in diﬀerent approximations can be done even analytically) becomes an introduction to the concepts of many-body theory, which have their bearings also in nuclear or astrophysics.

4.1 Free Electrons Without Interaction In order to have a simple first look at the homogeneous electron gas, we want to neglect for the moment the interaction between the electrons (the second term in (4.1)) and the interaction with the positive background charge (corresponding to the third term in (4.1)) together with the fourth term, the electrostatic interaction energy of the ion configuration. This simplification is reasonable because in the jellium model these terms compensate exactly if one also considers the electrons as being smeared out just as the positive ions (see Problem 4.1). Thus instead of the Hamiltonian H of (4.1) we consider HN =

N

Hl ,

where

Hl =

l=1

p2l , 2m

(4.2)

with the Schroedinger equation HN Ψ({rl }) = E Ψ({rl }),

(4.3)

where Ψ depends on the configuration {rl } = {r1 , . . . r N } of the N electrons. With HN being a sum of single-particle operators, this N -particle problem can be separated by assuming Ψ({rl }) to be a product of single-particle wave functions N * Ψ({rl }) = Ψl (r l ). (4.4) l=1

We employ Ritz’ variational principle + , δ Ψ|HN |Ψ − ǫl (ψl |ψl − 1) = 0 l

(4.5)

4.1 Free Electrons Without Interaction

77

to minimize the expectation value of HN under the constraint of normalized single-particle wave functions ψl (r l ). The contraint is weighted with the Lagrangian parameter ǫl . Using (4.2) and (4.4) we formulate (4.5) as variational derivative with respect to ψl∗′ + , δ (ψl |Hl − ǫl |ψl + ǫl ) = 0 (4.6) δψl∗′ l

to obtain the single-particle problem Hl ψl (rl ) = ǫl ψl (r l ) ,

(4.7)

which is the same for all electrons. Note, that the Lagrangian parameter ǫl takes the role of the single-particle energy. The total energy of the N -electron system can be represented as the sum of the energies ǫl E =

N

ǫl

(4.8)

l=1

of the occupied single-particle states. The single-particle problem (4.7) with Hl from (4.2) is that of a free particle and has the solution 1 ψk (r) = √ eik·r V

(4.9)

with the eigenvalues ǫ(k) =

¯ 2 k2 h . 2m

(4.10)

The single-particle wave function is normalized to the fictitious (or crystal) volume V , which encloses the considered N electrons (particle in the box). Let V be a cuboid with side lengths Li , i = 1, 2, 3 and assume periodic boundary conditions (see problem 3.3), according to which ψk (r) is periodic with the Li . These conditions restrict the components of k to the discrete values (see (3.18)) 2π ki = ni , ni integer, i = 1, 2, 3 (4.11) Li as shown in Fig. 4.1. The periodic boundary conditions are a trick to obtain discrete states, which can be counted. It should be clear, however, that the relevant physical quantity, on which all observables of the system depend, is the single-particle density n = N/V (and not the particle number N ). It is kept constant for V → ∞. The wave vector k (or else the triple n1 , n2 , n3 of (4.11)) characterizes the single-particle state except for the spin degree of freedom. Each of the (quasidiscrete) k takes a volume (2π)3 /V . The state with lowest energy at T = 0 K

78

4 The Free Electron Gas

ky 2π L

kF

kx

2π L

Fig. 4.1. (kx , ky ) plane with discrete k values due to the periodic boundary conditions. In the ground state at T = 0 K all points up to the radius kF , defining the Fermi sphere, are occupied

for N electrons in the volume V , the ground state, is obtained by occupying the single-particle states with lowest possible energy with the constraint of Pauli’s principle1 which allows two electrons with opposite spin at each k. This is known as the ideal Fermi gas or Sommerfeld 2 model [109]. To evaluate the ground state energy we first make use of the energy dispersion (4.10), which is isotropic in k, to find the number of states in a spherical shell with radius k and thickness dk: D(k)dk =

2V 4πk 2 dk . (2π)3

(4.12)

The factor 2 is due to the spin degeneracy. The ground state is defined by filling concentric spherical shells around k = 0 with increasing radius k until the stock of N electrons is exhausted: N =

kF

V D(k)dk = 2 π

0

kF

k 2 dk =

V 3 k . 3π 2 F

(4.13)

0

Here kF denotes the radius of the sphere in k space, the Fermi sphere, within which at T = 0 K all states are occupied (Fig. 4.1). This radius is determined by the density n = N/V of the homogeneous electron gas: kF = (3π 2 n)1/3 ,

1 2

radius of the Fermi sphere .

Wolfgang Pauli 1900–1958, Nobel prize in physics 1945 Arnold Sommerfeld 1868–1951

(4.14)

4.1 Free Electrons Without Interaction

79

Particles with kF have the energy EF =

¯ 2 kF2 h , 2m

Fermi energy

(4.15)

which separates the occupied states from the empty ones. The ground state energy of the system containing N electrons follows to be E0 =

kF

3 ¯ 2 k2 h D(k)dk = N EF 2m 5

(4.16)

0

which, when divided by N , gives the average energy per particle in the ground state ǫ0 =

3 E0 = EF . N 5

(4.17)

With (4.15) and (4.14), one obtains the relation between ǫ0 and the particle density n h2 3 ¯ ǫ0 = (3π 2 n)2/3 . (4.18) 5 2m Frequently, a dimensionless density parameter rs is used. In the proper unit of length (Bohr’s radius aB = 4πε0 ¯ h2 /me2 ) it defines for a given density the radius of the Wigner–Seitz sphere whose volume can be allocated to each particle, V /N = 1/n = 4π(rs aB )3 /3. One finds easily kF =

9π 4

!1/3

1 rs aB

(4.19)

and (in the unit 1 Ry = me4 /2(4πε0 ¯ h)2 of energy) 3 ǫ0 = 5

9π 4

!2/3

1 2.2099 Ry = Ry. rs2 rs2

(4.20)

Corresponding to this relation, the mean energy per particle increases with increasing density to the extent by which states with larger k are occupied under the constraint of Pauli’s principle. Characteristic quantities related to kF or EF are the Fermi velocity vF = h ¯ kF /m, the Fermi wave length λF = 2π/kF , and the Fermi temperature TF = EF /kB , where kB is the Boltzmann constant. Supplement: Parameters of fermion systems The concepts developed so far allow one to distinguish fermion systems by their diﬀerent particle densities n (or rs ) and the derived quantities. They are given for a selection of solid state fermion systems in Table 4.1. Potassium (K) has one valence electron and crystallizes in the body centered cubic lattice with two atoms and thus two electrons in the elementary cube. Copper (Cu), also with one valence electron

80

4 The Free Electron Gas

A−1 , vF in 108 cm s−1 , EF in eV, TF in Table 4.1. Parameters (n in cm−3 , kF in ˚ 104 K) for diﬀerent fermion systems realized in solids System K Cu n-GaAs n-GaAs

n 22

1.43 · 10 8.53 · 1022 1017 1020

rs

kF

vF

EF

TF

4.83 2.66 1.29 0.129

0.75 1.36 0.014 0.143

0.87 1.57 0.25 2.5

2.14 7.05 0.75·10−3 0.078

2.48 8.18 0.0009 0.09

per atom, forms a face centered cubic lattice with four atoms and consequently A and four electrons in the elementary cube. Given the lattice constants aK = 5.31 ˚ A one easily obtains the electron density n and the other system paramaCu = 3.61 ˚ eters depending on n. The values given in Table 4.1 for K and Cu are characteristic for normal and noble metals, respectively. For comparison we also quote numbers for n-GaAs with two diﬀerent doping concentrations. In evaluating the electron density of the metals, we have only considered the valence electrons outside of closed shells. Similarly, in doped semiconductors we count only the extra electrons provided by electron donating impurity atoms (e.g., Si substituting Ga in GaAs) to determine n. The properties of these extra electrons deviate, however, from those of free electrons in vacuum. Their energy–momentum relation is modified by the periodic crystal potential and becomes a band structure (see Chap. 5), in which these extra electrons occupy states at the bottom of the conduction band with a dispersion similar to free particles but with an eﬀective mass m∗ diﬀerent from the free electron mass m; their charge is screened with the (relative) dielectric constant ε. Correspondingly, the system Hamiltonian would contain these modified electron parameters in the kinetic energy and in the electron– electron interaction. Therefore, the atomic units are redefined as eﬀective units a∗B = (εm/m∗ )aB and 1 Ry∗ = (m∗ /mε2 )Ry. In these units, one obtains for the two doping concentrations in n-GaAs with m∗ /m = 0.066 and ε = 12.9 the corresponding parameter values given in Table 4.1 . In spite of their much lower particle density the doped semiconductors appear as interacting fermion systems with much higher density (smaller rs ) than the metals. This is due to the much larger eﬀective Bohr radius, which is the length scale for the interaction. Fermion systems with quite diﬀerent densities are realized in neutron stars and systems of atoms confined in electromagnetic traps. Neutron stars have a radius of about 10 km and evolve from the gravitational collapse of small stars with about the mass of the sun. They can be described as an ideal Fermi gas with the extremely high density of 1038 cm−3 ! The other extreme of very low densities is obtained by cooling ensembles of fermionic atoms confined in electromagnetic traps to extremely low temperatures (TF < 1μ K), at which they represent degenerate Fermi systems (see Physics Today, October 2003, p. 18).

Knowing the single-particle energy-momentum relation ǫ(k) of (4.10), one can derive the spectral distribution of the electrons using (4.12): % dk 2m √ mV dE = EdE . (4.21) D(E)dE = D(k) 2 2 dE π ¯ h ¯h2

4.1 Free Electrons Without Interaction

81

The number of states with energy E, the electron density of states, given by !3/2 √ 2m V D(E) = E (4.22) 2 2 2π h ¯ is the well-known square root relation shown in Fig. 4.2. Alternatively, one can calculate the density of states also from D(E) = 2 k δ(E − ǫ(k)) (see for comparison the phonon density of states (3.53)), by collecting in k space all single-particle energies with ǫ(k) = E. The factor 2 takes into account the spin degeneracy. At T = 0 K, all states with energy smaller than EF are occupied and the ground state energy E0 for N particles is obtained by integrating ED(E) up to the Fermi energy EF . More generally, to include finite temperatures the ground state energy follows from E0 (T ) =

∞

ED(E)f (E, μ, T )dE

(4.23)

0

D(E)

3D

E µ(T>0)

EF = µ(T=0)

2D

D(E)

T=0 T>0 µ(T)=EF

E

Fig. 4.2. Density of states of free electrons in three (upper part) and two (lower part) dimensions. At T = 0 K all states with E ≤ EF are occupied. For finite temperatures the occupation around the Fermi energy is smeared out and is determined by the chemical potential μ(T )

82

4 The Free Electron Gas

with the Fermi–Dirac distribution function f (E, μ, T ) =

1 e(E−µ)/kB T

+1

.

(4.24)

Here the chemical potential μ is fixed (similar to kF in (4.13)) by the particle density n = N/V : N = V

∞

D(E)f (E, μ, T )dE .

(4.25)

0

Due to the interplay between density of states and distribution function the chemical potential depends in general on T (Problem 4.2). However, for a two-dimensional electron system with constant density of states (as depicted in the lower part of Fig. 4.2) this is not the case (provided the Fermi energy is suﬃciently high up in the band). At T = 0 K, f (E, μ, T ) degenerates to a step function with a step at EF and one identifies μ(T → 0) → EF (see Fig. 4.2). But for finite T in general, quantities such as the ground state energy or the particle density, which require to integrate some power function in E with the Fermi–Dirac distribution function, can be expressed in terms of Fermi integrals (see Appendix). Taking the T dependent chemical potential from Problem 4.2 (which applies for T → 0), the thermal energy density of free electrons at fixed volume is found (Sommerfeld expansion, see Appendix) ǫ(T ) = ǫ0 +

π2 (kB T )2 D(EF ) + . . . 6

(4.26)

with ǫ0 from (4.20). Its derivative with respect to T is the electron contribution to the specific heat (at constant V ) ! ∂ǫ(T ) π 2 kB T π2 2 kB T D(EF ) = cV (T ) = nkB . (4.27) = ∂T 3 2 EF V This linear dependence on T is characteristic for metals and can be detected experimentally at low temperatures. Taking into account the simultaneously present lattice contribution also(see Chap. 3), we write cV (T ) = γT + A T 3 , and find in a plot of cV (T )/T versus T 2 the Sommerfeld coeﬃcient γ (see Fig. 4.3). Sommerfeld parameters obtained from experimental data of the specific heat (γexp ) can be compared with those from the Sommerfeld model (γS ) calculated with the respective particle density (see Table 4.2). The comparison can be quantified by realizing that according to (4.27) the specific heat is proportional to the density of states at the Fermi energy and thus to the particle mass. Thus the ratio of the Sommerfeld parameters corresponds to the ratio of the fermion mass obtained from the measured specific heat (mcV ) to the free electron mass m. In Table 4.2 we find good agreement between the

4.1 Free Electrons Without Interaction

c v mJ T mol K 2

83

Au

5 4 3 2

Ag

1

Cu 1

4

9

(T/K) 2

Fig. 4.3. Measured values of cV /T for Au, Ag, Cu vs. temperature. The extrapolated value for T = 0 K gives the Sommerfeld coeﬃcient γexp

Table 4.2. Sommerfeld coeﬃcients from the free electron model (γS ) and from experiment (γexp ) (in mJ Mol K−2 ) and the corresponding mass ratio [29] System

γS

γexp

mcV /m

Na K Al Cu Ag Au Fe Ni

1.38 2.08 1.35 0.695 0.646 0.729 4.98 7.02

1.094 1.668 0.912 0.505 0.645 0.642 0.498 0.459

1.26 1.25 1.48 1.38 1.00 1.14 10.0 15.3

Sommerfeld model and the experimental data, i.e., mcV /m ≃ 1, for the alkali and noble metals, whose valence shells contain s and p electrons. However, even dramatic deviations, mcV /m ≫ 1, are observed for transition metals and rare earth compounds. As an example, the mass ratio of about 15 for Ni is given in Table 4.2. It can be ascribed to the influence of the periodic potential (Chap. 5) and understood as a single-particle or band structure eﬀect. However, mass ratios of the order of 103 found for the rare earth compounds are in addition due to electron–electron interaction or electronic correlation (see Sect. 7.5). The latter materials are called heavy fermion systems [110]. In this section, we have assumed a three-dimensional electron system in deriving e.g., the density of states in (4.12) and (4.22) and the parameters of the Fermi sphere. But electron systems of lower dimension (d = 2, 1, and 0) can also be realized at semiconductor hetero-interfaces, by using modern device technologies to produce quantum wires or dots, or in molecular systems like carbon nanotubes [21, 111–114]. These low-dimensional electron systems exhibit characteristic phenomena which can be understood in parts from the

84

4 The Free Electron Gas

dimension-specific density of states (see e.g., Fig. 4.2 for the two-dimensional case and Problem 4.3).

4.2 Free Electrons in a Magnetic Field The orbital motion of electrons and their energy spectrum change in a homogeneous magnetic field. At the same time, the magnetic field acts on the electron spin. Both eﬀects result in some of the magnetic properties of the electron system. For their description the single-particle Hamiltonian Hl in (4.2) is to be replaced by H =

1 e¯h (p + eA)2 + σ·B . 2m 2m

(4.28)

Here we denote the magnetic induction B = μ0 H, connected with the external homogeneous field by the vacuum permeability μ0 , as magnetic field. If its direction is along z, B = (0, 0, B), it can be described in the Landau gauge3 by the vector potential A = (0, Bx, 0). In a classical description, the first term in (4.28) leads to the circular motion as a consequence of the Lorentz force. It represents a harmonic oscillator with the cyclotron frequency ωc = eB/m, which in the quantum mechanical treatment gives the characteristic oscillator spectrum. The second term, containing the vector σ of the Pauli spin matrices, is the Zeeman4 term responsible for the removal of the spin degeneracy by the magnetic field. Let us calculate first the eigenvalues of H by considering the commutation relations for the components of the vector operator π = p + eA: [πx , πy ] = e(px Ay − Ay px ) = −i¯ heB,

[πz , πx ] = [πy , πz ] = 0 .

(4.29)

They can be written in the compact form of a vector relation π × π = −ie¯ hB .

(4.30)

Because of the translational invariance of H in the z-direction (parallel to the magnetic field) we have [pz , H] = 0 , and the linear momentum pz = h ¯ kz is a constant of motion. The Hamiltonian H =

1 ¯ 2 kz2 h (πx2 + πy2 ) + + μB σz B 2m 2m

(4.31)

(with the Bohr magneton μB = e¯ h/2m) can now be rewritten by making use of oscillator operators a, a† according to

3 4

Lev Davidovich Landau 1908–1968, Nobel prize in physics 1962 Pieter Zeeman 1865–1943, Nobel prize in physics 1902

4.2 Free Electrons in a Magnetic Field

i¯ h πy = − √ (a† − a), 2l

h ¯ πx = √ (a† + a), 2l

85

(4.32)

where a and a† fulfill the commutation relation

and l =

#

[a, a† ] = aa† − a† a = 1

(4.33)

(¯h/eB) is the magnetic length. Thus H takes the form ¯ 2 kz2 h 1 + μB σz B . H = h ¯ ωc (a† a + ) + 2 2m

(4.34)

Its eigenvalues are given by: ǫ(n, kz , ±) = h ¯ ωc

1 n+ 2

!

+

g ¯ 2 kz2 h ± μB B . 2m 2

(4.35)

In comparison with the magnetic field-free case (4.10), these single-particle energies diﬀer by replacing ¯h2 (kx2 + ky2 )/2m → ¯hωc (n + 1/2) and adding the Zeeman term. The first term describes the equidistant energy levels of the cyclotron oscillator (Landau levels), the second the free motion parallel to the magnetic field, and the third the Zeeman spin-splitting with the g factor, which for the free electron is (very close to) two. The spectrum is visualized in Fig. 4.4 . Experimentally, this spectrum can be detected by absorption of electromagnetic waves propagating parallel to the magnetic field, i.e., with an electric field vector in the (x, y) plane. From the resonant absorption at ω = ωc (cyclotron resonance) with

ε

n= 5 4 n= 5

4 3 2

3

1 0

2

1

0

B0

B

B=B0 kz

Fig. 4.4. Dependence of the single-particle energies ǫ(n, kz ) on the magnetic field at kz = 0 (left) and on kz at the magnetic field B0 (right). The Zeeman splitting is not shown

86

4 The Free Electron Gas

¯hωc = ǫ(n + 1, kz , ±) − ǫ(n, kz , ±) =

¯ eB h 2m

(4.36)

and the electron-spin resonance (also known as spin–flip or paramagnetic resonance) at ω = ωesr with hωesr = ǫ(n, kz , +) − ǫ(n, kz , −) = gμB B ¯

(4.37)

the system parameters m and g can be determined. This possibility is important because for electron systems in solids these parameter values diﬀer from those of free electrons in vacuum due to the influence of the periodic potential and become material specific (see Sect. 4.1 and Chap. 5). Thus, cyclotron and electron spin resonance allow to determine the material specific eﬀective mass m∗ and g factor g ∗ of free charge carriers in a solid. The drastic change of the free electron energy spectrum by the magnetic field results in a change of the density of states and of the electronic wave functions. While in the magnetic field-free case the eigenfunctions are plane waves with the wave vector k determined by the periodic boundary conditions, now we have oscillator functions with the quantum number n in the plane perpendicular to the magnetic field. A closer inspection shows that the Landau levels are highly degenerate with a degeneracy depending on the magnetic field. It can be obtained by starting from the Hamiltonian H (4.28) with the Landau gauge in the form H =

1 2 p + (py + eBx)2 + p2z + μB σz B , 2m x

(4.38)

satisfying the commutation relations [H, py ] = [H, pz ] = 0 . Thus, besides kz , ky is also a good quantum number. Making use of this fact we write H =

1 p2x ¯ 2 kz2 h + mωc2 (x − x0 )2 + + μB σz B . 2m 2 2m

(4.39)

Here x0 = h ¯ ky /mωc = l2 ky is the so-called center coordinate of the classical cyclotron orbit. It becomes quantized by applying periodic boundary conditions in y (as in Sect. 4.1) with ky = 2πny /Ly . Assuming in x direction a width Lx of the solid, the natural condition 0 < x0 < Lx for the center coordinate leads to 0 < ny <

Lx Ly Lx Ly eB. = 2πl2 2π¯h

(4.40)

This defines the degree of degeneracy, which is the maximum number of states (counted by ny ) in a Landau level at fixed kz for a given spin in a system with area Lx Ly perpendicular to the magnetic field B. As BLx Ly is the flux Φ of the magnetic induction threading the system area, the degeneracy can also be written as Lx Ly Φ eB = (4.41) 2π¯ h Φ0 where Φ0 = h/e is the elementary flux quantum.

4.2 Free Electrons in a Magnetic Field

87

For a two-dimensional system with N electrons in the plane perpendicular to the magnetic field, the ratio ν = N Φ0 /Φ gives the number of filled Landau levels (note that in this case there is no dispersion with kz ) and is called filling factor. Its inverse gives the number of elementary flux quanta per electron. As we shall see in Sects. 5.7 and 7.6, the integer (fractional) filling factor marks the characteristic features in the magneto-transport data connected with the integer (fractional) quantum Hall eﬀect. For three-dimensional systems one has to consider also the dispersion parallel to the magnetic field. By applying periodic boundary conditions to make kz countable, the number of states at kz is 2Lz Φ/2πV Φ0 = eB/2π 2 ¯h and leads to the energy dependent density of states eBV dkz . 2π 2 ¯ h dE

D(E, B) =

(4.42)

Solving (4.35) for kz as a function of the single-particle energy (here denoted by E) this can be expressed as D(E, B) =

eBV 2¯ 2π h n,σ=±1

2m h2 ¯

!1/2

1 2

E − ¯hωc (n +

− 21 1 σ + ) 2 2

(4.43)

or, using the cyclotron frequency ωc , as

D(E, B) =

V

√ hω c ¯ 8π 2

2m h2 ¯

!− 12 !3/2 E 1 σ . (4.44) − n+ + hω c ¯ 2 2 n,σ=±

The sum includes only contributions with positive radicands. For 3D electrons, the density of states D(E, B) (4.44) is a superposition of inverse square root functions resulting from the free particle motion parallel to the magnetic field (see the solution of Problem 4.3 for the one-dimensional case). The singularities of the inverse square-root functions mark the discrete energies of the Landau levels. They move with changing B and also their strength (related to the degeneracy) changes with B. This density of states diﬀers strongly from the case for B = 0 (see Fig. 4.2) and gives rise to the characteristic properties of the free electron system in a magnetic field. They will be discussed in the following paragraph. The Fermi energy (at T = 0 K) is given in analogy to (4.13) by the relation N =

EF (B)

D(E, B)dE 0

1 = 4π 2

2m hω c ¯ h2 ¯

!3/2 n,σ

EF (B) 1 σ − n+ + hωc ¯ 2 2

!1/2

,

(4.45)

88

4 The Free Electron Gas

D(E,B)

(a)

N/N0

(b)

1

0.7 1

2

EF (B)/ hωc

Fig. 4.5. (a) Density of states of free electrons with (solid line) and without (dashed line) magnetic field; (b) Ratio of electron numbers for occupying all states with and without magnetic field up to EF (B)

which, however, cannot be solved for EF (B). Therefore, we first compare with the case without magnetic field !3/2 EF 2m 1 N0 = (4.46) E D(E)dE = 2 F 3π ¯h2 0 by writing N 3 = N0 4

¯ ωc h EF

!3/2 n,σ

EF (B) 1 σ − n+ + hω c ¯ 2 2

!1/2

.

(4.47)

Let us now fill the densities of states D(E, B) and D(E) up to the same energy EF (B) and look for the ratio N/N0 (see Fig. 4.5) : For 0 < EF (B)/¯hωc < 1 we start with N/N0 > 1 but N/N0 decreases monotonously with increasing EF (B), finally falling below 1 (Fig. 4.5b). In the interval 1 < EF (B)/¯hωc < 2 the ratio N/N0 first increases strongly, reaches a maximum value > 1 and then for EF (B) → 2¯hωc falls again below the value 1. This behavior is repeated with increasing values of EF (B) and we find an oscillating function with a characteristic period in the dependence on 1/B given by ∆1/B = e¯ h/mEF (B) .

(4.48)

This oscillation is a consequence of the quantization into Landau levels. The Fermi energy EF (B) for the given number of electrons N = N0 is found by the

4.2 Free Electrons in a Magnetic Field

E F (B)

3 2

89

1

E F (0)

n=0 magnetic quantum limit

B Fig. 4.6. Energies of Landau levels and Fermi energy versus magnetic field (spinsplitting is not shown)

lowering (raising) of EF (B) relative to EF (0) in intervals in which, according to Fig. 4.5b, N/N0 > 1 (N/N0 < 1). The result is shown in Fig. 4.6 with EF (B) oscillating around EF (0) with the period ∆1/B . This behavior of the Fermi energy, resulting from the Landau quantization, leads to a characteristic oscillating dependence of solid state properties on the magnetic field, the quantum oscillations. By measuring this period the magnetic field dependent Fermi energy EF (B) and thus the particle density can be determined. With increasing magnetic field the oscillations become weaker (and smoothen out at finite temperature). Therefore, they can be detected only at suﬃciently high magnetic fields and low temperatures given by h ¯ ωc > kB T . For h ¯ ωc > EF (B), when only states in the lowest Landau level are occupied, we reach the so-called magnetic quantum limit (Problem 4.4). Let us draw attention to the magnetization as the basic material property in this context. Starting point for describing the magnetization is the Gibbs free energy F = μN + Ω, with the chemical potential μ, the particle number N , and the grand-canonical potential (β = 1/kB T ) 1 1 ln(1 + e−β(ǫα −µ) ) . Ω = − lnZG = − β β α

(4.49)

The sum is over the single-particle states with ǫα = ǫ(n, kz , σ) (see (4.35)). For its evaluation, we refer to the literature [115, 116]. The magnetization, which for the isotropic system of free electrons has only one nonvanishing component parallel to the external magnetic field, is obtained from the free energy F according to ! 1 dF Mz = − (4.50) V dB T,µ

90

4 The Free Electron Gas

and reads (with x = EF /μB B and y = πkB T /μB B) . ∞ sin(π/4 − nπx) 1 3 N μ2B (−1)n 1/2 √ cos(nπ) . B 1 − + y(x) Mz = 2 V EF 3 n sinh(nπy) n=1

(4.51)

Here we have replaced the chemical potential μ(B, T ) by its value at T = 0 K, briefly written as EF . In the context of linear response, the magnetization, expressed in the form Mz = χM B/μ0 , is determined by a response function, the magnetic susceptibility χM which (up to the vacuum permittivity μ0 ) is the derivative ! ∂Mz M χ = μ0 (4.52) ∂B T of Mz from (4.51). χM is an experimentally accessible function of B and T . The individual terms in (4.51) have the following physical meaning: •

Pauli spin paramagnetism: The first term, showing an increase with increasing magnetic field, is a paramagnetic contribution. It can be traced back to the spin-splitting due to the Zeeman term. Suppressing the Landau quantization, we obtain the picture of Fig. 4.7: The energy levels and therefore the density of states D↑↓ (E) of electrons with diﬀerent spin are shifted against each other by the Zeeman energy μB B. Filling the states with D↑↓ (E) up to EF , there are more electrons with spin down than with spin up. Assuming EF ≫ μB B one finds the resulting magnetization (Problem 4.5) 3 N μ2B Mzpara = B. (4.53) 2 V EF

E

µB B

D (E)

EF

D (E)

Fig. 4.7. Density of states for spin-up and spin-down electrons (Landau quantization is suppressed). The dashed lines show the situation for B = 0

4.2 Free Electrons in a Magnetic Field

91

As can be seen from Fig. 4.7, Mzpara depends on the density of states (or the electron mass) and on the g factor at the Fermi energy. In solids, we expect material specific deviations of the experimental values for the Pauli susceptibility χspin =

•

(4.54)

from the free electron value indicating the already mentioned modifications of these parameters due to the periodic potential and the electron–electron interaction. In fact such deviations are particularly strong, e.g., for heavy fermion systems, as was the case for the specific heat. Landau–Peierls5 diamagnetism: The second term is negative, acting against the external magnetic field according to the Lenz rule applied to the cyclotron motion of the electrons. The magnetic moment connected with this motion has a direction opposite to the external magnetic field and results in a diamagnetic contribution. For free electrons in a periodic potential, which can be described as particles with an eﬀective mass m∗ , the diamagnetic contribution is to be multiplied by a factor (m∗ /m)2 : Mzdia

•

2 dMzpara = D(EF )μ2B , dB 3

1 N μ2B B =− 2 V EF

m∗ m

!2

(4.55)

For mass ratios m∗ /m ≃ 1 (see Table 4.2), we expect comparable values of the para- and dia-magnetic contributions. But there are also systems with m∗ /m ≫ 1 due to orbital contributions of d and f electrons, for which the diamagnetic term dominates. de Haas–van Alphen6 eﬀect: The third term describes the already mentioned oscillating contribution, periodic in 1/B with a period determined by the Fermi energy, which is characteristic for Landau quantization. In fact, this behavior of the magnetic susceptibility is found in metals and known as de Haas–van Alphen eﬀect. It can be used to determine the Fermi energy or, more precisely, the parameters of the Fermi surface, which for solids can deviate from the spherical form (see Sect. 5.7). The sum in the oscillating contribution converges rapidly due to the denominator and usually it suﬃces to consider only the first term with n = 1: χosc (B) ≃ −μ0

3 N π 2 kB T 2 V B2

∗ !1/2 m∗ m∗ cos( π4 − mmπ x) . (4.56) x ) cos(π ∗ m m sinh( mmπ y)

Here we have considered again possible deviations of the eﬀective mass from the free electron mass. An example of de Haas–van Alphen oscillations measured for Cu is shown in Fig. 4.8. 5 6

Sir Rudolf Ernst Peierls 1907–1995 Wander Johannes de Haas 1878–1960, P.M. van Alphen 1906–1967

92

4 The Free Electron Gas

M

B Fig. 4.8. Oscillating part of the magnetization (de Haas–van Alphen oscillations) for Cu (after [117])

B,kz

k . k

∇k ε(k)

Fig. 4.9. Fermi sphere and extremal cross section perpendicular to the magnetic field to illustrate the cyclotron motion in k space

It should be noted that the Landau quantization also shows up in other magnetic field dependent electronic properties of solids. One important example represent the Shubnikov7 –de Haas oscillations in the magneto conductivity, which for two-dimensional electron systems at high magnetic fields and low temperature evolve into the quantum Hall eﬀect (see Sect. 5.6). All these quantum oscillations can be understood in a semi-classical description. It starts from the classical equation of motion of an electron under the Lorentz force p˙ = e(v × B) , (4.57) leading to the cyclotron motion. In semi-classical terms, we use p = h ¯ k and the group velocity v = ∇k ǫ(k)/¯ h to obtain the corresponding equation of motion in k space dk e = 2 ∇k ǫ(k) × B . dt ¯ h

(4.58)

While in real space the electrons perform the classical cyclotron motion, their momentum (or wave vector) moves along the contour defined by cutting the Fermi sphere with a plane perpendicular to the magnetic field (see Fig. 4.9), 7

Lev Vasiljevich Shubnikov 1901–1937

4.3 Occupation Number Representation for Electrons

93

which is the circle with radius kF . The period of this motion is obtained after separation of variables as ¯2 1 h Θ= e B

ǫ(k)=EF

dk , |∇k ǫ(k)⊥ |

(4.59)

where the denominator is the modulus of the gradient of ǫ(k) in the direction perpendicular to B. The contour integral equals the area of the extremal cross section of the Fermi sphere. As will be shown in Chap. 5, the relation between the period of these quantum oscillations (in 1/B) and extremal cross sections perpendicular to the magnetic field applies to Fermi surfaces of nonspherical shape as well and is used for their exploration.

4.3 Occupation Number Representation for Electrons In Sect. 4.1, we have introduced the many-particle wave function of the N electron system in the form of (4.4) as a product of single particle wave functions and considered the Pauli principle in the construction of the Fermi sphere only by avoiding double occupancies. However, the important property of a many-fermion wave function, which has to be antisymmetric with respect to interchanging two particles, is not fulfilled by this form. Therefore, we replace it now by an antisymmetrised product. For N = 2 the two-particle wave function is 1 Ψ(x1 , x2 ) = √ (ψα1 (x1 )ψα2 (x2 ) − ψα1 (x2 )ψα2 (x1 )) 2

(4.60)

with ψαi (xj ) being ortho-normalized single-particle wave functions for states characterized by a complete set of quantum numbers αi (for free electrons αi = ki σi ) and xj denotes space and spin variables. The corresponding generalization for N electrons is the Slater8 determinant (with i short for αi ) " " " ψ1 (x1 ) ψ1 (x2 ) · · · ψ1 (xN ) " " " 1 " ψ (x ) ψ2 (x2 ) · · · ψ2 (xN ) "" . (4.61) ΨSlater ({xj , j = 1 . . . N }) = √ "" 2 1 N ! " . . . . . . . . . . . . . . . . . . . . . . . . . . "" " ψN (x1 ) ψN (x2 ) · · · ψN (xN ) "

By construction, it is antisymmetric under the exchange of two rows or columns and thus equals zero if two rows, i.e., the corresponding α, are identical. Thus double occupancy is avoided. Calculations with the Slater determinant are very clumsy and a more elegant formulation, the occupation number or Fock9 representation, is used instead. 8 9

John Clarke Slater 1900–1976 Vladimir Fock 1898–1974

94

4 The Free Electron Gas

The N -electron state can be represented in the form similar to the one used for phonons in Sect. 3.3 |Ψ = |{nα , all α} ,

(4.62)

but here the occupancy nα of the single-particle state with quantum numbers α is restricted to 0 or 1 to account for the Pauli principle. The total number of electrons in the state |Ψ is given by N = α nα . N -particle states can be obtained in a systematic way from the fermion vacuum |0 = |{nα = 0, all α}

(4.63)

by successively applying N fermion creation operators |Ψ = c†α1 c†α2 . . . c†αN |0 .

(4.64)

In contrast to the case of the creation operators for phonons introduced in Sect. 3.3, now the order in which the operators are applied becomes important. The anti-symmetry of the N particle state is guaranteed by the anti-commutation relations for fermion creation and annihilation operators {c†α , cα′ } = c†α cα′ + cα′ c†α = δαα′

{cα , cα′ } =

{c†α , c†α′ }

(4.65)

= 0.

(4.66)

Let us first look at the 2-electron state which can be generated by applying two creation operators to the fermion vacuum: according to the commutation relation we have |Ψ = c†α1 c†α2 |0

but

c†α2 c†α1 |0 = −|Ψ .

(4.67)

The adopted order of the applied fermion operators (with α1 = α2 ) determines the result obtained when applying cα1 or cα2 . One finds cα1 |Ψ = cα1 c†α1 c†α2 |0 = (1 − c†α1 cα1 )c†α2 |0 = c†α2 |0,

(4.68)

but cα2 |Ψ = −c†α1 |0 , because in the first step cα2 has to be interchanged with c†α1 which results in a sign change. This example for N = 2 demonstrates the influence of the order in which the fermion operators are applied to the vacuum state to create |Ψ. In general, for an N -electron state |Ψ = c†α1 . . . c†αN |0 = |{nα , all α; N = α nα } we have cαi |Ψ = (−1)νi nαi | . . . , nαi − 1, . . . ,

νi =

c†αi |Ψ = (−1)νi (1 − nαi )| . . . , nαi + 1, . . . ,

i−1

nαk

(4.69)

k=1

(4.70)

where the dots indicate unchanged entries. Note that the sum of the occupation numbers in the first equation is N − 1, because one fermion of |Ψ is

4.3 Occupation Number Representation for Electrons

95

annihilated, while for the second equation it is N + 1 due to the addition of one particle to |Ψ. The meaning of occupation number representation can be also illustrated by applying c†αi cαi to the N -particle state |Ψ. We find c†αi cαi |Ψ = (−1)νi nαi c†αi | . . . nαi − 1 . . . 2νi

= (−1) nαi (1 − (nαi − 1))| . . . nαi . . . |Ψ nαi = 1 = 0 nαi = 0

(4.71) (4.72) (4.73)

or in short c†αi cαi |Ψ = nαi |Ψ .

(4.74)

c†αi cαi

is the occupation number nαi of the singleThus the eigenvalue of particle state with quantum numbers αi in |Ψ and ˆ = c†α cα = n ˆα (4.75) N α

α

ˆ |Ψ = N |Ψ. is the number operator with the eigenvalue equation N The occupation number or Fock representation allows one to formulate the system Hamiltonian in terms of creation and annihilation operators. We give here only the rules for how the one-(Aˆ1 ) and two-particle (Aˆ2 ) terms of the Hamiltonian (or any other operator) have to be rewritten and refer for details to the Appendix (A.6): (4.76) ψαi |Aˆ1 |ψαj c†αi cαj Aˆ1 → ij

Aˆ2 →

ijkl

ψα(1) ψα(2) |Aˆ2 |ψα(1) ψα(2) c†αi c†αj cαk cαl . i j l k

(4.77)

These expressions have an obvious meaning: The one-particle operator (depending on the coordinates of only one particle) replaces a particle with quantum numbers αj by a particle with αi , the two-particle or interaction operator removes first a particle in the state αl and then another one in the state αk and adds particles in the states αj and αi . The weight of the operators are matrix elements of the original operators taken with the singleparticle wave functions ψα of an orthonormal set. The upper index (1) and (2) in the matrix element of the two-particle operator indicates the pairs of wave functions which have the same coordinates when integrating in position representation. This concept is now applied to the system of free electrons for which {c†kσ , ck′ σ′ } = δkk′ δσσ′

(4.78)

96

4 The Free Electron Gas

and ˆ = N

1 ψkσ (x) = √ eik·r χσ (s) , V

n ˆ kσ = c†kσ ckσ ,

n ˆ kσ ,

kσ

(4.79)

where χσ (s) is a Pauli spinor and s denotes the spin variable. The Hamiltonian for the jellium model of free electrons (with the jellium term H+ from Problem 4.1) Hjell =

p2 e2 1 l + + H+ , 2m 8πε0 k,l |r k − rl | l

(4.80)

k=l

can be rewritten (by making use of (4.76) and (4.77)) with 2 ¯ p2 h2 k † l → c ckσ 2m 2m kσ l

(4.81)

kσ

and k,l k=l

e2 e2 → ψk1 σ1 ψk2 σ2 | |ψk′1 σ1′ ψk′2 σ2′ (4.82) 4πε0 |rk − r l | 4πε0 |r 1 − r2 | = δσ1 σ1′ δσ2 σ2′

vk δk1 ,k′1 +k δk2 ,k′2 −k

(4.83)

k

where vk = e2 /ε0 V k 2 is the Fourier transform of the Coulomb potential (see Appendix). With the replacements σ1 = σ1′ = σ, σ2 = σ2′ = σ ′ and k′1 → p, k′2 → q we find Hjell =

¯ h2 k 2 kσ

2m

c†kσ ckσ +

1 vk c†p+kσ c†q−kσ′ cqσ′ cpσ + H+ . 2 p,q,k

(4.84)

σσ′

The interaction term becomes divergent for k → 0. However, this divergence is exactly compensated by the jellium term for which we obtained as solution of Problem 4.1 H+ = − 21 v0 N 2 while 21 p,qσσ′ v0 c†pσ c†qσ′ cqσ′ cpσ = 1 ˆ pσ qσ′ n ˆ qσ′ = 21 v0 N 2 . Thus we arrive at the Hamiltonian for free pσ n 2 v0 electrons in the jellium model Hjell =

¯ h2 k 2 kσ

2m

c†kσ ckσ +

1 vk c†p+kσ c†q−kσ′ cqσ′ cpσ . 2 p,q,k=0

(4.85)

σσ′

Its first term accounts for the single-particle energy of free noninteracting electrons, while the second term describes the electron–electron interaction. The interaction term is visualized in Fig. 4.10 as a Feynman diagram with the solid lines representing incoming and outgoing electrons with their respective

4.4 Hartree–Fock Approximation

q-kσ ¢

97

>

>

p+kσ

>

>

k pσ

qσ ¢

Fig. 4.10. Feynman diagram of the electron–electron interaction. The solid lines represent propagating electrons, the dashed line the Coulomb potential

wave vector and spin and the dashed line symbolizes the Coulomb interaction. This Hamiltonian does not only apply to electron systems in vacuum, but also to those realized in matter (see Table 4.1), however,by replacing the free electron mass by an eﬀective mass and screening the Coulomb interaction with the dielectric constant of the matter. Moreover it applies also to electron systems in energy bands, however, with a matrix element between full Bloch states (see Chaps. 5 to 7). In the following sections we develop the concepts for approximate solutions of the ground state problem for the jellium model.

4.4 Hartree–Fock Approximation In Sect. 4.1, we have characterized the ground state of the noninteracting electron system (at T = 0 K) as the filled Fermi sphere. The corresponding N -particle wave function, that takes into account the Pauli principle, is a Slater determinant composed of the wave functions of occupied single-particle states, i.e., plane waves with |k| ≤ kF but it can also be obtained by applying creation operators to the fermion vacuum to fill the Fermi sphere. Let us denote this ground state by |Ψ0 . We may now calculate the expectation value of Hjell from (4.85) with |Ψ0 giving the ground state energy in Hartree10 –Fock (HF) approximation E0HF = Ψ0 |Hjell |Ψ0 =

¯ 2 k2 † h c ckσ |Ψ0 2m kσ kσ 1 + Ψ0 |vk c†p+kσ c†q−kσ′ cqσ′ cpσ |Ψ0 . (4.86) 2 k=0

Ψ0 |

q,p,σσ′

The first term reproduces the ground state energy 35 N EF of the noninteracting electron gas obtained in the Sommerfeld model (Sect. 4.1). The second term is the first order correction due to the Coulomb interaction, which is treated 10

Douglas Rayner Hartree 1897–1958

98

4 The Free Electron Gas

q

p

q p kF

kF

direct

exchange

Fig. 4.11. Illustration of the direct and exchange terms in the Fermi sphere

here as a perturbation. Before we discuss the quality of this approximation, this correction has to be calculated. Two contributions, depicted in Fig. 4.11 can be distinguished: 1. Proceeding from right to left the two annihilation operators remove the electrons with pσ and qσ ′ from the filled Fermi sphere (|Ψ0 ) and the two creation operators add electrons at p + kσ and q − kσ ′ , respectively. In order to obtain the filled Fermi sphere (or Ψ0 |) k has to be zero. However, this term is excluded from the sum over k in the interaction term as a result of the jellium model and hence does not modify the ground state energy of the Sommerfeld model. It should be mentioned, however, that this so-called direct term describes the electrostatic interaction energy for a system with charge density n (see Problem 4.1), which becomes important for inhomogenous electron systems (see Chaps. 5 and 6) and defines the Hartree approximation. 2. As before, the two annihilation operators remove two electrons from the Fermi sphere, while the two creation operators put them back but now with interchanged wave vectors. This exchange within the Fermi sphere is possible only for electrons having the same spin (σ = σ ′ ) and whose wave vectors diﬀer by k = q − p (see Fig. 4.11). This exchange term results from the Pauli principle and does not have a classical analogue (in contrast with the direct term). The Hartree–Fock approximation is also understood to have considered the exchange contribution. The exchange term is evaluated under the constraints of the exchange process (k = q − p, σ = σ ′ ) giving E0HF = E0H +

σ,q,p q=p

e2 Ψ0 |c†qσ c†pσ cqσ cpσ |Ψ0 2ε0 V |q − p|2

(4.87)

with E0H being the ground state energy in the Hartree approximation. Because of p = q the operators, c†pσ and cqσ can be interchanged with a sign change and the four-operator expression becomes a product of number operators or their respective eigenvalues, the occupation factors nqσ , npσ . Thus one finds

4.4 Hartree–Fock Approximation

E0HF = E0H −

σ,q,p q=p

e2 nqσ npσ . 2ε0 V |q − p|2

99

(4.88)

Taking E0H from (4.16) and after summing over σ and considering the occupation factors, gives E0HF =

3 N EF − 5

q,p,q=p |q|,|p|≤kF

e2 . ε0 V |q − p|2

(4.89)

The evaluation of the sums is explicitly carried out in the Appendix (A.5). It gives 3 N e2 kF . (4.90) Eexch = − 16π 2 ε0 and leads to the important result of this calculation, the ground state energy of the electron gas in HF approximation E0HF =

h2 kF2 3 3 ¯ N − N e2 kF . 5 2m 16π 2 ε0

(4.91)

Due to the negative contribution of the exchange interaction it is lower than the result of the Sommerfeld model. In the formulation with the dimensionless density parameter rs defined in Sect. 4.1 it reads + !2/3 !1/3 , 1 2 3 9π 3 9π HF E0 = N Ry. (4.92) − 2 5 4 rs 4π 4 rs As mean value energy per electron the result is written in the form 1 HF 2.2099 0.916 E Ry. ǫHF = = − 0 N 0 rs2 rs

(4.93)

It is expressed in terms of inverse powers of the density parameter rs , which for the same electron density n can have quite diﬀerent values depending on the material (see Table 4.1) in which it is realized (Problem 4.6). In order to go beyond the HF approximation (which is a first order perturbation calculation) one may consider the second order correction becausse of the Coulomb interaction. This is formulated in Problem 4.7. As it will turn out, this term is logarithmically divergent. Thus higher order perturbation calculation seems not to be the proper way to treat Coulomb interaction and special many-body techniques are required to improve the result. But we can take from (4.93) the message that the HF approximation works well in the high density limit (rs ≪ 1). At this point, two considerations about the HF approximation can be made. The first one aims at an approximate formulation of Hjell (4.85) as an eﬀective single-particle operator

100

4 The Free Electron Gas

Hjell ≃

† ǫHF kσ ckσ ckσ

(4.94)

kσ

with single-particle energies ǫHF kσ , which besides the kinetic energy of free particles contain a contribution from the Coulomb interaction according to the HF approximation. A systematic way to calculate ǫHF kσ is the formulation of the equation of motion for the time-dependent fermion operator 0 d † 1 / ckσ = Hjell , c†kσ . dt i¯ h

(4.95)

d † 1 HF † c = ǫ c dt kσ i¯ h kσ kσ

(4.96)

The commutator of ckσ with the interaction term generates terms consisting of three fermion operators. Here the HF approximation means to replace these terms by the product of an expectation value of the number operator and the remaining fermion operator. This classifies the HF approximation as a mean field approximation, see Sect. 6.5. The result takes the form

with the single-particle energy ǫHF kσ =

e2 ¯ 2 k2 h − 2m ε0 V

q=k |q|≤kF

1 . |q − k|2

(4.97)

The evaluation of the sum over q in the Fermi sphere can be performed as an integral in polar coordinates with ϑ being the angle between q and k

qk |q|≤kF

1 V = 2π 2 |q − k| (2π)3 V = (2π)2

kF

0

+1

−1

kF

dq 0

dq q 2

q2

d cos ϑ + k 2 − 2kq cos ϑ

(q − k)2 q ln . 2k (q + k)2

(4.98)

The last integral can be found in an integral table. Finally, we obtain for the single-particle energy in HF approximation ! e2 kF kF2 − k 2 "" kF + k "" h2 k 2 ¯ − . (4.99) 1 + ǫHF = ln " " kσ 2m 4π 2 ε0 2kkF kF − k

A plot of ǫHF kσ is shown in Fig. 4.12 for a particular choice of the electron density n in reduced units together with the free electron dispersion. As for the ground state energy, we find for the single-particle energies in the HF approximation a lowering with respect to the free particle energy. Taking the sum over these single particle energies in the Fermi sphere leads back to the result of (4.93). Replacing the Hamiltonian of the jellium model (4.85) as indicated in (4.94) corresponds to describing the interacting free electrons as

4.4 Hartree–Fock Approximation

101

ε/ε°F 2 1 0

0.5

1

1.5

(k/kF)

–1 Band width –2

Fig. 4.12. Single-particle HF energy (solid line) for rs = 4 in comparison with free particle energy (dashed line), after [12]

noninteracting quasi-particles, whose single-particle energies incorporate part of the Coulomb interaction. This quasi-particle concept is frequently used in many-body theory. The second consideration concerns the charge density connected with the exchange interaction. For this purpose, we may rewrite the exchange correction with the Fourier transform of the Coulomb interaction as i(q−k)·r e2 1 e e2 1 − d3 r = − 2 ε0 V |q − k| 4πε0 q=k V r q=k |q|≤kF

= −

|q|≤kF

eρHF k (r) 3 4πε0 r

d r.

(4.100)

The last expression corresponds to the potential energy of an electron at r = 0 in a charge density ρHF k (r) (the HF or exchange charge density) resulting from all electrons in the Fermi sphere having the same spin: e i(q−k)·r e ρHF eiq·r . (4.101) e = e−ik·r k (r) = V q=k V q=k |q|≤kF

|q|≤kF

By performing the sum over q as integral over the Fermi sphere in polar coordinates one obtains ρHF k (r) =

2e e−ik·r (sin (kF r) − kF r cos (kF r)) . (2π)2 r3

(4.102)

The mean value of ρHF k (r) over the Fermi sphere gives the averaged HF charge density

102

4 The Free Electron Gas

g(r) 1 1/2 Fermi or exchange hole

0

1

2

3 k r 4 F

Fig. 4.13. Pair–distribution function g(r) = 1 − ρHF (r)/en showing the depression in the vicinity of r = 0 known as Fermi or exchange hole

ρHF (r) =

2 HF ρk (r) N k |k|≤kF

e 9N 2 (sin(kF r) − kF r cos(kF r)) . = 2V (kF r)6

(4.103)

The limits ρHF (r → ∞) = 0

and ρHF (r → 0) =

1 en 2

(4.104)

indicate the modification of the constant charge density −en of the noninteracting electron gas due to the exchange correction. In the vicinity of an electron with given spin, the charge density is reduced by 50% corresponding to the interaction with all electrons having the same spin. With increasing r, the density approaches the value of the noninteracting electron system. This result can be expressed in terms of the pair–distribution function (see Sect. 1.3) g(r) = 1 − ρHF (r)/en (4.105) for free electrons in the HF approximation, which is depicted in Fig. 4.13. It shows a depression around r = 0 which is the Fermi or exchange hole and visualizes the correlation due to the exchange interaction or the Pauli principle.

4.5 The Dielectric Function The dielectric function is the response of the system (here the homogeneous electron system) to a perturbation by an electromagnetic field. It is convenient to consider this perturbation as having the characteristic space and

4.5 The Dielectric Function

103

time dependence given by exp(i(q · r − ωt)) typical for a Fourier component with q and ω. One can distinguish this perturbation caused by a moving charge (as in a particle scattering process) from that of an electromagnetic wave. The latter case, leading to the transverse dielectric function has been the subject of Sect. 3.5. Here we consider the former case: The moving charge is connected with a longitudinal field (directed parallel to the momentum of the moving particle) and the response is the longitudinal dielectric function ε(q, ω). For ω = 0 this case includes also the response of the electron system to a static charge placed into the electron system: The electrons will arrange around this static charge due to Coulomb attraction (if the static charge is positive) or repulsion (if it is negative), thus increasing or reducing the otherwise homogeneous density. This eﬀect is known as static screening. In Sect. 2.5, the inverse dielectric function has been introduced already in an exact formulation for an arbitrary system as density–density correlation function between density and number fluctuations. Here we want to derive the inverse dielectric function for the homogeneous electron gas. Starting point is the electron system described by Hjell and the external perturbation by Vext H = Hjell + Vext .

(4.106)

Having in mind the results of Sect. 2.5, it is advantageous to †use the forˆk = mulation of Hjell in terms of number fluctuations N pσ cp+kσ cpσ (see Problem 4.8) Hjell =

¯ h2 k 2 kσ

2m

c†kσ ckσ +

1 ˆ† ˆ vk Nk Nk − N 2

(4.107)

k=0

and write the perturbation as ˆ−q Next e−iωt Vext = −vq Next eiq·r−iωt = −vq N

(4.108)

ˆ−q . where we have replaced exp(iq · r) by N The response of the homogeneous electron system to a perturbation by an external charge will be a deviation from the homogeneous density in terms of number (or density) fluctuations. For Vext as given in (4.108) the induced number fluctuations are described by ˆq eiq·r−iωt . N (r, t) = N

(4.109)

In the context of Poisson’s equation, the external test charge eNext and the induced charge (density) fluctuation eN (r, t) play the role of free and polarization charges, respectively, or in other words: eNext/V is the source of the dielectric displacement field D while both eNext and eN (r, t) determine the electric field E, both fields being connected by the dielectric function: D(q, ω) = ε0 ε(q, ω)E(q, ω). Thus, as in Sect. 2.5 we may write

104

4 The Free Electron Gas

e ˆq Next + N V ε0 e iq · D(q, ω) = Next V iq · E(q, ω) =

(4.110) (4.111)

and obtain from these equations the relation ˆq 1 N =1+ . ε(q, ω) Next

(4.112)

ˆ−q and N ˆq as the operators B ˆ and A, ˆ From (4.108) and (4.109) we identify N respectively, of the response formalism, to write the exact expression for the inverse longitudinal dielectric constant as (see (2.80)) 1 i ∞ ˆq (τ ), N ˆ−q (0)]0 dτ eiωτ −Γ τ [N (4.113) = 1 + lim vq Γ →0 ε(q, ω) h 0 ¯ where the damping factor regularizes the integral. The thermal expectation value under the integral is to be evaluated with the eigenstates of the Hamiltonian Hjell of the unperturbed system. As these are not known, this can be done only approximately. For T = 0 K the thermal expectation value reduces to the expectation value of the ground state |Ψ0 (see Problem 2.2) and we may write & ˆq (τ ), N ˆ−q (0)]|Ψ0 = ˆq (τ )|Ψm Ψm |N ˆ−q (0)|Ψ0 Ψ0 |[N Ψ0 |N m

' ˆ−q (0)|Ψm Ψm |N ˆq (τ )|Ψ0 . (4.114) − Ψ0 |N

The matrix elements can be evaluated by making use of ˆq (τ ) = e h¯i Hjell τ N ˆq e− h¯i Hjell τ N to yield ¯hωm0 = Em − E0 ˆq (τ ), N ˆ−q (0)]|Ψ0 = Ψ0 |[N

m

ˆ−q = N ˆq† and N

(4.115)

( ) ˆq |Ψm |2 e−iωm0 τ − eiωm0 τ . (4.116) |Ψ0 |N

Thus we obtain the exact expression for the longitudinal dielectric function at T = 0 K 1 vq 1 ˆq |Ψm |2 = 1 + lim |Ψ0 |N Γ →0 ¯ ε(q, ω) h m ωm0 + ω + iΓ 1 . (4.117) + ωm0 − ω − iΓ Formally, this result is the same as for χ(ω) derived in Sect. 3.5 for a transverse perturbation, where however the matrix elements are those of the dipole operator while here they are those of the operator of number fluctuations.

4.5 The Dielectric Function

105

The inverse dielectric function has poles at ω = ±ωm0 − iΓ in the lower complex ω-plane (see Fig. 3.5), thus measurements of 1/ε(q, ω) provide information about the exact excitation energies h ¯ ωm0 = Em − E0 . Such experiments are performed as inelastic scattering of charged particles (electrons), in which energy and momentum is transferred to the electron system. This experimental technique is known as energy–loss spectroscopy. Let us evaluate the inverse dielectric function at T = 0 K in HF approximation. This is done by calculating the matrix elements of the number fluctuations between the HF ground state (which is the filled Fermi sphere) and excited states, which are obtained by removing an electron from the Fermi sphere and placing it into a state outside (see Fig. 4.14). In other words, the excited states are the Fermi sphere plus an electron–hole excitation. The ˆ q = c† matrix elements can be evaluated and yield N kσ k+qσ ckσ ˆq′ |Ψ0 = Ψkqσ |N

|k| ≤ kF , |k + q| > kF otherwise

δqq′ 0

(4.118)

and one finds (with h ¯ Γ = δ) 1 = 1 + lim vq HF δ→0 ε (q, ω) −

|k|≤kF ,σ |k+q|>kF

1 hω + ǫk+q − ǫk + iδ ¯

1 hω − ǫk+q + ǫk + iδ ¯

.

(4.119)

Here the energies ǫk ought to be the HF single particle energies, but can be replaced by h ¯ 2 k 2 /2m which is a good approximation for rs ≪ 1. The restricted sum can be considered by using the Fermi distribution function fk and if in addition we replace in the first term k + q → −k the rhs of (4.119) reads fk (1 − fk+q ) f−k−q (1 − f−k ) − . (4.120) = 1 + lim vq δ→0 hω + ǫ−k − ǫ−k−q + iδ ¯hω − ǫk+q + ǫk + iδ ¯ k,σ

k+q,σ

kF

k,σ

Fig. 4.14. Schematic view of an excited state: Fermi sphere plus electron–hole pair

106

4 The Free Electron Gas

Finally with f−k = fk , ǫk = ǫ−k , and fk fk+q = 0 (one of the states is not occupied) one obtains 1 = 1 + vq π0 (q, ω) εHF (q, ω)

(4.121)

where π0 (q, ω) = lim

δ→0

k,σ

fk − fk+q . hω − ǫk+q + ǫk + iδ ¯

(4.122)

is the polarization function of the free noninteracting electron system. A standard approximation for the inverse longitudinal dielectric function beyond HF is the so-called random phase approximation or RPA. Although its derivation in the original literature (where also the name is justified) is quite involved (see [5]) it can be obtained simply by replacing the bare Coulomb interaction vq by the screened one vq /ε(q, ω). This gives rise to the following formulation: 1 vq = 1+ π0 (q, ω) ε(q, ω) ε(q, ω) = 1 + vq π0 (q, ω) 1 + vq π0 (q, ω)(1 + vq π0 (q, ω)(. . .)) 2 = 1 + vq π0 (q, ω) + vq π0 (q, ω) + . . . 1 . (4.123) = 1 − vq π0 (q, ω) The structure of this expression, reminding of the Born series known from scattering theory, is typical for results obtained for interacting particles and can be cast into a graphic representation in terms of Feynman diagrams [4, 5, 64]. Here we give only the formula for the dielectric function in RPA, also known as the Lindhard function, εRPA (q, ω) = 1 − vq π0 (q, ω) = 1 − vq lim δ→0

k,σ

fk − fk+q . hω − ǫk+q + ǫk + iδ ¯

(4.124)

It should be mentioned, that this result which goes beyond the HF approximation, is not obtained by a more accurate description of the eigenstates rather than by the physical argument of screening applied to the electron–electron interaction in the system. The same result can be obtained in an alternative way known as the self-consistent field approximation. The idea here is to consider the induced number density fluctuations together with the external charge as the perturbation of the system but evaluate the response with the HF ground state. This calculation is the subject of Problem 4.9.

4.6 Discussion of the Dielectric Function

107

4.6 Discussion of the Dielectric Function Starting from (4.117), the exact expression for the inverse longitudinal dielectric function at T = 0 K can be separated with ! 1 1 lim =P ∓ iπδ(x) (4.125) δ→0 x ± iδ x into its real and imaginary part: 2¯hωm0 1 ˆq |Ψm |2 = 1 + vq (4.126) |Ψ0 |N 2 ε(q, ω) (¯hω) − (¯ hωm0 )2 m 1 ˆq |Ψm |2 {δ(¯ = −πvq Im |Ψ0 |N hω − ¯hωm0 ) − δ(¯hω + h ¯ ωm0 )} . ε(q, ω) m Re

(4.127)

In Sect. 1.3 we have introduced the dynamic structure factor ∞ 1 ˆq (t)N ˆ−q (0)dt e−iωt N S(q, ω) = 2π −∞

(4.128)

ˆq (t), introducwhich for T = 0 K (after resolving the Heisenberg picture for N ing a complete set of exact eigenstates, and performing the integration over t) takes the form ˆq |Ψm |2 δ(ω − ωm0 ). S(q, ω) = |Ψ0 |N (4.129) m

Thus it is possible to express the imaginary part of the dielectric function as Im

1 π = − vq S(q, ω) − S(q, −ω) . ε(q, ω) h ¯

(4.130)

Moreover, the diﬀerential cross section for inelastic scattering is related with the dynamic structure factor: d2 σ k ′ m 2 = |vq |2 S(q, ω). dΩdω k 2π¯ h

(4.131)

This relation tells us how to extract information about the excitations in the interacting electron system from inelastic scattering experiments: The crosssection will be enhanced if the experimental settings, q and ω, correspond to the energy and momentum of an excitation in the electron system. The structure factor or likewise the dielectric function can be expressed in terms of the characteristic excitation energies of the system. Due to this relation Im(1/ε(q, ω) is known also as the energy–loss function. In order to discuss this in more detail, we look at the dielectric function in RPA εRPA (q, ω) = ε1 (q, ω) + iε2 (q, ω) with

108

4 The Free Electron Gas

ε1 (q, ω) = 1 + vq

kσ

ε2 (q, ω) = πvq

kσ

fk+q − fk hω − ǫk+q + ǫk ¯

(4.132)

(fk − fk+q ) δ(¯hω − ǫk+q + ǫk ) .

(4.133)

The energy–loss function can be written in terms of ε1 (q, ω) and ε2 (q, ω) as Im

ε2 (q, ω) 1 =− 2 2 . ε(q, ω) (ε1 (q, ω)) + (ε2 (q, ω))

(4.134)

The numerator ε2 (q, ω) gives nonvanishing contributions only for ¯2 h (2k + q) · q (4.135) 2m which in view of the particle–hole excitations out of the Fermi sphere is possible at h ¯ ω = 0 for all q with 0 ≤ |q| ≤ 2kF and for h ¯ ω > 0 for the same range of |q| but shifted to higher values. This excitation spectrum is known as the particle–hole continuum (see Fig. 4.15). Significant contributions to the energy-loss function are also expected if the denominator in (4.134) vanishes, i.e., for ε1 (q, ω) = ε2 (q, ω) = 0. Let us consider the case ¯hω ≫ ǫk+q − ǫk for which ε2 (q, ω) = 0 and (after replacing k + q in the first term of ε1 (q, ω) by k) fk fk ε1 (q, ω) = 1 + vq − hω − ǫk + ǫk−q ¯ ¯hω − ǫk+q + ǫk kσ 2ǫk − ǫk+q − ǫk−q . (4.136) fk = 1 + vq (¯hω − ǫk + ǫk−q )(¯ hω − ǫk+q + ǫk ) hω = ǫk+q − ǫk = ¯

kσ

2 2

Using ǫk = h ¯ k /2m, the numerator can be rewritten as −¯h2 q 2 /m, and simplifying the denominator for h ¯ ω ≫ ǫk+q − ǫk , one obtains for suﬃciently small |q|

ω(q) plasmon particle - hole continuum

1

q/2kF

Fig. 4.15. The spectrum of excitations derived from the energy–loss function: RPA spectrum

4.6 Discussion of the Dielectric Function

ε1 (q, ω) ≃ 1 − vq

¯ 2 q2 1 h N e2 fk = 1 − 2 m (¯hω) ε0 V mω 2

109

(4.137)

k

or for q → 0 ǫ1 (q, ω) ≃ 1 −

ωp2 , ω2

ωp2 =

e2 N . ε0 m V

(4.138)

Apparently, the energy–loss function has a singularity at the plasma frequency ωp . As the typical result of RPA it represents a collective excitation, the plasmon, in which all particles of the electron system participate. Moreover, it obeys a sum rule, which can be expressed in terms of ωp (Problem 4.10). We note that (4.138) does not contain h ¯ . This indicates the possibility of an interpretation in terms of classical physics. In fact, the classical equation of motion of a single electron in an electric field E, determined by the dielectric polarization E = −P /ε0 = ner/ε0 of the electron system with density n = N/V , reads e2 N ¨ = −ωp2 r . r or r m¨ r=− (4.139) ε0 V In the long wavelength limit each electron of the system experiences a restoring force characterized by the plasma frequency when (in the collision with a fast charged particle) the system is displaced against the jellium background. For finite q, one obtains from ε1 (q, ω) = 0 ωp2 (q)

=

ωp2

! h2 kF2 3 q2 ¯ 1+ + ... , 10 mωp2

(4.140)

where the second and higher order terms indicate quantum mechanical corrections to the classical result. The typical RPA spectrum, consisting of the particle–hole continuum and the plasmon mode, as depicted schematically in Fig. 4.15, can be translated for small momentum transfer into the real and imaginary part of the dielectric function as shown in Fig. 4.16: ε2 (q, ω) is determined by the particle–hole continuum giving contributions only at low frequencies, while ε1 (q, ω), showing a more complex frequency dependence, starts at ω = 0 with the value defined by the Thomas–Fermi screening parameter kFT (see Problem 4.11), changes sign around the upper cut-oﬀ frequency of the particle–hole excitations, passes through zero at the plasma frequency, and approaches 1 for high frequencies. These characteristic features are also found in the experimental data, e.g., [5, 89]. In metals with electron densities of the order of 1023 cm−3 (see Table 4.1), the plasmon energy h ¯ ωp is about 1 Ry and determines the response in an energy-loss experiment in this energy range. In contrast, doped semiconductors represent diluted metallic systems with much lower plasmon energies, which can be tuned by the concentration of the dopands in the range of

110

4 The Free Electron Gas 2 k FT q2

q<
1 ωp

ω

ε1 Fig. 4.16. The frequency dependent real and imaginary part of the dielectric function as obtained for the RPA spectrum

optical phonons at q = 0. The response of the system (measured by inelastic light scattering [118]) is then characterized by coupled plasmon–phonon modes with frequencies determined by the zeros of the dielectric function in the long-wavelength limit 2 ωp2 ωL2 − ωT ε(ω) = ε∞ 1 − 2 + 2 , (4.141) ω ωT − ω 2 which is the sum of the plasmon and phonon contributions (see Sect. 3.5). Note that the plasmon contribution is modified by the dielectric constant ε∞ of the semiconductor accomodating the diluted plasma, while the electron mass becomes the eﬀective mass m∗ . The biquadratic equation ε(ω) = 0 has two solutions, which in dependence on the carrier concentration show an anticrossing behavior typical for the coupling between the two modes, which is also experimentally verified.

4.7 Electronic Correlation The mean energy per electron in the ground state at T = 0 K ǫ0 =

1 E0 = ǫHF 0 + ǫc N

(4.142)

can be separated into the HF energy and the correlation energy. The former is determined by the filling of the Fermi sphere and considering the exchange interaction (as a consequence of the Pauli principle). The latter takes into account the electron–electron interaction beyond the HF approximation and will lead to a modified picture of the pair–distribution function of Fig. 4.12, in which the exchange hole is replaced by the correlation hole.

4.7 Electronic Correlation

111

It is helpful to start with the expectation value (per electron) of the electron–electron interaction in (4.107) for the exact ground state |Ψ0 ǫxc =

1 vq Ψ0 |(Nq† Nq − N )|Ψ0 , 2N

(4.143)

q =0

where the index xc refers to exchange and correlation. It can be reformulated by making use of the expression for the dynamic structure factor +∞ 1 ˆq (t)N ˆ−q (0)|Ψ0 dt S(q, ω) = e−iωt Ψ0 |(N (4.144) 2π −∞ which can be written also as S(q, ω) =

m

ˆq |Ψm |2 |Ψ0 |N

1 2π

+∞

e−i(ω−ωm0 )t dt .

The integral is the δ function and thus ∞ ˆq |Ψm |2 S(q, ω)dω = |Ψ0 |N 0

(4.145)

−∞

m

ˆq N ˆ−q |Ψ0 = N S(q) . = Ψ0 |N

(4.146)

Here S(q) is the static structure factor which now enters the (still exact) expression for ǫxc : 1 ǫxc = vq (S(q) − 1). (4.147) 2 q=0

Due to the isotropy of the free electron system we have S(q) = S(q) and we may write after integration over the angles ! ∞ 1 e2 kF ǫxc = − 2 − (S(q) − 1)dq . (4.148) 2π ε0 2kF 0 By introducing the dimensionless function ∞ 1 γ(kF ) = − (S(q) − 1)dq 2kF 0

(4.149)

and expressing kF by the density parameter rs we find (in Ry) 4 ǫxc (rs ) = − πrs

9π 4

!1/3

γ(rs ) .

(4.150)

The static structure factor S(q) is shown in Fig. 4.17 for diﬀerent values of rs . It should be kept in mind here that the expectation value of the kinetic energy (to be taken for the exact ground state) depends on the interaction.

112

4 The Free Electron Gas

S(q) 1.0 0.8 0.6 rs = 0.001

0.4

rs = 2 rs = 6 rs = 10

0.2 0

0

0.5

1.0

1.5

2.0

q/kF

2.5

Fig. 4.17. The static structure factor S(q) for diﬀerent values of the density parameter rs (after [119])

This is accounted for within the following consideration known as the ground state theorem: Let the system Hamiltonian be given by H(g) = Hkin + gHint

(4.151)

where g = e2 is the strength of the (electron–electron) interaction. The eigenstates of H(g), including the ground state |Ψ0 (g), depend on g and Eint (g) = Ψ0 (g)|gHint |Ψ0 (g)

(4.152)

is the exact contribution of the interaction to the ground state energy E0 (g). Then, according to the theorem g 1 E0 (g) = E0 (0) + E (g ′ )dg ′ . (4.153) ′ int 0 g To prove this theorem one has simply to write ! dE0 (g) dH(g) d = Ψ0 (g)| |Ψ0 (g) + Ψ0 (g)| H(g)|Ψ0 (g) dg dg dg ! d |Ψ0 (g) +Ψ0 (g)|H(g) dg 1 d = Eint (g) + E0 (g) Ψ0 (g)|Ψ0 (g). (4.154) g dg The last term vanishes, because |Ψ0 (g) is normalized, and the resulting differential equation can be integrated directly to yield the statement of the theorem.

4.7 Electronic Correlation

113

For the homogeneous electron gas we can immediately write ǫ0 =

kF 3 EF − 2 5 2π ε0

e2

γ(g ′ )dg ′ .

(4.155)

0

Thus, in terms of the density parameter rs , the exact ground state energy per electron reads (in Ry) !1/3 1 9π 4 2.2099 − γ(λrs )dλ. (4.156) ǫ0 (rs ) = rs2 πrs 4 0 If γ or S(q) or Im1/ε(q, ω) is given, the ground state energy per electron ǫ0 (rs ) is also known. The first term is the contribution of the kinetic energy in the Sommerfeld (or noninteracting electron) model, the second term describes the exchange and correlation energy (including also modifications of the kinetic energy due to the electron–electron interaction), and the exact correlation energy per electron can be written as !1/3 1 9π 0.916 4 ǫc (rs ) = − γ(λrs )dλ. (4.157) rs πrs 4 0 By definition ǫc (rs ) vanishes in the HF approximation. Calculations of the correlation energy have been carried out in diﬀerent approximations for ε(q, ω) (or S(q) or γ), whose accuracy depends on the density (or rs ) [64, 119]. Typically one uses ε(q, ω) = 1 −

vq π0 (q, ω) 1 + vq G(q)π0 (q, ω)

(4.158)

with diﬀerent forms of G(q). One particularly simple form is G(q) = 0, for which the dielectric function takes the RPA form of (4.125). But also in general, considering correlation means some kind of screening the Coulomb interaction. Results are visualized by plotting the pair–distribution function (see Problem 4.12) which for the homogeneous electron system reads ∞ 3 g(r) = 1 + q sin(qr) (S(q) − 1) dq . (4.159) 2rkF3 0 As can be seen in Fig. 4.18, the HF result, with g(r → 0) = 1/2, is recovered for dense electron systems (rs ≪ 1) while in the low-density case (large rs ) g(r → 0) → 0 and the exchange or Fermi hole evolves into a correlation hole because the electrons try to avoid each other (irrespective of their spin). We will see in the next chapter how these concepts of exchange and correlation can be considered in describing the electronic states of the inhomogeneous electron systems found in solids. Let us also briefly discuss the low density limit rs ≫ 1, for which the electron–electron interaction is expected to dominate the kinetic energy. The

114

4 The Free Electron Gas

g(r) 1.0 0.8 rs = 0.001

0.6

rs = 1 rs = 5

0.4

rs = 15 rs = 1

0.2 0

rs = 5

0

0.5

1.0

1.5

2.0 2.5 r/rsa0

Fig. 4.18. The pair–distribution function g(r) for diﬀerent values of the density parameter rs (after [119]). The symbols refer to Monte–Carlo simulations of [120]

ultimate limit, when the kinetic energy can be completely neglected, leads to a problem of classical electrostatics: the ground state is defined by the configuration of point charges with the lowest potential energy. This configuration is a crystalline one: the Wigner crystal. Its electrostatic energy shall be evaluated here in the jellium model, in which each electron is assigned the volume of the Seitz sphere with radius r0 = rs aB . The potential energy of an electron in the center and the sphere filled with the compensating positive charge of the jellium background is r0 e2 3 ǫWC nd3 r = − Ry. (4.160) = − − 4πε r r 0 s 0 The potential energy of the positive jellium background in the Seitz sphere, ǫ+ , is to be calculated from the electrostatic potential of the homogeneously charged sphere ! e2 r2 +C (4.161) V (r) = 2 3 − rs aB 2 where C = 3e2 /2rs aB has to be chosen to establish charge neutrality, thus r 2 2 ! 1 3− Ry (4.162) V (r) = rs rs aB and ǫ+ =

1 2

0

r0

V (r)nd3 r =

6 Ry. 5rs

(4.163)

4.7 Electronic Correlation

115

The electrostatic energy per electron in the Wigner crystal amounts to ǫWC = ǫWC − + ǫ+ = −

1.8 Ry. rs

(4.164)

This expression corresponds to the sum of exchange and correlation energy. With the latter being given by ǫx = −0.916/rs Ry, the correlation energy per electron in the Wigner crystal turns out to be ǫc (rs ≫ 1) = −

0.88 . rs

(4.165)

We may finally compare with the electrostatic energy of the smeared (instead of the point-like) electron in the Seitz sphere: it is given by ǫjell − = −2ǫ+ = −

2.4 Ry rs

(4.166)

which is larger than ǫWC − . Thus, the Wigner crystal is the predicted ground state of the strongly diluted homogeneous electron system. The experimental verification of the Wigner crystal represents a tremendous challenge because the realization of a diluted homogeneous electron system turns out to be very diﬃcult.

Problems 4.1 Calculate the interaction energy of a homogeneous electron gas (density n = N/V ) with a homogeneous jellium background of positive charges with the same density and the electrostatic potential energy of this background. Hint: Make use of the Fourier transform of the Coulomb interaction. 4.2 According to (4.25) the chemical potential μ(T ) is determined by the particle density n. For the three-dimensional electron gas, calculate the temperature dependence of μ for low temperatures (kB T ≪ EF ) up to the order (kB T /EF )2 ! 4.3 Calculate the density of states for a system of electrons, which can move freely in a plane (two-dimensional electron gas) or along a straight line (1-dimensional electron gas). Give an expression for the density of states of bound electrons (zero-dimensional electron gas). Plot and discuss these results. 4.4 Typical electron densities are n ≃ 1023 cm−3 for metals and n ≃ 1014 cm−3 for doped semiconductors (see Table 4.1) Consider the magnetic quantum limit, when all electrons are in the lowest Landau level, and derive for this case a relation between the magnetic field and n. For which of the two systems is it possible to reach the magnetic quantum limit in a laboratory?

116

4 The Free Electron Gas

4.5 The Pauli spin-paramagnetism is determined by the spin-splitting of the electronic energy spectrum due to the Zeeman term (see Fig. 4.7). Express this contribution to the magnetization by the number of noncompensated electron spins and compare the result with (4.53). 4.6 Discuss the quality of the HF approximation starting from ǫHF (4.93) 0 for metals and semiconductors (see parameters given in Table 4.1). In terms of the density parameter rs , which system is of higher density? 4.7 In order to go beyond the HF approximation (which is a first order perturbation calculation) one may consider the second order perturbation correction E2 due to the Coulomb interaction. (a) Make use of the Fermi sphere to find excited states which yield a contribution to E2 . (b) Distinguish between direct and exchange contributions to E2 and show that the direct term can be formulated as E2dir = −4m

kqp

vk2

np (1 − np+k )nq (1 − nq+k ) h2 k · (k + q + p) ¯

(4.167)

where np is the occupation number. (c) Show by expanding np+k for small k that p np (1 − np+k ) ∼ k and use this result to find that (for small k) E2dir is logarithmically divergent. Interpret this result! 4.8 Field operators are defined by Ψ(r) = cα Ψα (r), Ψ† (r) = c†α ψα∗ (r) (4.168) α

α

with fermion operators cα , c†α and a complete orthonormal set of single √ particle wave functions ψα (r) (for free electrons ψk (r) = exp(ik · r)/ V ). (a) Discuss the meaning of Ψ(r) and Ψ† (r). (b) Derive the commutation relations for Ψ(r) and Ψ† (r). (c) The density operator is given by n ˆ (r) = Ψ† (r)Ψ(r). Give for free electrons the Fourier components n ˆ defined by q n ˆ (r) = q eiq·r n ˆ q . What is the meaning of n ˆ q for q = 0 and q = 0? d) Express the Coulomb interaction in terms of the number fluctuation ˆq = V n operators N ˆq . 4.9 Calculate the thermal expectation value of the number fluctuation ˆq by starting from the exact expression operator N ∞ ˆq (τ ), N ˆ−q (0)]0 Next dτ ˆq = lim vq 1 eiωτ −Γ τ [N (4.169) N Γ →0 i¯ h 0 ˆq RPA ) with by replacing the bare external charge eNext by e(Next + N ˆ the induced number fluctuation Nq RPA in RPA and evaluating the expectation value of the commutator not in the exact ground state of the interacting system but in the ground state of the system without interaction. This is the self-consistent field approximation.

Problems

117

4.10 Prove the sum rule

0

∞

Im

1

2 π 1 ωdω = − ωp2 ǫ(q, ω) 2

(4.170)

where ωp is the plasma frequency. Hint: Calculate the ground state expectation value of [[Hjell , n ˆ q ], n ˆ −q ] by making use of the commutation relations of ck and c†k and by introducing a complete set of eigenstates of Hjell . 4.11 Evaluate the real part of the dielectric constant (4.132) for the static case (ω = 0) in the long-wavelength limit |q| → 0 (Thomas–Fermi approximation). It will depend on the parameter kFT defined by 2 = kFT

3N e2 . 2ε0 V EF

(4.171)

Consider the screened Coulomb potential vq /ε1 (q, 0): How does the Fourier transform of this potential look like and what is the obvious −1 meaning of kFT ? 4.12 The ground state density–density correlation function is defined as 1 Ψ0 |ˆ p(r) = n(r + r ′ )ˆ n(r ′ )|Ψ0 dr ′ . (4.172) N 3 The Fourier transform of p(r) is S(k) = eik·r p(r)d3 r. It can be obtained from the dynamic form factor S(k, ω) with 1 S(k) = S(k, ω)dω. (4.173) N Use the relation between S(k, ω) and Im (1/ǫ(q, ω)) to calculate p(r) in HF approximation.

5 Electrons in a Periodic Potential

In Chap. 4, the crystal structure of the solid has been suppressed by smearing out the periodic configuration of ions in the jellium model, because the many-body aspects of the electron–electron interaction stood in the focus of interest. Now, our attention will be, in addition, at the eﬀect of the periodic potential formed by the ions in the configuration {R0n,τ } of a crystal lattice. Therefore, we reverse the introduction of the jellium term in Sect. 4.3 with the replacement H+ ⇒

n,τ ,l

v(r l − R0n,τ ) +

1 V (R0nτ − R0n′ τ ′ ) = V (r l ). 2 n,n′

(5.1)

l

τ ,τ ′

Note the simplified notation of the single-particle potential V (r l ), which does not explicitly refer to the ion positions. It is invariant under lattice translations and is the same for all electrons. The Hamiltonian for a system of N electrons in the crystal volume HN =

N l=1

! N p2l 1 + V (r l ) + v(r l − r k ), 2m 2 k,l=1

(5.2)

k=l

with the Coulomb potential v(r l − r k ) =

e2 , κ|r l − rk |

κ = 4πε0

(5.3)

defines the starting point for this chapter. The complexity of the eigenvalue problem of HN comes from the simultaneous presence of the periodic lattice potential and the electron–electron interaction. As we have already seen in Chap. 4, the latter prevents a rigorous solution of the eigenvalue problem and some approximate treatment has to be applied. In some textbooks, e.g., [6, 7, 13, 121], the potential energy terms of HN , namely l V (rl ) and the electron–electron interaction, are replaced U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 5,

120

5 Electrons in a Periodic Potential

by an eﬀective single-particle potential: HN

⇒

l

! p2l + Veﬀ (r l ) 2m

(5.4)

which incorporates the many-body aspect in an approximate way. Other textbooks, e.g., [8, 9, 122], provide arguments, on how this replacement can be justified and we shall follow this line in Sect. 5.1. Given such a potential, electrons can be understood as independent particles and in a wave picture, are expected to undergo Bragg reflections due to the periodic potential (as in the case of elastic or electromagnetic waves). The consequence is that the energy spectrum becomes a band structure with energy intervals for which propagation of electrons is possible. They are separated by gaps, where this is not the case. In this approximation, the Hamiltonian is a sum of identical single-particle terms and can be separated with a product Ansatz into one and the same single-particle Schroedinger equation for all electrons: ! h2 ¯ ∆ + Veﬀ (r) ψα (r) = Eα ψα (r). − (5.5) 2m Here, α denotes a complete set of single-particle quantum numbers, which, as will be outlined in Sect. 5.2, are the band index n and the crystal momentum or wave vector k in the first Brillouin zone. Spin–orbit coupling and a spin index can be added if required. The energy eigenvalues Eα = En (k) form the energy band structure of electrons. The foremost task is the justification of this approach and the definition of the eﬀective single-particle potential Veﬀ (r).

5.1 Density Functional Theory The many-particle problem is defined by the Hamiltonian HN in (5.2). The eigenvalue equation HN ΨN = EΨN , is the same as (2.17) or (2.25) with the electron–ion interaction considered for the crystalline equilibrium configuration of the ions. But it corresponds also to that of the jellium model treated in Chap. 4 with the exception, now that the crystal structure of the ions is considered. Thus, we may use the concepts developed for the homogeneous electron gas. The simplest approach would be to apply the Hartree approximation with the product Ansatz N * ΨN (r 1 . . . rN ) = ψα (rα ). (5.6) α=1

The single-particle wave functions ψα have to be determined from the condition that the expectation value of HN takes a minimum under the constraint of their normalization. This leads to Ritz’ variational problem (see Sect. 4.1) & ' Eα ψα |ψα − 1 = 0 (5.7) δ ΨN |HN |ΨN − α

5.1 Density Functional Theory

121

where the normalization condition enters with the Lagrangian parameters Eα . The expression in the curly brackets is a functional of the single-particle wave functions ψα , thus the variation is to be performed as the functional derivative with respect to ψα or ψα∗ . If this is done with ! N p2 3 ∗ + V (r) ψα (r) d rψα (r) ΨN |HN |ΨN = 2m α=1 N 1 d3 rd3 r′ ψα∗ (r)ψβ∗ (r ′ )v(r − r′ )ψα (r)ψβ (r ′ ) + 2 α,β=1 α=β

(5.8) one obtains the Euler–Lagrange equations of the variational problem (note that the variational derivative gives two contributions from the double sum of the interaction term) e2 β =α |ψβ (r ′ )|2 h2 ¯ 3 ′ − ψα (r) = Eα ψα (r). ∆ + V (r) + d r (5.9) 2m κ|r − r ′ | They have the form of (5.5) with an eﬀective potential composed of the potential V (r) due to the periodic ion configuration and a contribution from the electron–electron interaction, which can be identified as the Hartree potential: it describes the direct Coulomb interaction of the electron in the state α with the charge distribution of all the other electrons. Remember, that in the jellium model the Hartree contribution exactly compensates the jellium term, to which V (r) simplifies in this case. Three aspects are to be mentioned here: the Hartree potential for ψα e2 VH,α (r) = d3 r ′ |ψβ (r ′ )|2 (5.10) ′ κ|r − r | β=α occ

depends on all the other occupied single-particle states with β = α with the consequence that (5.9) can be solved only if the solutions (the eigenfunctions) are known. How such a self-consistent solution can be achieved, will be explained at the end of this section. Moreover, the Hartree potential depends also on the state α, and one has to deal with equations containing diﬀerent potentials. Finally, as we know from Sect. 4.3, the product Ansatz does not prevent double occupation of single-particle states, thus being in contrast with the Pauli principle. Nevertheless, the Hartree approximation indicates in some way, how to arrive at the eﬀective single-particle potential we are looking for, without reaching it. In the next step, we assume ΨN in the form of a Slater determinant as in (4.61) constructed from single-particle wave functions ψα (x), where x comprises space and spin variables and α is a complete set of quantum numbers including spin, but proceed as before. By carrying out the variation

122

5 Electrons in a Periodic Potential

(Problem 5.1) we arrive at the Hartree equations (5.9) augmented by an additional term Vx,α (r)ψα (r) = − d3 r ′ v(r − r′ ) (5.11) ψβ∗ (r ′ )ψα (r ′ )ψβ (r). β=α, occ

This exchange term results from the fact that the expansion of the Slater determinant contains products of single-particle wave functions with interchanged particle coordinates, which do not exist in the simple product Ansatz of the Hartree approximation. Here, Vx,α (r) is not a simple potential acting as a factor on ψα (r) but a nonlocal integral operator, because ψα (r ′ ) appears under the integral. The -sign in the sum under the integral indicates the restriction to contributions from states β = α with parallel spins and results from the summation over the spin variables. We may include the contributions for α = β in the Hartree and the exchange terms, which cancel each other, to obtain the Hartree–Fock equations

¯2 h − ∆ + V (r) + 2m −

d3 r ′ n(r ′ )v(r − r′ )

d3 r ′ nxα (r, r′ )v(r − r ′ ) ψα (r) = Eα ψα (r).

Here we have introduced the density |ψβ (r)|2 n(r) =

(5.12)

(5.13)

β,occ

and the Hartree–Fock (HF) or exchange density nxα (r, r′ ) =

ψβ∗ (r ′ )ψα (r ′ )ψα∗ (r)ψβ (r) β, occ

ψα∗ (r)ψα (r)

.

(5.14)

√ For free electrons with ψα (r) = eik·r / Vc (here we denote the crystal volume by Vc ) (5.14) leads back to the results in Sect. 4.4 (Problem 5.2). As the Hartree potential (5.10), also the exchange potential is statedependent and the HF equations have to be solved self-consistently. However, by replacing the state-dependent HF density by its average over all occupied states N 1 x nxα (r, r ′ ) ⇒ n ¯ HF (r, r′ ) = n (r, r ′ ), (5.15) N α=1 α we obtain an eﬀective single-particle potential 2 ′ ¯ HF (r, r ′ ) 3 ′ e n(r ) − n HF . Veﬀ (r) = V (r) + d r κ|r − r′ |

(5.16)

5.1 Density Functional Theory

123

which is the same for all electrons. Besides the electrostatic interaction with the periodic ion configuration, it contains the electron–electron interaction in approximate form. It should be noted, however, that although deriving from the Ansatz with the Slater determinant, it is not exactly the HF approximation because the averaging of the exchange density in (5.15) represents a nonsystematic step. The question arises if one could go systematically beyond HF and consider in Veﬀ (r) correlation contributions besides exchange also as discussed in Chap. 4. The answer to this question comes from the density functional theory (DFT) [9, 122–126], which since its formulation in the sixties has developed into the most frequently used concept in calculating the electronic structure of atoms, molecules, and solids and earned in 1998 the Nobel prize in Chemistry shared by Kohn and Pople. Here we outline the three essential steps of DFT: 1. The Hohenberg–Kohn1 theorem [127] 2. The Kohn–Sham equations [128] 3. The local density approximation (LDA) The basic idea of the Hohenberg–Kohn theorem is that the ground state energy of the N electron system for a given external potential (here that of the periodic configuration of the ions V (r)) is a unique functional of the singleparticle density n(r). This idea is conceivable: when adding the electrons to the external potential, they will arrange in a unique way to establish the state with lowest energy. This state will be characterized by a many-particle wave function Ψ0 ({r l }) and a single-particle density (the spin degree of freedom is suppressed here) n(r) = . . . d3 r2 . . . d3 rN |Ψ0 (r, r2 . . . r N )|2 , N = d3 r n(r). (5.17) The statement of the theorem is: Let n(r) be the (inhomogeneous) single particle density for the ground state of a system of interacting electrons in an external potential V (r) and let the density n′ (r) have the same relation to the external potential V ′ (r). Then it follows from n(r) = n′ (r) that V (r) = V ′ (r) up to a constant. The proof of the theorem is indirect: Assume two systems with external potentials V (r) = V ′ (r), which diﬀer by more than just a constant, but have identical densities n(r) = n′ (r) in the ground state. Then one has the ground state energies E0′ = Ψ′0 |T + V ′ + U |Ψ′0 ,

E0 = Ψ0 |T + V + U |Ψ0

(5.18)

where T and U denote the kinetic energy and the electron–electron interaction, respectively, and |Ψ0 and |Ψ′0 the ground states of the system with external potential V (r) and V ′ (r), respectively. As |Ψ0 is not the ground state of the system with the external potential V ′ (and |Ψ′0 not that of the system 1

Walther Kohn *1923, shared the Nobel prize in chemistry 1998 with J. Pople.

124

5 Electrons in a Periodic Potential

with V ), we may formulate the following relations E0′ < Ψ0 |T + V ′ + U |Ψ0 = E0 + Ψ0 |V ′ − V |Ψ0

(5.19)

E0 < Ψ′0 |T + V + U |Ψ′0 = E0′ + Ψ′0 |V − V ′ |Ψ′0 .

(5.20)

and Due to the assumption of identical densities, the last terms in both relations are identical (except for the sign) and by taking the sum of these expressions we find the contradictory relation E0 + E0′ < E0 + E0′ . Thus, the assumption must be incorrect, while the statement of the theorem is correct. (We note that this proof applies, if the ground state is not degenerate. A more general proof was given by Levy [124].) The Hohenberg–Kohn theorem can be formulated also by saying that for the given external potential V (r) the exact ground state energy E0 is a unique functional of the exact ground state density n(r): E0 = Ψ0 |T + V + U |Ψ0 = EV [n(r)] .

(5.21)

It tells us, that in order to find the ground state energy of the N -particle problem with HN , it is not required to find the exact many-particle wave function Ψ0 (which is a function of the coordinates of N electrons), it suﬃces to find the exact single-particle density n(r) (which depends only on the coordinates of one particle). Note, that the same statement as for the ground state energy E0 can be made for the expectation value of any observable in this state (and this includes also the response functions). One has to keep in mind, however, that the theorem is restricted to the system ground state. Now, the problem to be solved is to find E0 as minimum of the energy functional EV [n(r)] = T [n(r)] + V (r)n(r)d3 r n(r)n(r ′ ) 3 3 ′ e2 d rd r + Exc [n(r)], + 2 κ|r − r′ |

(5.22)

where the first and last terms describe the exact functionals of the kinetic and exchange–correlation energy, respectively, of the interacting electron system, while the second and third term are the electrostatic energy of the electron density en(r) in the external potential V (r) and the Hartree energy for this charge density. In mathematical language (5.23) represents, for a given external potential, a mapping of the densities onto the energies of which we have to find the minimum. T [n] and Exc [n] are not exactly known and will be considered in an approximate way, as outlined in the following. The variational problem can be treated by assuming the representation of the density by a complete set of single-particle wave functions

5.1 Density Functional Theory

n(r) =

N

α=1

2

|ψα (r)| ,

ψα |ψα = 1,

N=

n(r)d3 r.

125

(5.23)

Taking the functional derivatives with respect to the ψα∗ (which corresponds to the variation with respect to the density n(r)) and assuming T [n] as for non-interacting electrons ! ¯h2 ∗ T [n(r)] ≃ T0 [n(r)] = ∆ ψα (r)d3 r ψα (r) − (5.24) 2m α one arrives at a set of Schroedinger equations for the single-particle functions ψα

−

¯2 h ∆ + V (r) + 2m

e2 n(r ′ ) 3 ′ d r + V (r) ψα (r) = Eα ψα (r), (5.25) xc κ|r − r ′ |

the Kohn–Sham equations. They have the form of (5.5), but are integro– diﬀerential equations due to the fact that, because of Vxc (r)ψα (r) =

δ δn Exc [n] ∗ , δn δψα

(5.26)

Vxc (r) is a nonlocal integral operator and we may write similar to (5.11) Vxc (r)ψα (r) = d3 r′ Σxc (r, r′ ; Eα )ψα (r ′ ). (5.27) Here, Σxc (r, r′ ; Eα ) is the exchange–correlation self-energy or mass operator. In the last step, the Kohn–Sham equations can be reduced to the form of (5.5) by applying the local density approximation (LDA) with the replacement LDA Exc [n(r)] ⇒ Exc [n(r)] = n(r)ǫxc n(r) d3 r, (5.28)

where ǫxc (n) is the xc energy per electron of a homogeneous system with density n. For each r of the inhomogeneous system (with external potential V (r) = const) the xc energy of the jellium model (Chap. 4) is taken with the local density n = n(r). This allows one to write the xc term as a local single-particle potential LDA Vxc (r) =

d nǫxc (n) n=n(r) dn

(5.29)

with ǫxc (n) = ǫxc (rs ) from (4.150). Thus, we have found the eﬀective singleparticle potential to be used in (5.5) LDA LDA Veﬀ (r) = V (r) + n(r ′ )v(r − r′ )d3 r ′ + Vxc (r). (5.30)

126

5 Electrons in a Periodic Potential

The considerations of this section apply to any interacting fermion system and external potential. The DFT-LDA has been and is currently applied to atoms, molecules, and condensed matter. With the available computer power it has become the dominant tool for solving quantum many-body problems. The input with respect to the exchange–correlation energy comes from the homogeneous electron gas (which accounts also for the replacement made in (5.24)), which we have studied in detail in Chap. 4. The DFT concepts have been refined to include spatial inhomogeneity of the density in the xc energy (in the generalized gradient approximation), spin polarization (in the spin or SDFT), and time dependent perturbations (in the DFT perturbation theory, mentioned in Chap. 3) [100, 123]. Also, the restriction to the system ground state has been overcome by taking care of the discontinuity of the correlation energy in its dependence on the quasiparticle energy across the Fermi energy (quasi-particle corrections in the GW approximation) for which we refer also to Chap. 7 [123, 129, 130]. But even without the latter, the eigenvalues Eα of the Kohn–Sham equations with the single particle potential (5.30) are usually taken as single particle energies not only for states occupied in the ground state, but also for unoccupied states. This identification is widely supported by the agreement between experimental mapping of the energy bands from photoelectron spectroscopy (PES) and results from DFT calculations [131]. It should be noted, however, that this agreement is not found for optical properties in semiconductors and insulators due to the band gap problem: as it turns out, the separation between conduction and valence band states, as obtained from DFT-LDA calculations is too small [130]. This discrepancy can be resolved systematically by considering diﬀerent xc energies for conduction and valence band states, as is done in the already mentioned GW approximation. As we have seen in this section, the Hartree, the Hartree–Fock, and the eﬀective LDA potential (5.30) depend on the solutions of the corresponding Schroedinger equations. For this type of problem an iterative procedure applies: for the given external potential one has to choose a single-particle density n0 (r) to start with, calculate the Hartree and LDA contributions and solve the Schroedinger equation. The occupied states of this solution define a density n1 (r), which in general will be diﬀerent from n0 (r) and is taken to create a modified Hartree and LDA potential. With these modified potentials (or a mixture with the potential of the previous step) the Schroedinger equation is solved again and the procedure is repeated until the calculated densities (or energy eigenvalues) for two successive iteration steps are reproduced within desired limits of accuracy. At this level, the obtained single-particle energies, wave functions and the potential to which they contribute via the occupied states are self-consistent.

5.2 Bloch Electrons and Band Structure

127

5.2 Bloch Electrons and Band Structure In the previous section we have reduced the many-body problem of the electrons in a solid to the single-particle problem Hψα = Eα ψα ,

H =−

¯2 h ∆ + Veﬀ (r) 2m

(5.31)

with the periodic eﬀective potential Veﬀ (r + R0n ) = Veﬀ (r).

(5.32)

The periodicity of the potential means invariance of the Hamiltonian under lattice translations [H, TR0n ] = 0 ,

with

i

0

TR0n = e− h¯ p·Rn

(5.33)

and the momentum operator p. The properties TR0n ψα (r) = ψα (r − R0n )

and |TR0n ψα (r)|2 = |ψα (r)|2

(5.34)

of the translation operator allow one to write 0

TR0n ψk (r) = e−ik·Rn ψk (r)

(5.35)

where the phase factor, the eigenvalue of the translation operator, is characterized by the wave vector k (note that h ¯ k is the eigenvalue of the momentum operator p). Due to the property (5.35) the eigenfunctions of (5.31) can be split into an exponential and a lattice periodic function ψk (r) = eik·r uk (r) ,

uk (r + R0n ) = uk (r).

(5.36)

This form, a modulated plane wave, is characteristic of electrons in a periodic lattice: the Bloch2 function. It is schematically depicted in Fig. 5.1. The

ψk(r) r Fig. 5.1. Schematic view of a Bloch function (solid line) and its plane wave part (dashed line), the dots mark the lattice points 2

Felix Bloch 1905–1983, Nobel prize in Physics 1952.

128

5 Electrons in a Periodic Potential

characterization of the Bloch function by the wave vector is unique within the 1st Brillouin zone, because the phase factor in (5.35) is the same for all k diﬀering by a reciprocal lattice vector. For a given k, the eigenvalue equation (5.31) has in general an infinite set of independent solutions, which are distinguished by an energy quantum number, the band index n. Thus the solutions of (5.31) are classified by the complete set of quantum numbers α = (n, k) (a spin quantum number can be added where required) ! h2 ¯ ∆ + Veﬀ (r) ψnk (r) = En (k)ψnk (r) − (5.37) 2m and the energy eigenvalues En (k) for all n and k from the 1st Brillouin zone define the energy band structure of electrons in a periodic potential. By applying periodic boundary conditions, one can easily verify, that there are as many diﬀerent k in the Brillouin zone as there are unit cells in the crystal volume. Thus, each energy band can accommodate one electron of either spin per Wigner–Seitz cell (Problem 5.3). Supplement: Symmetry of En (k): Besides the translation symmetry En (k) = En (k+G) considered already by restricting k to the first Brillouin zone, there are other symmetries of the problem according to which the band structure repeats within the Brillouin zone. 1. Due to time reversal invariance the Hamiltonian is hermitian, H = H † , i.e., it is real and we may write the complex conjugate equation to (5.37) ∗ ∗ = En (k)ψnk . Hψnk

(5.38)

∗ The complex conjugate Bloch function ψnk , written in the form of (5.36), is characterized by a phase factor with −k and belongs to the solutions of (5.37) in the form (5.39) Hψn−k = En (−k)ψn−k .

Both these equations are eigenvalue equations for the same operator and yield the same spectrum. If we consider the electron spin (whose direction changes under time inversion) as additional quantum number, we may write {En↑ (k) , all n} = {En↓ (−k) , all n} ,

(5.40)

where the curly brackets denote the whole set of eigenvalues. This degeneracy, following from time reversal symmetry, is known as Kramers degeneracy. 2. We mentioned already in Chap. 1 the symmetry of the periodic lattice under the operations S of the point group, which now means invariance of the Hamiltonian of (5.37) [H, S] = 0 ,

or

SHS −1 = H.

(5.41)

Application of the point group operation S to the wave function ψ(r) changes r into S −1 r and we may write

5.3 Almost Free Electrons and Pseudo-Potentials Sψnk (r) = ψnk (S −1 r) = eik·(S

−1

r)

unk (S −1 r)

129 (5.42)

−1

which because of k · (S r) = Sk · r means that the Bloch function (5.42) is one with wave vector Sk and we conclude, as for the time reversal symmetry, that {En (Sk) , all n} = {En (k) , all n} .

(5.43)

Thus, the energy band structure reflects completely the point group symmetry of the crystal structure. This property of the band structure can be exploited in performing sums over k, which can be restricted to the so-called irreducible wedge of the Brillouin zone. 3. Consider a combination of 1 and 2: if time reversal invariance combines with the point group symmetry, we have for operations S with Sk = −k that {En↑ (k) , all n} = {En↓ (k) , all n} .

(5.44)

In contrast to (5.40), this situation with the up and down spin states of a given k having the same energy, is called spin degeneracy.

Let us have a look back to the beginning of this chapter with the N electron Hamiltonian (5.2) as starting point. Within the DFT concepts we have reduced this many-particle problem to a single-particle one (of course with the restrictions already mentioned). Nevertheless, we can formulate HN in an approximate way (compare with (4.94)) En (k)c†nk cnk (5.45) HN ≃ nk

by making use of the band structure and by introducing fermion creation and annihilation operators c†nk and cnk with {c†nk , cn′ k′ } = c†nk cn′ k′ + cn′ k′ c†nk = δnn′ δkk′ .

(5.46)

In this approximate formulation, the electron–electron interaction is incorporated in the single-particle properties and we may address these Bloch electrons as the quasi-particles of the density functional theory.

5.3 Almost Free Electrons and Pseudo-Potentials Starting with this section we present concepts of band structure calculations which are outlined in many textbooks of Solid State Theory [11, 121, 132–134]. The task is to solve the Schroedinger equation (5.37) ! h2 ¯ ∆ + Veﬀ (r) ψnk (r) = En (k)ψnk (r) − (5.47) 2m for Bloch electrons with ψnk (r) = eik·r unk (r).

(5.48)

The simplest possible concept is to make use of the periodicity of unk (r) and expand it in the complete set of normalized plane waves with reciprocal

130

5 Electrons in a Periodic Potential

lattice vectors G ψnk (r) =

G

1 Cnk (G) √ ei(k+G)·r . Vc

(5.49)

The problem of solving (5.47) consists in finding the expansion coeﬃcients Cnk (G). For this purpose we use the expansion (5.49) in (5.47), multiply from left with a normalized plane wave with k + G′ , and integrate over the crystal volume Vc . This leads to a set of coupled homogeneous linear equations ¯ h2 (k + G)2 − E δGG′ + V (G − G′ ) Cnk (G) = 0. (5.50) 2m G

The first term results from the operator of kinetic energy and the orthogonality of the plane waves, the second is the Fourier component of the eﬀective potential ′ 1 ′ V (G − G ) = Veﬀ (r)ei(G−G )·r d3 r. (5.51) Vc The energy eigenvalues are obtained from the secular problem $ 2 $ $ ¯ $ $ h (k + G)2 − E δGG′ + V (G − G′ )$ = 0. (5.52) $ 2m $ The eﬃciency of this concept depends essentially on the convergence of the plane wave expansion (5.49) or on the strength of the periodic potential in terms of its Fourier coeﬃcients. When looking at Fig. 5.1 it is conceivable that a strong modulation of the Bloch function (caused by a strong potential) requires more terms in the plane wave expansion than a Bloch function with a weak modulation. In general, one can say that the plane wave expansion is expected to work well if in the secular problem (5.52) the kinetic energy terms (the diagonal terms of the matrix) dominate over the (non-diagonal) potential terms. This is the case for almost free electrons, for which the periodic potential acts as a weak perturbation. In fact (5.52) represents an expression of first order perturbation theory applied to the free electron states and leads to a Brillouin–Wigner perturbation expansion. The anticipation of a weak periodic potential and the occupation of the lowest bands deriving from a free electron parabola with the valence electrons, is consistent with the idea put forward at the beginning of this book that the constituents of the solid are ions and valence electrons. But we have not yet shown how this concept is considered in the eﬀective potential. Intuitively, the potential of an ion seen by a valence electron is that of a screened uncompensated charge outside of the closed shell radius but approximately zero (due to charge compensation) inside of this shell. This leads to the model of an empty core potential 0 r < Rc Vec (r) = (5.53) Ze2 −kFT r − 4πε0 r e r > Rc

5.3 Almost Free Electrons and Pseudo-Potentials

Vec (r)

131

vec(q)

Rc r

q

Fig. 5.2. Empty core pseudo-potential and its Fourier transform (the arrows indicate the length of reciprocal lattice vectors)

where Z is the ion charge, Rc the core radius, and kFT the inverse Thomas–Fermi screening length (see Problem 4.11). This potential is depicted in Fig. 5.2 together with its Fourier transform vec (q) = −

Ze2 vion (q) 2 ) cos qRc = ε(q) , ε0 (q 2 + kFT

(5.54)

which can be expressed as the ion potential vion (q) = −Ze2 /4πε0 q 2 divided by the dielectric constant ε(q) in the Thomas–Fermi approximation. Characteristic values of the parameters Z, Rc , and kFT , yield values for the Fourier coeﬃcients of the periodic potential of a few tenths of the Rydberg energy, which can in fact be considered as a perturbation on the scale of the free electron energies (which are of the order of 1 Ry). The empty core potential is a prototype pseudo-potential adapted to the anticipated construction of the solid out of ions and valence electrons [134, 139]. If we had started from nuclei and all electrons, then the eﬀective singleparticle potential would have the characteristic form depicted in Fig. 2.1 with strong attractive parts close to the nuclei. Such a potential would have strong Fourier coeﬃcients up to large reciprocal lattice vectors, which are required to obtain a converging representation, especially of the most strongly bound electrons in the inner shell. But as already discussed, these electrons are not relevant for solid state properties and we are interested here only in the valence electrons. In view of the all-electron problem, the valence electrons are in states which are orthogonal to those in closed shells and we can understand the problem of calculating the band structure as that of looking for eigensolutions of the all-electron Hamiltonian in that part of the Hilbert space, which is orthogonal to the core states. This view opens the principle access to pseudopotentials. Let us expand for this purpose the Bloch function ψnk (r) = r|nk in terms of plane waves r|k + G which by construction are made orthogonal to the core states |nk = (1 − P)|nkPW , |nkPW = Cnk (G)|k + G, (5.55) G

132

5 Electrons in a Periodic Potential

where P=

ν∈core

|νkνk|

(5.56)

is the projection operated onto the core Bloch states (see (5.67)). Making use of this expansion in the eigenvalue problem with the single-particle Hamiltonian but with all electrons considered in the periodic potential Vall , one finds p2 Hall |nk = + Vall − Eν |νkνk| |nkPW 2m ν∈core = E 1− |νkνk| |nkPW . (5.57) ν∈core

An eigenvalue equation for this expansion is obtained by rearranging (5.57) in the form p2 + Vall + (E − Eν )|νkνk| |nkPW = E|nkPW (5.58) 2m ν∈core with the nonlocal pseudo-potential operator Vpsp = Vall + (E − Eν )|νkνk|.

(5.59)

ν∈core

By construction, the solutions of (5.58) are orthogonal to the core electron states and yield the electron states of valence and conduction bands. When taking the expectation value of the nonlocal operator with |nkPW one finds (note, that we calculate valence electron states with E > Eν ) (E − Eν )|νk|nkPW |2 > 0, (5.60) ν∈core

i.e., the additional potential cancels (partially) the attractive potential of the nuclei and converts the all electron potential into a weak pseudo-potential. In this sense, the eﬀective potential in (5.47) is to be understood as a pseudopotential and we can safely assume, that the plane wave expansion converges with reasonable eﬀort. Initially, the Fourier coeﬃcients of the pseudo-potentials have been used as empirical parameters [134, 139] which were determined by fitting a calculated band structure to experimental data, such as Fermi surface parameters or energy gaps. In the course of time, the pseudo-potential concepts have been developed and it is now possible to perform ab initio band calculations which make use of so-called norm conserving and soft pseudo-potentials, which are free of adjustable parameters [140, 141]. Instead of treating the full problem of solving (5.52), we may look for situations where the free electron picture, with the parabolic dispersion of the

5.3 Almost Free Electrons and Pseudo-Potentials

133

Ist BZ E

- 4π a

- 2π a

0

2π a

4π a

k

Fig. 5.3. Free electron energy bands in one dimension: extended and repeated zone scheme

energy, is only slightly perturbed by the periodic potential. Figure 5.3 shows the free-electron band structure for a one-dimensional system without periodic potential but with the periodicity taken into account by introducing Brillouin zone boundaries. Due to the translation symmetry, the energy dispersion is periodic in k with periods 2π/a, i.e., the free-electron parabola may start at each value 2nπ/a with integer n. In order to avoid the redundancy in this repeated zone scheme, it is suﬃcient to consider the dispersion only in the first Brillouin zone (reduced zone scheme). This picture of the reduced zone scheme can be obtained also from the parabola starting at 0, by shifting those parts, which are outside the first Brillouin zone, by multiples of 2π/a to bring them back to this zone. This concept is not restricted to the free electron dispersion in 1D but applies as well to any realistic band structure not only of electrons (see e.g., the situation in Problem 3.1 for the linear chain, when the two masses become equal). Coming back to the secular problem (5.52) we recognize in Fig. 5.3 degeneracies of the free-electron energy dispersion at the Brillouin zone boundaries and in its center, e.g., the parabolas starting at 0 and at 2π/a cross at k = π/a. Writing the secular problem for the corresponding states one obtains $ $ $ $ E(k) − E V (G) $ $ $ = 0 , G = 2π , $ (5.61) $ $ a $ V (G) E(k − G) − E $

134

5 Electrons in a Periodic Potential

where E(k) is the kinetic energy of free electrons, with the zeros E± (k) =

1 E(k) + E (k − G) 2 1/2 1 2 2 (E(k) − E (k − G)) + |V (G)| . ± 4

(5.62)

2 For k = π/a one has E(k) = E(k − G) = h ¯ 2 /2m π/a and E±

π a

" ¯ 2 π 2 "" h = ± "V 2m a

!" 2π "" . a "

(5.63)

The eﬀect of the periodic potential is to remove the degeneracy of the free electron states (here at the boundary of the Brillouin zone) and to create an energy gap or band gap in the otherwise continuous spectrum. The gap is determined by a Fourier coeﬃcient of the potential for the reciprocal lattice vector connecting the degenerate plane waves. The dispersion of the lowest energy bands E± (k) around k = π/a is shown in Fig. 5.4 together with the corresponding free electron dispersion (thin dashed lines). In general, the reciprocal lattice vectors of degenerate plane waves fulfill the Bragg condition (Problem 5.4), thus the plane waves being reflected at the crystal planes is characterized by the reciprocal lattice vector form standing waves. This becomes evident by looking at the eigenfunctions. In our simple example with the energies of (5.63), they take the form ψ+ πa (x) ∼ i sin

πx a

and ψ− πa (x) ∼ cos

πx a

(5.64)

and their modulus shows minima (maxima) at the lattice points, which leads to the raising (lowering) of the energy due to the periodic potential in comparison with the free electron case. (Note that the potential is attractive for electrons in the vicinity of the ion positions.) The energy gaps and the horizontal slope of energy bands (preferentially) at the boundary of the Brillouin

E

π/2a

k

π/a

Fig. 5.4. Opening of a gap at the Brillouin zone boundary due to a periodic potential

5.3 Almost Free Electrons and Pseudo-Potentials

X M Γ

X

135

Γ

M

X Γ Fig. 5.5. Brillouin zones of the two-dimensional square lattice and Fermi contours

zone, which corresponds to a vanishing group velocity, are the most prominent features of the energy band structure. Let us now proceed to the situation of a 2D square lattice and look at the Fermi contours as depicted in Fig. 5.5. In the left part, we recognize the reciprocal lattice points and the construction lines for the Brillouin zones: the solid lines mark the central square of the first Brillouin zone (first BZ); the dashed lines, augmenting the first BZ to a square of double size, mark the second Brillouin zone. The four triangles outside of the first BZ can be rearranged by translations with reciprocal lattice vectors to form a square reproducing the 1st BZ as indicated in the right upper part of Fig. 5.5. The third BZ is obtained from the dash-dotted smaller triangles which again can be rearranged to a square of the size of the first BZ around the M point (see lower right part of Fig. 5.5). The free-electron dispersion in the extended zone scheme (Fig. 5.5, left) is a paraboloid with its minimum at the Γ point or k = (0, 0). In the presence of a 2D periodic potential, (causing the square lattice with lattice constant a) this continuous energy dispersion will be deformed in the vicinity of the Brillouin zone boundaries in connection with the opening of gaps (Problem 5.5). Instead of looking at the band dispersion, we want to discuss here the consequences of the Fermi circle shown in the left part of Fig. 5.5 for an assumed electron density, i.e., the area inside this circle defines the occupied states of the free-electron dispersion at T = 0 K. When constructing the Brillouin zones the Fermi circle, breaks into pieces which by rearrangement of the extended to the reduced zone scheme, give the grey areas indicating occupied states in diﬀerent bands. The first BZ or the lowest energy band is completely occupied, the second and third band (plotted for a BZ around the point M ) are only partially filled and the grey areas are only faintly reminiscent of deriving from a circle. In the presence of a weak periodic potential, these contours, which separate the empty from the occupied states, are distorted when the gaps open at the BZ boundaries. These Fermi contours (which in 3D become Fermi

136

5 Electrons in a Periodic Potential

energy in 2π

2

( a ( Ry

3

2

1

L

Γ

X

K

Γ

Fig. 5.6. Free electron energy bands in an fcc lattice

surfaces) can experimentally be detected by analyzing the quantum oscillations of the magnetic susceptibility (deHaas–vanAlphen eﬀect) or of the magneto-resistivity (Shubnikov–deHaas eﬀect). These methods are of central importance in band structure investigations of metals. Modern semiconductor technology has led to man-made two-dimensional electron systems (see Sect. 5.6) with a periodic potential (lateral surface superlattices). The band structure of these systems and their Fermi contours have been analyzed with the band theory of almost free electrons in two dimensions [135]. Let us look at the 3D case. The free electron dispersion in a fcc lattice shown in Fig. 8.6 (Problem 5.6) will serve as the basis to understand in the following some of the electronic properties of a noble metal (Ag), a normal (trivalent) metal (Al), and a semiconductor (Si). The free electron parabola, starting at the Γ point is backfolded to account for Bragg reflection at the BZ boundaries. These backfolded branches can be understood also in the repeated zone scheme as deriving from parabolas starting at reciprocal lattice vectors G. Some of these branches are degenerate. The characteristic energy scale (and the length of the axes) changes with the lattice constant, but is typically of the order of the Rydberg energy. A weak potential with Fourier coeﬃcients of a few tenths of the Rydberg energy slightly changes this picture by lifting degeneracies (seen here e.g., along the line Γ − K − X) and opening gaps as depicted in Fig. 5.7 for Al. The density of states for these energy bands, shown in Fig. 5.8, exhibits the similarity with the square root dependence (dashed line) found for free electrons, while the energy gaps are responsible for the deviations seen in the solid curve. Imagine the filling of these states with electrons: in the free electron picture we have the Fermi sphere in k space, which is mapped onto the band structure in the reduced zone scheme, where each band can accommodate two electrons per atom in the unit cell (see Problem 5.3). For a monovalent metal

5.3 Almost Free Electrons and Pseudo-Potentials

137

Al

energy [Ry]

1.0

EF

0.8 0.6 0.4 0.2

L

Γ

X

K

Γ

Fig. 5.7. Energy band structure of Al after [134] 0.3 states atom eV

0.2

D(E) in

Al

0.1

0

EF

0

0.2

0.4 0.6 0.8 energy in Ry

1.0

1.2

Fig. 5.8. Density of states of electrons in Al

(like Ag), this leads to a half filled lowest band and the Fermi energy cuts the band structure in a region away from zone boundaries where the free electron dispersion is almost unchanged by the periodic potential. In this case, the Fermi surface more or less maintains its spherical shape as for free electrons. Nevertheless, depending on the lattice constant, the Fermi contour can come suﬃciently close to the BZ boundaries which (for an fcc lattice around the L points) are closest to the Γ point and one finds the situation shown in Fig. 5.9. It exhibits the Fermi surface of Ag (representative of the noble metals) in the repeated zone scheme. The Fermi sphere is distorted here by the formation of necks close to the L points as a consequence of gap formation. In deHaas–vanAlphen measurements for diﬀerent orientations of the magnetic field, the extremal cross sections of the Fermi surface perpendicular to the magnetic field (indicated as N for neck, B for belly, and DB for dumb–bell in Fig. 5.9) are detected by the periods of their corresponding oscillations (see discussion at the end of Sect. 4.2).

138

5 Electrons in a Periodic Potential

Fig. 5.9. Extended zone scheme with Fermi surfaces and extremal cross sections (see text) of Ag after [136] W

Γ W X

Γ

X

L

(a)

(b)

Fig. 5.10. Fermi surfaces of Al in the second and third Brillouin zone after [137]

For the trivalent metal Al the Fermi energy is indicated in Fig. 5.7. All states of the lowest band (or the first BZ) are filled, while the states in the second and third band are only partially filled. The Fermi contours, derived from the Fermi sphere of the free electrons for these bands, are depicted in Fig. 5.10. They will be slightly changed (essentially by rounding oﬀ the sharp edges) by gap formation due to the periodic potential. Looking at these strange surfaces, one can imagine that their analysis from deHaas–vanAlphen oscillations, which correspond to extremal cross sections of the Fermi surface perpendicular to the applied magnetic field (see Sect. 4.2), can be quite an involved task. The last example to be presented in the context of Fig. 8.6 is the band structure of Si. It crystallizes in the diamond structure with two atoms with each four electrons per unit cell. The structure factor of diamond diﬀers from that of the simple fcc lattice and as a consequence, the non-vanishing Fourier components of the periodic potential and the gap structure are diﬀerent

5.4 LCAO and Tight-Binding Approximation

6

L3

3 energy [eV]

L1

0 –3

Γ2’

Si

K1

Γ15 Γ25’

X1 X4

L 3’

K3 K2 K1

–6 –9 –12

K3

L1 L 2’

L

139

X1

K1

Γ1

Γ

k

X

U,K

Γ

Fig. 5.11. Energy band structure of Si after [138]

(Problem 5.7). This is clearly seen by comparing the band structure of Si (Fig. 5.11) with that of Al (Fig. 5.7) which both derive from the free electron dispersion of the fcc lattice (Fig. 8.6). The eight valence electrons per unit cell in Si, fill the four lowest energy bands in Fig. 5.11 (the valence bands), which are separated from the empty conduction bands by an energy gap, which extends throughout the BZ. This particular situation of the energy band structure with a gap separating the occupied from the unoccupied states, i.e., the Fermi energy is somewhere in the gap, characterizes the system as a semiconductor or insulator depending on the size of the gap compared to the thermal energy kB T . At room temperature, and for gap energies around 1 eV (typical for semiconductors like Si) free carriers can be thermally excited and become available for electrical transport, while these solids would be insulating at low temperatures. This band structure is responsible also for a characteristic optical property of semiconductors: at low temperature, light can be absorbed only for hν > Egap by creating electron–hole pairs (see Chap. 10).

5.4 LCAO and Tight-Binding Approximation An alternative approach to solve (5.47) starts from the isolated atom, for which we may formulate the Schroedinger equation ! h2 ¯ − ∆ + v(r) φν (r) = Eν φν (r) (5.65) 2m with an eﬀective single-particle potential v(r) of the isolated atom. When arranging the atoms to a solid we can imagine the eﬀective single-particle

140

5 Electrons in a Periodic Potential

potential of the solid emerging from a superposition of the atomic potentials v(r − R0n ), (5.66) Veﬀ (r) = n

while the overlapping atomic orbitals form Bloch functions 1 ik·R0n ψνk (r) = √ e φν (r − R0n ), N n which can be used as a complete set for the expansion ψnk (r) = Cnν ψνk (r).

(5.67)

(5.68)

ν

It is a linear combination of atomic orbitals (LCAO). In contrast to the plane wave expansion used in Sect. 5.3, which exploited the periodicity and weakness of the eﬀective potential, is the LCAO expansion intimately related to the atomic orbitals of valence electrons, which are distorted in the crystalline environment. We make use of the expansion (5.68) in (5.47) to determine the expansion coeﬃcients Cnν . The lhs of (5.47) can be written H ψnk (r) 1 ik·R0n h2 ¯ ∆+ − v(r − R0n′ ) φν (r − R0n ) e Cnν √ 2m N n ν n′ 1 ik·R0n Eν + e v(r − R0n′ ) φν (r − R0n ). (5.69) Cnν √ = N n ν n′ =n

=

To arrive at the last line, the solution of (5.65) has been used for the lattice site R0n . By multiplying from the left with ψν∗′ k and integrating over the crystal volume, one obtains a set of homogeneous coupled linear equations for the Cnν . In contrast to the plane wave expansion the basis set used here, the atomic orbitals φν (r − R0n ), are not orthogonal when centered around diﬀerent lattice sites. Thus, one has to calculate (making use of the lattice periodicity) 1 ik·(R0n −R0 ′ ) n φ∗ν ′ (r − R0n′ )φν (r − R0n )d3 r e N ′ nn 1 ik·R0n φ∗ν ′ (r)φν (r − R0n )d3 r e = N m n 0 eik·Rn Sν ′ ν (R0n ) =: Sν ′ ν (k) (5.70) = δνν ′ + n=0

5.4 LCAO and Tight-Binding Approximation

with the two-center overlap integral Sν ′ ν (R0n ) = φ∗ν ′ (r)φν (r − R0n )d3 r.

141

(5.71)

These integrals depend on the relative position of the centers and on the spatial orientation of the atomic orbitals. Their value decreases with increasing separation of the centers, because the orbitals are localized around the lattice sites. The terms containing the atomic potentials depend in general on three centers 1 ik·(R0n −R0m ) φ∗ν ′ (r − R0m )v(r − R0n′ )φν (r − R0n )d3 r. e N m n ′ n =n

(5.72)

Because the atomic orbitals and the atomic potential decrease rapidly away from their respective centers, these integrals will be small if the three centers are diﬀerent from each other. A reasonable approximation is to consider only terms where two of the three centers coincide. They can be written Kν ′ ν = φ∗ν ′ (r)v(r − R0n )φν (r)d3 r (5.73) n

′

for m = n = n and J

ν′ν

(k) =

e

ik·R0n

n

φ∗ν ′ (r − R0n )v(r − R0n )φν (r)d3 r

(5.74)

for m = n = n′ . As for the two-center overlap integral these expressions are obtained by making use of the periodicity of the lattice. The term Kν ′ ν , which does not depend on the wave vector k, describes the eﬀect of the oﬀcenter atomic potentials on the orbitals at the origin. This crystal field term accounts for the lowering of the rotational symmetry of the isolated atom (according to which atomic orbitals can be classified by angular momentum quantum numbers) in the crystalline environment. It can lead to a removal of degeneracy of atomic orbitals which is known as crystal field splitting (Problem 5.8). The term Jν ′ ν (k) is similar to the overlap integral Sν ′ ν (k) except for the additional potential under the integral. It contributes together with Sν ′ ν (k) to the formation of energy bands. The eigenvalue problem can now be written (5.75) Cnν Hν ′ ν (k) − ESν ′ ν (k) = 0 ν

with the Hamiltonian matrix Hν ′ ν (k) = Eν Sν ′ ν (k) + Kν ′ ν + Jν ′ ν (k)

(5.76)

142

5 Electrons in a Periodic Potential

and the eigenvalues are obtained from the secular equation Hν ′ ν (k) − ESν ′ ν (k) = 0.

(5.77)

This secular problem deviates from the standard form because of the overlapping atomic orbitals. It can be shown, that due to the properties of the overlap matrix, one can find a unitary transformation to reduce (5.77) to the standard form (Problem 5.9). Frequently, the problem is further simplified by considering in the sum over the lattice sites n only nearest neighbors (which is justified for suﬃciently tightly bound atomic orbitals). This tight-binding approximation will be used in the following examples. Supplement: Energy bands in tight-binding approximation: Let us assume in all these examples the simplified form of the secular problem (5.77) with Sν ′ ν (k) ≃ δν ′ ν (see Problem 5.9). 1. The simplest example is the energy band in a cubic lattice that derives from an atomic s orbital. For this case ν = ν ′ = s and the energy band is immediately given by Es (k) = Hss (k) ≃ Es +Jss (k). Note that the s orbital is not degenerate (except for spin) and the crystal field causes only a shift which can be absorbed in the zero of the energy scale. Let us evaluate Jss (k) by summing up the nearest neighbors in a fcc lattice, which under normal conditions is the crystal structure of normal and noble metals. The 12 nearest neighbors are R0n :

a 2

(±1, ±1, 0), (±1, 0, ±1), (0, ±1, ±1) (±1, ∓1, 0), (±1, 0, ∓1), (0, ±1, ∓1).

(5.78)

Due to the spherical symmetry of the s orbital, the matrix element Jss (R0n ) does not depend√on the individual lattice vector but only on the nearest neighbor distance a/ 2 and one finds a Es (k) = Es + 12Jss √ f (k) (5.79) 2 with f (k) =

ky a ky a kz a kz a kx a kx a 1 cos + cos cos + cos cos cos . 3 2 2 2 2 2 2

(5.80)

The function f (k) is characteristic of the crystal structure (Problem 5.10), while the overlap of the atomic orbital together with the strength of the atomic potential, determines the value of the matrix element Jss . The energy band (for obvious reasons called s band) is depicted in Fig. 5.12 for the lines Γ − X √and Γ − L. The width of the band is determined by the matrix element Jss (a/ 2), which for a → ∞ decreases to zero and the band shrinks to the discrete level at Es of the isolated atom. This band exhibits a clear similarity with the lowest band of the free electron dispersion in Fig. 8.6, if we take into account that a periodic potential gives rise to gaps around the points X and L, leading to Fig. 5.7. This similarity indicates the strong influence of the crystal structure on the energy bands.

5.4 LCAO and Tight-Binding Approximation

143

E Es 16 IJsI

L

Γ

k

X

Fig. 5.12. Dispersion of an energy band in an fcc crystal structure deriving from an s orbital

E

z

d3z2- r2

a 2

s

a 2

x,y Γ

X

Fig. 5.13. Polar diagrams of atomic orbitals at nearest neighbors for a fcc transition metal (left) and schematic of energy bands along Γ − X resulting from s − d hybridization (right)

2. Transition metal atoms diﬀer from those of normal (and noble) metals by the successive filling of d electron states, which are energetically close to the s level of the valence shell. In a solid, when bands are formed due to the wave function overlap, these bands deriving from s and d states fall into the same energy range and intersect each other. This situation can well be described in the tight-binding approximation. We choose ν, ν ′ = s, d to account for the configuration of the 3d transition metals (e.g., Fe:4s2 3d6 ). For simplicity we assume here, besides the s orbital (shown as polar plot in Fig. 5.13, left part, placed at the origin) only the d orbital with the spatial dependence 3z 2 − r 2 (in Fig. 5.13 at the nearest neighbor site) and consider the dispersion along k = (0, 0, k). With the simplifying assumption Sν ′ ν = δν ′ ν the secular problem (5.77) takes the form $ $ $ Hss (k) − E Hsd (k) $ $ $ (5.81) $ Hsd (k) Hdd (k) − E $ = 0 with the solutions

144

5 Electrons in a Periodic Potential E± (k) =

)1/2 (1 2 1 . Hss (k) + Hdd (k) ± Hss (k) − Hdd (k) + |Hsd (k)|2 2 4 (5.82)

Here Hss (k) (and similarly Hdd (k) and Hsd (k)) takes the form ka Hss (k) = Es + 4Jss 1 + 2 cos . 2

(5.83)

Without the s − d coupling one gets two similar bands deriving from the s and d orbitals. Their width depends on the corresponding two-center overlap integral Jνν , ν = s, d. In Fig. 5.13 we have depicted in the left part the polar diagrams of two orbitals on neighboring lattice sites. If s orbitals are considered on both sites one obtains a significantly larger overlap than from d orbitals (which are stronger localized to the atomic site). Consequently, the s band is much broader than the d band as depicted by the dashed lines in the right part of Fig. 5.13. The overlap between s and d orbitals (described by Hsd (k)) leads to a coupling and an anti-crossing of these bands as shown by the solid lines in Fig. 5.13. It is known as s − d hybridization and is typical for transition metals. A similar situation can be found for rare earths due to hybridization with f orbitals. The energy bands for some of the 3d transition metals, as obtained from a realistic band structure calculation, are shown for k (001) in Fig. 5.14. Note the crystal field splitting of the d bands, which at the Γ point leads to a twofold (Γ12 ) and a threefold (Γ25′ ) state (see Problem 5.8). The group theoretical notation refers to the irreducible representations of the point group. The crystal structure changes with the filling of the d shell, it is body-centered cubic for V and Fe but face-centered cubic for Co and Cu. With the filling of the d shell the corresponding bands get narrower and shift to lower energy until for Cu (Fig. 5.15) they are all below the Fermi energy. In Fig. 5.15 one easily identifies the broad band deriving from the s-orbitals, which is reminiscent of the free electron band of Fig. 8.6. But one also recognizes the small distortions of this band close to the L point at the Fermi energy, which leads to the necks of the almost spherical Fermi surface (see Fig. 5.9). 3. Hybridization can take place not only due to overlap of atomic orbitals at diﬀerent lattice sites but also due to linear combinations of diﬀerent orbitals Fe (bcc)

V (bcc) 2

4s 3d

energy [Ry]

1.2

3

2

4s 3d

H15

6

Co (fcc) H15

1.0

2

4s 3d

1.2

Cu (fcc)

7

10

4s3d

1.0 H’25

Γ12 Γ’25

Γ2

X’4 H’25

Γ12 Γ’25

0.4

H12

H12

0.2 Γ1 H

X3 X1

0.4 0.2 Γ1

Γ1

Γ

X5 X2

Γ12 Γ’25

Γ

H

X’4 X5 X2 X3 X1

Γ12 Γ’25

Γ1

Γ

X

Γ

X

Fig. 5.14. Energy bands of V, Fe, Co, and Cu for k (001) after [142]. The dashed line indicates the Fermi energy

5.4 LCAO and Tight-Binding Approximation

145

1.5 Cu X1

energy [Ry]

X1

1.0 EF Γ12

0.5

Γ25’

L1

X4’

X4’

X5

L3

L2

X2 X3

L3

X1

L1

X5

Γ12

X2 X3

Γ25’

X1

Γ1

Γ1

0.0 X

Γ

W

L

K X

Γ

Fig. 5.15. Energy bands of Cu after [143]. The dashed line indicates the Fermi energy

z

-

-

pz +

+

-

+

+

-

φ+

+ x,y + px,y - s

r Fig. 5.16. Polar diagram of an sp3 hybrid orbital and its decomposition into s and p contributions (left) and bonding molecular orbital formed by superposition of sp3 orbitals from neighboring lattice sites (right) at the same lattice site. It leads to directed orbitals, which are important in covalent binding. Let us consider the secular problem (5.77) with the orbitals ν, ν ′ = s, px , py , pz but use instead their linear combinations 1 φ s + φ px 2 1 φ s + φ px = 2 1 φ s − φ px = 2 1 = φ s − φ px 2

φ111 = φ1−1−1 φ−11−1 φ−1−11

+ φ py + φ pz − φ py − φ pz + φ py − φ pz

− φ py + φ pz .

(5.84)

A polar diagram of the orbital φ111 and how it is composed of the s and p orbitals is depicted in the left part of Fig. 5.16. Just by accounting for the signs of the latter, it is clear that the resulting orbital has a pronounced positive lobe in the (111) direction, which is the direction towards one of the four nearest neighbors

146

5 Electrons in a Periodic Potential in the diamond structure. Likewise, the other orbitals point to the direction of the other nearest neighbors. These directed orbitals, called sp3 hybrid orbitals, are favorable for establishing a network with tetrahedral coordination based on covalent bonds as realized in the diamond structure. In the same way sp2 hybrid orbitals can be used to establish planar networks based on covalent bonds as in graphite. In this case the pz orbitals sticking out of the plane form π bonding and anti-bonding orbitals, which lead to a peculiar band structure with vanishing gap (Problem 5.11). The band structure of this planar network of carbon atoms (graphene) is the basis also for the electronic structure of carbon nanotubes, which result from rolling up the carbon sheets into cylinders. For the diamond structure, which we pursue here, there are two atoms in the Wigner–Seitz cell with each four directed sp3 hybrid orbitals, which can be superimposed, as depicted in the right part of Fig. 5.16, such that the positive lobes overlap in the nearest neighbor direction or with the opposite signs to form bonding and anti-bonding orbitals, respectively, 1 & τ ' τ φ+ν r + ± φ−ν r − . (5.85) φ± ν (r) = N± 2 2 Here ±τ /2 are the positions of the two atoms in the Wigner–Seitz cell and ±ν refers to the directions of the positive lobes of the sp3 hybrid orbitals, which change sign between the nearest neighbor sites. In order to calculate the band structure, Bloch functions have to be composed of these bonding and antibonding orbitals and in general, an 8 × 8 secular problem has to be solved. The bonding orbitals yield the valence bands, the anti-bonding orbitals the conduction bands. Both groups are separated by an energy gap and we obtain the characteristic band structure of a semiconductor. As an example, the valence bands of Ge calculated by diagonalising the 4 × 4 matrix for the four bonding orbitals with nearest and next nearest neighbor interaction are shown in Fig. 5.17. These bands exhibit a strong similarity with those of Fig. 5.11 obtained from a pseudo-potential calculation. It is interesting also to look at the spatial electron distribution

L 3’

0

Γ25’

energy [eV]

–2

Χ4

–4 –6

L1

–8 –10

L2

–12 –14

L

Χ1 Γ1 Γ

Χ

K

Γ

Fig. 5.17. Valence band dispersion (Ge) from an LCAO calculation with nearest and next nearest neighbor coupling after [144]

5.5 Eﬀective-Mass Approximation

147

Fig. 5.18. Contour plots of the valence electron charge density of Si showing the bond charge: experiment (left), theory (right) after [145]

ρ(r) = −e

n,k,occ.

|ψnk (r)|2

(5.86)

determined by all occupied (valence band) states. A calculated contour plot is shown in Fig. 5.18 for Si together with experimental data obtained from X-ray scattering. The dominant feature is the accumulation of charge between the neighboring lattice sites. It is the bond charge characteristic of covalent binding.

The density of states and the energy band dispersion can be determined experimentally using light–matter interaction (see Chap. 10). Photo-electron spectroscopy (PES) is used to analyse the energy distribution of photo-emitted electrons and yields approximately the density of states from which the excitation takes place, while angular resolved photo-electron spectroscopy (ARPES), which analyses the geometry of the photo-emission process, in addition, allows to map out the energy bands.

5.5 Eﬀective-Mass Approximation Semiconductors diﬀer from metals by the fact that the Fermi energy is in the gap between valence and conduction bands, which is a region with vanishing density of states. At low temperature, there are no free carriers, which could react on a weak perturbation by an applied electric field. At elevated temperatures, the Fermi distribution function is smeared out and thermal population of the lowest conduction band states takes place together with depopulation of the topmost valence band states. The empty valence band states correspond to missing electrons or holes which can be understood as particles with positive charge. In doped semiconductors, the Fermi energy is shifted from the middle of the gap to the impurity levels, which can be close to the conduction band minimum (n doping) or to the valence band maximum (p doping) and the thermal population/depopulation takes place between the impurity states and the nearby band edge. The number of the thermally excited carriers will be small compared with the number of valence electrons and only states close to

148

5 Electrons in a Periodic Potential

the conduction band minimum or valence band maximum become occupied with electrons and holes, respectively. These carriers can follow an applied electric field and carry an electric current. Similarly, optical excitation with photon energies exceeding the gap energy creates electron–hole pairs by lifting electrons from the valence band (leaving holes behind) to the conduction band (see Chap. 10). Thus, the near band edge states determine some of the characteristic properties of semiconductors and deserve special attention. In fact, most of the technological applications with semiconductors (transistors, sensors, lasers) are based on these states. The lowest minimum of the conduction bands is found in Si along the axis from Γ to X (see Fig. 5.11), in Ge at the L point, but it is at the Γ point for most of the compound semiconductors A3 B5 (A2 B6 ) in the zinc blende structure, with one fcc lattice occupied by atoms from the third (second) and the other by atoms from the fifth (sixth) column of the periodic table. For all these tetrahedrally coordinated semiconductors, the valence band maximum is at the Γ point. Si and Ge with valence band maximum and conduction band minimum at diﬀerent points of the Brillouin zone are called indirect gap semiconductors, while those with the band extrema at the same point of the BZ (here the Γ point) are called direct gap semiconductors. The method to describe the dispersion of energy bands around a given point k0 in the BZ has been developed in the early days of semiconductor physics [146, 147]. It is an expansion around this point, which in principle can be extended throughout the Brillouin zone but is used mainly under the condition |k − k0 | ≪ 2π/a. Let us assume the Schroedinger equation (5.37) to be solved for k0 Hψnk0 (r) = En (k0 )ψnk0 (r).

(5.87)

The periodic parts unk0 (r) (with fixed k0 and all n) of the Bloch functions form a complete set of lattice periodic functions. With the Bloch functions at k written as ψnk (r) = ei(k−k0 )·r eik0 ·r unk (r)

(5.88)

the Schroedinger equation reads h ¯2 h2 (k − k0 )2 ¯ − ∆ + Veﬀ (r) + 2m 2m h ¯ + (k − k0 ) · p eik0 ·r unk (r) = En (k)eik0 ·r unk (r). (5.89) m

Now we make use of the complete set in the expansion cnn′ (k − k0 )un′ k0 (r), unk (r) = n′

and insert in (5.89) to find

(5.90)

5.5 Eﬀective-Mass Approximation

n′

149

& h2 (k − k0 )2 ¯ −E cnn′ (k − k0 ) En′ (k0 ) + 2m ' h ¯ + (k − k0 ) · p eik0 ·r un′ k0 (r) = 0. (5.91) m

Here the eigenvalue equation at k0 was applied and the energy En (k) we are looking for is now called E. The last term in the curly brackets containing the momentum operator and the diﬀerence in k vectors can be treated as a perturbation and that is why this concept is called k · p perturbation theory. The expansion coeﬃcients can be found from the set of coupled linear equations n′

& ¯ 2 (k − k0 )2 h − E δnn′ cnn′ (k − k0 ) En′ (k0 ) + 2m ' h ¯ + (k − k0 ) · pnn′ (k0 ) = 0 m

(5.92)

∗ which are obtained from (5.91) by multiplying ψnk (r) from the left and inte0 grating over the crystal volume. Here the matrix elements of the momentum operator pnn′ (k0 ) = u∗nk0 (r)p un′ k0 (r)d3 r (5.93)

establish a coupling between the diﬀerent bands. Solving the secular problem $ $ ¯ 2 (k − k0 )2 h h ¯ $ $ − E δnn′ + (k − k0 ) · pnn′ (k0 )$ = 0 $ En (k0 ) + 2m m

(5.94)

yields the dispersion relations En (k) around k0 for given En (k0 ) and pnn′ (k0 ). Formally, (5.94) is the matrix of a first order perturbation calculation for degenerate states. The oﬀ-diagonal coupling by the momentum matrix elements can be eliminated to any desired order, by matrix perturbation theory as will be outlined below. From Sect. 5.4 we identify the states at the conduction band minimum at k0 = 0 as s anti-bonding states, while those of the valence band maximum are p bonding states. As seen in Fig. 5.17, the dispersion around the valence band maximum is more complex than that around the conduction band miminum. Let us start, therefore, with describing the states at the conduction band edge. Conduction band edge at k0 = 0: We choose the band index n = c for the lowest conduction band with the Bloch function uc0 (r) = r|S deriving from the s anti-bonding orbital. This state is coupled with the p bonding (or valence band) and with the p anti-bonding (or higher conduction band) states (n′ ). Denoting these states |X, |Y , |Z and |X ′ , |Y ′ , |Z ′ , respectively, we have the momentum matrix elements

150

5 Electrons in a Periodic Potential

P =

¯ h ¯ h h ¯ S|px |X = S|py |Y = S|pz |Z m m m

(5.95)

and similar ones for the primed states. These matrix elements are equal due to the symmetry of the diamond or zinc blende structure according to which the cubic axes are equivalent. By eliminating to lowest order the coupling of the s like state at the conduction band minimum to the p like states n′ = c, one arrives at the second order perturbation expression " "2 h2 "S|k · p|n′ " ¯ h2 k 2 ¯ (5.96) + 2 Ec (k) = Ec (0) + 2m m ′ Ec (0) − En′ (0) n =c

and by making use of (5.95) one arrives with " " " " "S|k · p|n′ "2 = "S|px |X"2 kx2 + ky2 + kz2

(5.97)

n′ =X,Y,Z

at the approximate parabolic dispersion around the conduction band minimum " "2 2 "S|px |Xn′ " h2 k 2 ¯ Ec (k) = Ec (0) + . (5.98) 1+ 2m m ′ Ec (0) − En′ (0) n (p)

The summation here is over all p like states. This dispersion tells us, that the electrons occupying these states behave like free electrons but with an eﬀective mass m∗ given by " "2 m 2 "S|px |Xn′ " . (5.99) =1+ m∗ m ′ Ec (0) − En′ (0) n (p)

Depending on the energy denominators, only a few terms need to be considered in the sum. For narrow gap semiconductors the contribution of the topmost valence band dominates, while in general the coupling to the p anti-bonding states needs to be considered. Characteristic values are m∗ /m = 0.067 for GaAs with a band gap Egap = Ec (0) − Ev (0) = 1.52 eV and m∗ /m = 0.0135 for InSb with Egap = 0.25 eV. By applying a magnetic field, a fan chart of Landau levels evolves out of this parabolic dispersion with typical spacings of h ¯ ωc∗ = eB/m∗ and the detection of the cyclotron resonance frequency (see Sect. 4.2) provides the information about the eﬀective mass. We note in passing that higher order terms in the dispersion relation account for the flattening and anisotropy of the energy band away from the Γ point [148] as can be seen in e.g., Fig. 5.12. This nonparabolicity and warping would translate into an energy dependence and anisotropy of the eﬀective mass. Conduction band edge of Si at k0 = (0, 0, k0 ): As mentioned already, Si is an indirect gap semiconductor, the minimum of the conduction band is along the ∆ axis close to the point X (see Fig. 5.11).

5.5 Eﬀective-Mass Approximation

151

The lowest conduction band derives from the sp3 anti-bonding states with ˜ Along the ∆ axis, a k dependent hybridization. Let us denote them by |S. the threefold valence band maximum splits into a rather flat twofold band connected with the states |X and |Y and a band, which evolves from a ˜ Instead of mixing of |Z with the s bonding state, is denoted here as |Z. (5.95) we now have for the momentum matrix elements PX =

h ˜ ¯ ¯ ˜ h ¯ ˜ h ˜ = PZ , S|px |X = S|p S|pz |Z y |Y = m m m

and the perturbation series reads with k′ = k − k0 " " ˜ ′ · p|n′ "2 h2 "S|k ¯ h2 k ′2 ¯ ′ . + 2 Ec (k ) = Ec (k0 ) + 2m m ′ Ec (k0 ) − En′ (k0 )

(5.100)

(5.101)

n =c

The inequality in (5.100), caused by the reduced symmetry of the group of the wave vector k0 , leads to the anisotropic dispersion relation (with k′ = (kx , ky , kz − k0 )) ′ h2 kx2 + ky2 ¯ kz2 ′ (5.102) + ∗ Ec (k ) = Ec (k0 ) + 2 m∗⊥ m with m∗ and m∗⊥ being the masses parallel and perpendicular to the ∆ axis. For Si, the conduction band minimum is at k0 = 0.85 in units of 2π/a. Due to the cubic symmetry there are minima also along the other equivalent directions in the Brillouin zone, thus the conduction band of Si has six minima, which (thermally or by doping) become equally populated. In the presence of a magnetic field, fan charts of Landau levels evolve from these minima, which depend on the orientation of the magnetic field with respect to the inverse mass ellipsoid and so does the cyclotron mass (Problem 5.12). Valence band maximum at k0 = 0: The valence band maximum of semiconductors with diamond or zinc blende structure derives from the p bonding orbitals and is (without spin) threefold degenerate. The energy dispersion is obtained from the 3 × 3 determinant (ν, ν ′ = X, Y, Z) $ h2 ν|k · p|nn|k · p|ν ′ $ ¯ ¯ 2 k2 h $ $ − E δνν ′ + 2 $ = 0. (5.103) $ Ev (0) + 2m m E (0) − E (0) ν n ′ n=ν,ν

Symmetry arguments, for which we refer to the literature [146], allow to write the terms bilinear in the components of the wave vector as a 3 × 3 matrix (known as the Shockley3 matrix [147]) 3

William B. Shockley 1910–1989, received the Noble prize in physics 1956 jointly with J. Bardeen and W.H. Brattain.

152

5 Electrons in a Periodic Potential

⎛

⎞ Lkx2 + M (ky2 + kz2 ) N kx ky N kx kz ⎠ N kx ky Lky2 + M (kx2 + kz2 ) N ky kz M=⎝ N kx kz N ky kz Lkz2 + M (kx2 + ky2 )

(5.104)

with L, M, N ∼

ν|pα |nn|pβ |ν ′ , Eν (0) − En (0) ′

(5.105)

n=ν,ν

where diﬀerent intermediate levels contribute to L, M, and N . The secular problem M − Eδν,ν ′ = 0 is easily solved for special directions of k. For the ∆ axis or k = (0, 0, k) one finds Ev1 (k) = Lk 2

and Ev2 (k) = M k 2 ,

(5.106)

where the second solution is twofold, which by comparison with Fig. 5.17 can be identified with √ the upper branch of the valence band. For the Λ axis with k = (k, k, k)/ 3 the solutions are Ev1 (k) =

L + 2M − 2N 2 L + 2M + N 2 k and Ev2 (k) = k . 3 3

(5.107)

Again, the second solution is twofold and can be identified with the upper branch for this direction. A characteristic feature is the diﬀerent curvatures of the bands for the two solutions: the twofold band with the small curvature, corresponding to a large mass, is the heavy hole band, while the non-degenerate band with the larger curvature is the light hole band. The other feature is that these curvatures depend on the direction of k and the surfaces of constant energy are warped. We note in passing that for all the other directions the secular problem has three diﬀerent solutions. These degeneracies and the anisotropy of the valence bands are the same as those of the acoustic phonon branches discussed in Sect. 3.4. This is a consequence of the group of the wave vector, which contains at least threefold rotations only for the ∆ and Λ axes. In general, energy bands can be calculated numerically by diagonalizing (5.94) for a finite set of states, provided, the separation of their energy levels at k0 and the momentum matrix elements are known. For near band-edge states in direct gap semiconductors around the Γ point, a multi-band k · p model is frequently in use, which comprises five bands or (including spin) 14 states [148–150].

5.6 Subbands in Semiconductor Quantum Structures While in the previous sections of this chapter, we have looked at the electron states in three-dimensional solids, we shall now consider systems with reduced dimensionality, which are in the focus of interest since the development of the planar semiconductor technology of Si based field eﬀect transistors

5.6 Subbands in Semiconductor Quantum Structures

153

(MOSFET or metal oxide Si field eﬀect transistor)[151–153]. Here, we would like to describe a semiconductor heterostructure, which can be obtained by epitaxial growth of one semiconductor (say AlAs) onto another one (GaAs). Both systems have the same crystal structure (zinc blende) and almost the same lattice constant, i.e., across the interface it is only the chemical nature of the atoms that changes. Let the growth direction be along z and the interface at z = 0 extend in the x − y plane. The electronic structure of this system is determined by the bulk band structure of the two materials with their respective energy gaps and eﬀective masses. In particular, the band edges (the minima of the conduction band and the maxima of the valence band) are shifted against each other and the interface appears as a step-like potential for carriers at these edges. Correspondingly, a double heterostructure AlAs/GaAs/AlAs would represent a confining square well potential which for electrons can be described by the conduction band profile ⎧ z<0 ⎨ 0 0L

where the minimum of the energy band in AlAs is the zero point of the energy scale and V0 is the conduction band oﬀset between AlAs and GaAs. As it is only a few 100 meV, the confined states, being close to the conduction band minimum, can be described in the eﬀective-mass approximation with the Hamiltonian h2 ¯ HEMA = − ∆ + Ec (z). (5.109) 2m∗ The eigenfunctions of this Hamiltonian φn (k r) = eik ·r ζn (z) and the eigenvalues En (k ) = En +

¯ 2 k 2 h

(5.110)

(5.111) 2m∗ indicate the quantization with energies En due to the confinement potential in growth direction and the free motion parallel to the interface: electrons occupying these states would be free carriers in two dimensions. The quantum number n denotes the subbands evolving from the conduction band of the three-dimensional band structure. In both, the single heterostructure and in the quantum well, free carriers can be introduced by doping the AlAs barrier with Si. The Si atoms create additional states close to the conduction band minimum (donor states) with a small binding energy. Due to the negative band oﬀset, the donor electrons cross the interface and reside in the GaAs layer, while leaving a positively charged center behind in AlAs. This charge separation gives rise to an electrostatic potential, which is superimposed on the band edge profile. For a single

154

5 Electrons in a Periodic Potential

E

++

AlGaAs:Si

E + 2 + E 1

E0

AlGaAs GaAs

EF Ec

z

Fig. 5.19. Semiconductor heterostructure (left) and conduction band profile across the interface with subband levels and Fermi energy (right)

heterostructure this leads to the picture shown in Fig. 5.19 with an almost triangular potential at the interface, which now confines the carriers. The energy spectrum is similar to (5.111) but the subband energies are obtained here from a self-consistent solution of the one-dimensional Schroedinger equation ! h2 2 ¯ − ∗ ∂z + V (z) ζn (z) = En ζn (z) (5.112) 2m with the potential V (z) = V0 Θ(−z) + VH (z) + Vxc (z) and Poisson’s equation ∂z2 VH (z) =

e ̺(z). εε0

(5.113)

Here, VH (z) is the Hartree potential, determined by the charge distribution ̺(z) of ionized donors in the barrier (AlAs) and of the electrons occupying subband states. We recognize (5.112) as the one-dimensional analogue to the Schroedinger equation (5.5) for the three-dimensional case with an eﬀective single-particle potential, which contains also an exchange–correlation part Vxc (z). We note in passing that a more accurate description would take into account the change of the eﬀective mass across the interface. The eﬀective mass determines also the density of states, which (without nonparabolicity eﬀects) is a constant (see Problem 4.3 and Fig. 5.20, left part). The number of subbands with occupied states depend on the areal density, which can be controlled by the external gate voltage. The situation, when only states in the lowest subband are occupied, is called the electric quantum limit. A more detailed description shows, that the spin-degeneracy of the subbands is removed due to a spin–orbit coupling caused by the asymmetric confinement potential at the interface [153] (Problem 5.13). It can be tuned by an external gate voltage and is important in manipulating the spin dynamics in these structures (Rashba eﬀect). The eﬀective-mass Hamiltonian (5.109) is obtained from the conduction band dispersion (5.98) by replacing k 2 → −∆, which leads to the operator of

5.6 Subbands in Semiconductor Quantum Structures

E

E

155

N=4 3 2

E1

1

EF

EF

0

E0

E0 kII ,D(E)

B

Fig. 5.20. Subband dispersion and density of states for the lowest two subbands (left) and fan chart of Landau levels evolving from the lowest subband (right), spin splitting is suppressed

kinetic energy for carriers close to the conduction band minimum. The corresponding description of confined holes is more complex because of the valence band degeneracy and would lead (without spin) to three coupled Schroedinger equations, which for k = 0, decouple and yield subband states for heavy and light holes. Quantitative subband calculations including spin and nonparabolicity eﬀects are performed within the multiband envelope function approximation, which is based on the k · p theory [152–154]. Due to the confinement potential, the energy gap, i.e., the energy diﬀerence between the lowest electron and the highest hole subband state, is changed as compared with the bulk material. The possible choices of well and barrier material with their respective electron and hole masses and of the width L of the quantum well allow one to design materials with defined gap energies. This band gap engineering is exploited in optical devices (see Chap. 10). In heterostructures, the spatial separation between the free carriers in the confining potential and the ionized donors in the barriers (usually enlarged by growing a spacer layer at the interface) leads to improved mobilities of the carriers at low temperatures (when scattering with ionized impurities is dominating, see Chap. 9). Under these conditions, characteristic transport lengths become comparable or even larger than the lateral system size and phenomena of mesoscopic physics can be investigated [155–160]. Magnetotransport experiments in these systems have led to the discovery of the quantum Hall eﬀects and composite fermions (see Chap. 7). A magnetic field applied perpendicular to the plane of the two-dimensional electron system leads to a complete discretization of the energy spectrum. The subbands split into highly degenerate Landau levels (see Sect. 4.2) and the filling factor, i.e., the number of occupied Landau levels changes with the magnetic field. The ultimate case, when at suﬃciently high magnetic field all electrons can be accommodated in the lowest Landau level of the lowest

156

5 Electrons in a Periodic Potential h/2e2 12 RH

B = 18.9T T = 1.5 K

10 2

RH,RX /kΩ

h/3e 8

h/4e2 6 Rx

h/6e2

4

h/8e2 2

n=1

n=2

0 0

10

20

30

Vg/V

Fig. 5.21. Longitudinal and Hall resistance vs. magnetic field of a Si-MOSFET showing the quantum Hall eﬀect after [161]

subband, is known as the magnetic quantum limit (see Problem 4.4). Upon reducing the magnetic field starting from this limit, a point will be reached when all states in the lowest Landau level are occupied, which corresponds to filling factor ν = 2 (due to spin degeneracy). With further reduction of the magnetic field, the Fermi energy moves into the next higher Landau level. Thus, by sweeping the magnetic field, the Fermi energy will jump at field values corresponding to (even) integer filling factors (Fig. 5.20, right). When at suﬃciently high magnetic fields (and low temperatures) the Zeeman splitting can be resolved, these jumps take place at all (even and odd) integer filling factors. In a MOSFET or a gated heterostructure, it is possible to change the carrier density by varying the gate voltage. Thus, for a fixed magnetic field, the Fermi energy can be shifted through the Landau levels by changing the filling factor. At integer filling factors and suﬃciently low temperature, magnetotransport data show striking deviations from the classical behavior of the Drude model (Fig. 5.21): the longitudinal resistance (or conductance) vanishes (as typical for an insulator) and the Hall resistance (or conductance) exhibits plateaus, which to extremely high accuracy are inverse integer multiples of e2 /h [161]. This is known as the integer quantum Hall eﬀect, discovered by von Klitzing4 in 1980. The high accuracy of the plateau values gave rise to a new definition of the resistance, normally defined by precision measurements in a quantum Hall experiment of h/e2 , which corresponds to the Klitzing constant RK−90 = 25812, 807 Ω introduced by the Meter Convention in 1990. An introduction to the theory of the integer Quantum Hall eﬀect is [162]. 4

Klaus von Klitzing *1943, Noble prize in physics 1985.

Problems

157

In 1983, Tsui and St¨ ormer5 discovered signatures of the quantum Hall eﬀect at very low temperatures and high magnetic fields at the fractional filling ν = 1/3. The confirmation of this signature complemented by the discovery of a whole family of so-called fractional quantum Hall states (see Sect. 7.6) in the following years has stimulated much experimental and theoretical work on electron systems under such extreme conditions. The fractional quantum Hall eﬀect is now understood to be dominated by electron–electron interaction (see Chap. 7) while the integer quantum Hall eﬀect is essentially a single-particle eﬀect caused by disorder (see Chap. 9).

Problems 5.1 Derive the Hartree–Fock equations as Euler–Lagrange equations of the variational problem (5.7) with the Slater determinant as ground state wave function. Note that the single-particle wave functions depend on space and spin coordinates. 5.2 Show that the exchange potential HF n ¯ (r, r′ ) 3 ′ Vx (r) = −e2 d r (5.114) |r − r ′ | with the averaged nonlocal HF or exchange density (5.15) reduces for free electrons to the so-called Slater exchange potential 3e2 kF (n(r)). 2π Compare it with the exchange potential in the LDA Vx,Slater (r) = −

(5.115)

d nǫx (n) n=n(r) , (5.116) dn where ǫx (n) is the exchange energy per electron for the jellium model with density n. 5.3 Verify that each energy band can accommodate one electron of either spin per unit cell. Remember the corresponding statement concerning phonon modes. 5.4 A point at the Brillouin zone boundary can be characterized by wave vectors k and k′ drawn from diﬀerent reciprocal lattice vectors. Give the relation between k and k′ and verify that the condition of degeneracy of the corresponding free electron states, i.e., of having the same energy, is consistent with the condition for Bragg reflection. VxLDA (r) =

5

Daniel C. Tsui *1939, Horst L. St¨ ormer *1949, shared the Noble prize in Physics 1998 with R.B. Laughlin.

158

5 Electrons in a Periodic Potential

5.5 Calculate and plot the free electron energy dispersion for the square lattice along the lines Γ → M, M → X, and X → Γ for the lowest three bands (the reciprocal lattice vectors can be used as band index). Indicate the position of the Fermi energy, if each atom in the unit cell contributes one, two, or three electrons. Calculate the dispersion for the two lowest bands in the presence of a periodic potential and discuss the modification of the Fermi circle with radius kF = π/a by this potential. 5.6 Calculate the free electron energy bands of a fcc lattice as shown in Fig. 8.6. Find out the degeneracies of these bands. 5.7 Compare the band structures of Al, Si, and GaAs in the context of the almost free electron or pseudo-potential approximation. What is common (diﬀerent) in the free electron picture and which of the Fourier coeﬃcients of the periodic potential are responsible for opening of gaps? Which details of the band structure can be understood from this comparison? 5.8 The atomic d states are fivefold degenerate (without spin). Calculate the crystal field splitting of these states in a cubic lattice. Is there a diﬀerence between the sc, bcc, and fcc crystal structures? 5.9 Show that the secular problem (5.77) can be reduced to the standard form $ $ $˜ $ (5.117) $Hν ′ ν (k) − Eν δν ′ ν (k)$ = 0

by making use of the properties of the overlap matrix Sν ′ ν (k). Give the ˜ ν ′ ν (k) in terms of Hν ′ ν (k) and Sν ′ ν (k). explicit expression of H 5.10 Calculate the dispersion of an energy band deriving from an s orbital in a sc and bcc lattice for the lines Γ − X and Γ − L. Compare with each other and with the result for the fcc lattice. How does the width of the energy band depend on the nearest neighbor configuration? 5.11 Graphene is a single sheet of carbon atoms arranged in a (twodimensional) hexagonal lattice with primitive lattice vectors in the (x, y) plane √ √ 3 3 3 3 a1 = d , a2 = d − (5.118) , , 2 2 2 2 and two carbon atoms in each cell at τ 1 = (0, d) and τ 1 = (0, 2d) with the nearest neighbor distance d. The s, px , and py orbitals can be combined to sp2 hybrid orbitals directed towards the nearest neighbors, while the pz orbitals stick out of the lattice plane and form πz bonding and anti-bonding orbitals. Formulate and solve the eigenvalue problem for graphene using LCAO for the πz orbitals in the tight-binding approximation and discuss the resulting energy bands. 5.12 Find the expression for the cyclotron mass of a spheroidal energy surface E(k) =

1 2 1 2 k = const. kx + ky2 + mt ml z

(5.119)

Problems

159

in dependence on the orientation of the magnetic field in the xz plane. Make use of the Peierls substitution and use a vector potential with Ax = Az = 0. Set up and solve the equations of motion in the xz plane. 5.13 Solve the eigenvalue problem of free two-dimensional electrons in the xy plane in the presence of a spin-orbit coupling HSO = α(k × ∇V ) · σ

(5.120)

due to the confinement potential V in z direction, i.e., ∇V (001). Calculate and visualize the expectation values of the spin operator σ in dependence on the direction of k. Identify the Kramers pairs.

6 Spin Waves: Magnons

The electron spin, which does not explicitly appear in the N -particle Hamiltonian of the solid (except for the spin-orbit coupling and Zeeman terms), will be in the focus of this chapter. In Chaps. 4 and 5 we have addressed already the relevance of spin in connection with magnetic properties, which shall be studied now in more detail. The issue here will be to consider the interacting electron system with dominating exchange interaction, which leads to a spin-ordered ground state, and to describe elementary excitations out of this ground state: spin waves or (in quantized form) magnons. In several aspects, spin dynamics is similar to lattice dynamics (see Chap. 3) with the masses coupled by springs, now being replaced with spins (or their magnetic moments) coupled with exchange interaction. Depending on the complexity of the crystal structures on which these spin systems are realized, their spin or magnetic order can be ferromagnetic, anti-ferromagnetic, ferrimagnetic, or anti-ferrimagnetic [163]. Most standard textbooks on Solid State Theory contain a chapter on spin waves or magnons and magnetic properties, but there are also special review articles [164, 165] and monographs [116, 166–172] on these topics. In solids with disorder, e.g. due to alloying, the magnetic order takes the form of spin glasses [16]. Theoretical concepts developed for spin glasses have turned out to be useful also for neural networks [173]. A variant of magnetic order in disordered solids are diluted magnetic semiconductors of which A3 B5 with Mn ions randomly replacing the A3 atoms show ferromagnetic order (see [174–176]). The latter materials have gained much interest in connection with spintronics [177, 178], a concept of electronics using the electron spin rather than its charge. Magnetic order exists as a ground state property only below some critical temperature at which a phase transition takes place. It can be described using the molecular or mean field approximation. For spin excitations at the surface of magnetic solids (surface magnons) we refer to [179, 180].

U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 6,

162

6 Spin Waves: Magnons

6.1 Preliminaries In Chap. 4 we have seen how magnetic properties of free electrons can be understood in terms of the single-particle picture and the Pauli principle. The Zeeman term causes a shift of the density of states of up and down spin electrons against each other (see Fig. 4.7). Occupying these states according to the Fermi–Dirac distribution function, leads to a magnetic moment determined by the excess of majority spins: this is the Pauli spin paramagnetism. On the other hand, for the case without external magnetic field, we have found a lowering of the single-particle and ground state energies due to the exchange term in the HF approximation, which becomes eﬀective for electrons with parallel spins. This eﬀect is expected to get stronger with an increase of the number N↑,↓ of aligned spins, i.e. with a spin polarization N↑ − N↓ of the system accompanied by a magnetization M = gμB (N↑ − N↓ ). Eventually, one may ask for a condition under which a state with all spins aligned could be the system ground state. In this consideration, we have completely neglected the influence of the periodic potential. Let us look, therefore, for a moment at the band structure of ferromagnetic Fe in Fig. 6.1 as an example. We note the characteristic features resulting from s-d hybridization discussed in Sect. 5.4, but find that for the ferromagnetic state the energy bands of up and down spin electrons diﬀer from each other, ǫnk↑ = ǫnk↓ , as is typical for Bloch electrons with incomplete d shells. A closer inspection shows that the bands of spin-up and spin-down electrons are very similar but shifted against each other, as can be seen in the density of states (Fig. 6.2) by comparing the position of the pronounced structures, caused by van Hove singularities, for majority and minority spins. This reminds us of the free electron case with external magnetic field, but

EF

energy [Ry]

0.8

0.6

0.4 Fe minority spin

0.2

majority spin 0

Γ

H

N P wave vector k

Γ

N

Fig. 6.1. Energy bands of ferromagnetic Fe after [181]

6.1 Preliminaries

163

3.5

6 Fe

3.0

5

2.5

4

2.0 3

Integrated DOS

1

1.0 0.5

majority spin

0

0 1

minority spin

0.5 1.0

2

1.5

3

DOS [states / eV - atom]

1.5

2

2.0 EF

4

2.5

5 –10 –8 –6 –4 –2 0 Energy E - EF [eV]

2

3.0

Fig. 6.2. Density of states (solid lines) and integrated density of states (dotted lines) for the energy bands of ferromagnetic Fe of Fig. 6.1 (after [181])

here, the shift is determined by the spin–polarization that exists even without an external magnetic field, i.e. there is a spontaneous magnetization below a critical temperature. Energy bands as those of Fig. 6.1 result from an extension of the DFT-LDA to the SDFT-LSDA (spin–density functional theory in the local spin–density approximation) [123, 182]. Turning back to the free electron system of Chap. 4, we recollect the expression for the ground state energy in HF approximation (without magnetic field) with each state with |k| ≤ kF occupied by a pair of ↑, ↓ spins (see (4.91)) HF = E0↑↓

h2 2 3 3 ¯ N N e2 kF . kF − 2ε 5 2m 16π 0 a

(6.1)

b

Remember the origin of the second term; it is the exchange part of the Coulomb interaction which results from parallel spins. Instead of occupying each state in the Fermi sphere with a pair of ↑, ↓ electrons one could align all spins to gain exchange energy, but this can be done only at the cost of increasing the kinetic energy because the size of the Fermi sphere has to be doubled, i.e. by now occupying all states with |k| ≤ 21/3 kF (see Fig. 6.3). The HF ground state energy for this ferromagnetic order (the completely spin-polarized free electron gas) is HF E0↑↑ = 22/3 akF2 − 21/3 bkF

(6.2)

164

6 Spin Waves: Magnons

kF 1/3

2 kF Fig. 6.3. Fermi spheres for a spin-unpolarized and a completely spin-polarized electron system HF HF < E0↑↓ , kF has to fulfill the with a and b from (6.1). In order to have E0↑↑ condition

kF <

b 1.1 −1 21/3 − 1 b = 0.44 = a a π B 22/3 − 1 a

(6.3)

with the Bohr radius aB or, with kF = (9π/4)1/3 1/rs aB , this yields for the density parameter !1/3 π 9π rs > = 5.45. (6.4) 4 1.1 According to this estimate the ferromagnetic ground state should be favored in diluted electron systems with rs > 5.45, which is still in the region of electron densities in metals (2 < rs < 6). However, this estimate should not be taken for a proof of the existence of a ferromagnetic ground state as it does not include the eﬀects of correlation or of the periodic potential. For a more rigorous treatment, we start from the N -electron Hamiltonian (5.2) that includes the periodic potential. It can be written with the help of (4.76) and (4.77) in terms of fermion operators as 1 H= ǫα c†α cα + (6.5) Vαββ ′ α′ c†α c†β cα′ cβ ′ , 2 ′ ′ α αβα β

′

′

where α, β, α , and β are complete sets of single-particle quantum numbers, which here are those of Bloch states, e.g. α = nkσ and ǫα are the corresponding single-particle energy values. The potential matrix elements e2 ψβ ′ (x)ψα′ (x′ ) (6.6) Vαββ ′ α′ = dx dx′ ψα† (x)ψβ† (x′ ) 4πε0 |r − r′ | can be expressed in terms of the single-electron wave functions ψα (x) (Problem 6.1), where x stands for space and spin coordinates. Summing up

6.1 Preliminaries

165

the spin coordinates results in getting the same spin quantum numbers for the states α, β ′ and for the states β, α′ . In the following, we shall separate the Hamiltonian (6.5) according to H = Hsp + Hint into its single-particle and interaction parts. Magnetism is known, from the introductory courses on solid state physics [29–31], to arise in materials with incomplete d or f shells. The corresponding electrons form narrow bands (see e.g. Fig. 6.1), for which the free electron picture hardly applies. We take this situation into account by switching to the LCAO (or tight-binding) approximation introduced in Sect. 5.4. In this picture, contrary to Fig. 6.3, the electron spin is not attached to delocalized electrons, and magnetism results from the magnetic moments of the total spin carried by local atomic d or f orbitals. Let us consider electrons in an energy band deriving from one such orbital. The Bloch function can be expressed (as in Sect. 5.5) by 1 ik·R e φσ (r − R), ψkσ (r) = √ N R

(6.7)

where we have dropped the band index to simplify notation while the index σ refers to the spin eigenstate. It is advantageous here to adopt the Wannier1 representation with localized orbitals, which are orthogonal for diﬀerent lattice sites, although they may still overlap (Problem 6.2). This allows one to switch with 1 ik·R ckσ = √ e cRσ (6.8) N R to fermion operators and write the single-particle term in the form ǫk c†kσ ckσ Hsp = kσ

=

1 ′ ǫk eik·(R−R ) c†R′ σ cRσ = tRR′ c†R′ σ cRσ . N RR′ σ k RR′ σ

(6.9)

tRR′

The dispersion of the energy band ǫk (assumed to be independent of σ) is now expressed in terms of the hopping or transfer matrix elements tRR′ , R = R′ which is a two-center integral of the type given in (5.74), while the term with R = R′ gives the atomic energy level from which the band derives (see Problem 6.2) . Likewise the electron–electron interaction takes the form (Problem 6.3) Hint = 1

1 VR1 R2 R′1 R′2 c†R1 σ c†R2 σ′ cR′2 σ′ cR′1 σ 2 RRσ 1 2 R′ R′ σ′ 1 2

Gregory Hugh Wannier 1911–1983.

(6.10)

166

6 Spin Waves: Magnons

with the interaction matrix element being now the four-center integral d3 r d3 r′ φ∗ (r − R1 )φ∗ (r ′ − R2 ) VR1 R2 R′1 R′2 = ×

e2 φ(r − R′1 )φ(r ′ − R′2 ). 4πε0 |r − r ′ |

(6.11)

Note, that the Wannier (or atomic) orbitals and the single particle energies are assumed to be independent of the spin quantum number σ. This means to neglect all spin-dependent eﬀects deriving from spin–orbit coupling and from the electrons in all other occupied bands of the solid, while here all spin related eﬀects derive from the interaction term.

6.2 The Heisenberg Hamiltonian The extreme case of strongly localized orbitals would lead to vanishing hopping matrix elements and to a single-particle part of the Hamiltonian, just counting the occupation of the sites multiplied by the atomic level energy. Contrary to the free-electron case (see Fig. 6.3), in a flat band there is no increase in the kinetic energy when all spins are aligned, which is the favorite configuration with respect to the exchange interaction. But this ferromagnetic configuration, with each site being occupied by a single electron with given spin (Fig. 6.4), minimizes the Coulomb repulsion also, because each two electrons are separated in space as much as possible in the given crystal structure. Thus, this configuration can be considered as the ground state. We have to keep in mind here that the assumption of strongly localized atomic orbitals does not apply to metallic ferromagnets such as the transition metals, which form d bands with a width of the order of one eV, (see Figs. 5.14 and 6.1) due to the overlap between nearest neighbors in a close-packed crystal structure (bcc or fcc). This overlap is, however, essentially reduced in transition metal compounds like MnAs, EuS, EuS with a larger spacing between the metallic ions in a lattice with basis. They appear as ferromagnetic insulators and are the materials, to which the model of Fig. 6.4 applies. With respect to elementary excitations, this ferromagnetic ground state plays the same role as the filled Fermi sphere for the free electrons. As in Chap. 4, when considering the HF approximation the interaction term in application to this ground state, we can distinguish between a direct and an

R –d

R

R+d

Fig. 6.4. Ferromagnetic configuration of spins on a linear chain

6.2 The Heisenberg Hamiltonian

167

exchange term. The direct process requires R1 = R′1 = R and R2 = R′2 = R′ and can be expressed as Hd =

† † 1 VRR′ RR′ cRσ cR′ σ′ cR′ σ′ cRσ . 2 ′ ′

(6.12)

σσ

RR

When applied to the ferromagnetic ground state, it removes two electrons at sites R and R′ with spins σ and σ ′ , respectively, and puts them back where they are taken from. It is clear that this requires R = R′ , because at each lattice site, there is only one electron that can be removed, and the operator part can be transformed to yield † † cRσ cRσ (6.13) cR′ σ′ cR′ σ′ = 1. σ′

σ

Thus, we find that Hd is the energy of the electrostatic interaction between the charge densities located around R and R′ , i.e. of the electronic configuration associated with the spin configuration of Fig. 6.4. The exchange term is obtained with R1 = R′2 = R and R2 = R′1 = R′ and reads † † 1 VRR′ R′ R (6.14) cRσ cR′ σ′ cRσ′ cR′ σ . Hx = 2 ′ ′ σσ

RR

The operator part can be transformed (again with R = R′ ): † † cRσ cR′ σ′ cRσ′ cR′ σ = − c†R↑ cR↑ c†R′ ↑ cR′ ↑ − c†R↓ cR↓ c†R′ ↓ cR′ ↓ σσ′

− c†R↑ cR↓ c†R′ ↓ cR′ ↑ − c†R↓ cR↑ c†R′ ↑ cR′ ↓ . (6.15)

The diﬀerent combinations of creation and annihilation operators at a given site have the following meaning with respect to the site R (here we drop the site index): c†↑ c↑ , c†↓ c↓ : count the ↑, ↓ electrons

c†↑ c↑ − c†↓ c↓ : counts the diﬀerence between ↑ and ↓ electrons c†↑ c↓ , c†↓ c↑ : cause spin flips.

Taking into account the commutation relation [c†↑ c↓ , c†↓ c↑ ] = c†↑ c↑ − c†↓ c↓

(6.16)

we recognize that the operators c†↑ c↓ , c†↓ c↑ , and c†↑ c↑ − c†↓ c↓ at each site R fulfill the commutation rules of the su(2) algebra known from the angular momenta (Problem 6.4): [S i , S j ] = iεijk S k ,

i, j, k = x, y, z

cycl. perm.

(6.17)

168

6 Spin Waves: Magnons

with the Levi–Civita symbol εijk , and we may identify S + = S x + iS y = c†↑ c↓ , S − = S x − iS y = c†↓ c↑ , S z =

1 † (c c↑ − c†↓ c↓ ). (6.18) 2 ↑

Thus, it is possible to replace the annihilation and creation operators by spin y x z vector operators S R = (SR , SR , SR ) at each lattice site. This is achieved by † † adding and subtracting (cR↑ cR↑ cR′ ↓ cR′ ↓ + c†R↓ cR↓ c†R′ ↑ cR′ ↑ )/2 on the rhs of (6.15) to obtain for the operator part of the exchange term † † + − − + z z cRσ cR′ σ′ cRσ′ cR′ σ = − SR SR′ + SR SR′ − 2SR SR′ σσ′

−

1 † c cRσ c†R′ σ′ cR′ σ′ . 2 ′ Rσ

(6.19)

σσ

The first two terms on the rhs can be written in the form + − − + z z SR′ + SR SR′ + 2SR SR′ = 2S R · S R′ SR

(6.20)

while the last term can be combined with the direct term, and we find for the interaction part, Hint of the Hamiltonian, the form (6.21) JRR′ S R · S R′ Hint = − RR′ R=R′

which is the Heisenberg2 Hamiltonian with the exchange integral e2 φ∗ (r − R)φ(r − R′ )φ∗ (r ′ − R′ )φ(r ′ − R) . (6.22) JRR′ = d3 r d3 r ′ 4πε0 |r − r′ | It can be shown that JRR′ > 0. We add the Zeeman term and replace the site index R by a number index i (which can be understood as replacing the 3-dimensional Bravais lattice with a linear chain but is meant also as short notation for Ri ) to write the spin or Heisenberg Hamiltonian in the form Siz . (6.23) Jij S i · S j − gμB Hext Hspin = − i,j i=j

i

This Hamiltonian is the canonical starting point in the theory of magnetism [183]. In spite of its simplicity, it bears fundamental properties, one of which is related to symmetry. Without external magnetic field, (6.23) is spherically symmetric, however, its ground state, the ferromagnetic configuration, has only axial symmetry. This situation is known as spontaneous symmetry breaking: Due to interaction between the spins and in order to gain energy, 2

Werner Heisenberg 1901–1976, Nobel prize in physics 1932.

6.2 The Heisenberg Hamiltonian

169

the system prefers a configuration with lower symmetry than that of the Hamiltonian. The further task of this chapter will be to describe elementary excitations out of this ground state. There exist several model Hamiltonians related to the Heisenberg Hamiltonian, which have been under investigation in the context of magnetic properties [171]: 1. the anisotropic Heisenberg model Hspin = −

i,j i=j

Jij (Six Sjx + Siy Sjy ) + J¯ij Siz Sjz , Jij = J¯ij ,

(6.24)

2. the Ising model (with Jij = 0)

J¯ij Siz Sjz ,

(6.25)

Jij (Six Sjx + Siy Sjy ).

(6.26)

HIsing = −

i,j i=j

3. and the XY model (J¯ij = 0) HXY = −

i,j i=j

The Ising and the XY models are used in statistical mechanics because of their simplicity compared to the Heisenberg model, while still representing reasonable approximations for real spin systems. As it turns out, however, the low-dimensional spin models (except the two-dimensional Ising model) fall short of describing the ferromagnetic phase transition (see Sect. 6.5). Instead the XY model produces a transition between two disordered phases of diﬀerent topology. This Kosterlitz–Thouless transition is caused by the appearance of spin vortices (see [171]). In all these models, the exchange integrals appear as parameters. The microscopic mechanism, although it is always the exchange interaction, can be quite diﬀerent depending on the lattice structure of the solid. The easiest case is that of the transition metals, where the exchange takes place directly between the d orbitals on nearest neighbor lattice sites of a closepacked structure like bcc and fcc. Therefore, it is called direct exchange to distinguish it from indirect or super-exchange in compounds of elements with incomplete d shells, as e.g. the ferromagnetic insulators MnAs, EuO, and EuS. In these solids forming lattices with basis, the orbitals which carry a magnetic moment, are separated by the non-magnetic ions and the exchange coupling has to be mediated by the orbitals of the intervening ions. A similar situation is found in metallic ferromagnetic compounds, where the exchange

170

6 Spin Waves: Magnons

coupling between distant magnetic ions is mediated by the free electrons. This is the Rudermann–Kittel exchange, which exists together with the direct exchange interaction and is frequently the dominating mechanism. Depending on the mechanism and the lattice configuration, the exchange constant can be positive (as anticipated so far), which leads to the ferromagnetic order, but also negative. In the latter case, neighboring spins align anti-parallel and exhibit the anti-ferromagnetic ordering. More complex crystal structures with sublattices accomodating diﬀerent spin carrying orbitals, i.e. the spin operator S depends on the sublattice, lead to the ferrimagnetic or anti-ferrimagnetic order.

6.3 Spin Waves in Ferromagnets In this Section, we want to describe low-energy excitations out of the ferromagnetic ground state. Let us assume for the moment the classical picture of spins in Fig. 6.4 as localized magnetic dipoles. They are coupled from site to site by the scalar product of their respective spin vectors, weighted by the exchange integral. If one of the spins is tilted against the preferential direction thus raising the energy, the neighboring spins tend to follow this tilt. If this spin is released from its tilted orientation the whole system will start to perform a collective motion just as the linear chain of masses connected by springs in Chap. 3: The masses now become the magnetic moments of the spins and the role of the springs is taken by the exchange coupling. This collective excitation of the localized interacting spins are the spin waves or, in quantized form, the magnons. In order to quantify this consideration, we start from the Heisenberg model in tight-binding approximation, for which the exchange coupling is taken into account only between nearest neighbors (n.n.i, j), thus the Hamiltonian (6.23) simplifies to Hspin = −J Siz . (6.27) S i · S j − gμB Hext n.n.i,j

i

and contains only one exchange integral. The nature of the spin, being an angular momentum, can be exploited by making use of the corresponding representation with Si± |SM i = {S(S + 1) − M (M ± 1)}1/2 |SM ± 1i ,

(6.28)

where |SM i is an eigenstate of the spin operators S 2i and Siz for the site i. It is convenient here to write the Hamiltonian with raising and lowering operators Si± (see (6.18)) 1 (Si+ Sj− + Si− Sj+ ) + Siz Sjz − gμB Hext Siz . (6.29) Hspin = −J 2 i n.n.i,j

6.3 Spin Waves in Ferromagnets

171

Let us first calculate the expectation value of Hspin without external magnetic field (Hext = 0) in the ferromagnetic ground state, formulated as |Ψ0 = 7 i |SSi : E0 = Ψ0 |Hspin |Ψ0 = −J SS|Siz |SSi SS|Sjz |SSj = −JνS 2 N,

(6.30)

n.n.i,j

where ν is the number of nearest neighbors and N the number of sites in the chain. Note, that the ferromagnetic ground state in angular momentum representation is composed of angular momentum eigenstates with maximum z component. Application of the raising operator to |SS gives zero, while application of the z components of the spin operators to the ground state yields * |SSi = N S|Ψ0 . (6.31) Sjz Sjz |Ψ0 = j

j

i

As we see from (6.30), the ground state energy is negative if the exchange integral J, whose sign depends on the configuration of nearest neighbors and the atomic orbital, is positive. The stability of the ground state increases with J, with the number of nearest neighbors ν, and with the total spin S. The latter changes with the number of electrons in an incomplete shell and takes a maximum value for half-filling of the shell according to Hund’s rule3 . Before studying the collective excitations of the spin lattice it is quite instructive to have a look at the dynamics of the individual spin operator S j . It is straightforward to derive the equation of motion with the Heisenberg Hamiltonian (6.23) (Problem 6.5) dS j i 1 = [Hspin , S j ] = − (H j × S j ) , dt ¯ h h ¯

(6.32)

where H j = i Jij S i + gμB H ext is an eﬀective magnetic field acting on the spin S j at the same site. The equation of motion (6.32) is that of an angular momentum, whose dynamics is determined by a torque, or of a magnetic dipole moving in a magnetic field. This ubiquitous equation of motion, originally formulated in the context of magnetic resonance phenomena, is known as Bloch equation [184]. In order to solve (6.32) we choose as before H ext = (0, 0, Hext), which gives for the transverse spin components the equations of motion h ¯ h ¯

3

dSjx =− Jij Siy Sjz − Siz Sjy + gμB Hext Sjy dt n.n.i

dSjy =− Jij Siz Sjx − Six Sjz − gμB Hext Sjx . dt n.n.i

Friedrich Hund 1896–1997.

(6.33)

172

6 Spin Waves: Magnons

Close to the ferromagnetic ground state, i.e. at low temperature, we may use the replacement Sjz → Sjz ≃ S (while Sjx , Sjy ≪ S) to simplify these equations: h ¯

dSjx = −S Jij Siy − Sjy + gμB Hext Sjy dt n.n.i

dSjy = −S h ¯ Jij Sjx − Six − gμB Hext Sjx . dt n.n.i They are combined with Sj± = Sjx ± iSjy in the equation dSj± ± ± ± = ∓i S Jij (Sj − Si ) + gμB Hext Sj . h ¯ dt n.n.i

(6.34)

(6.35)

The coupled motion of neighboring spins can be decoupled by exploiting the periodicity of the chain (or solid) with the Bloch representation 1 −ik·Rj ± Sk± = √ Sj . e N j

(6.36)

With these normal coordinates (see Chap. 3, where the same concept was applied in lattice dynamics) we arrive at ⎛ ⎞ ± h dSk ¯ = ⎝S (6.37) Jij 1 − e−ik·(Ri −Rj ) + gμB Hext ⎠ Sj± . i dt n.n.i,j

The exchange integrals Jij to the ν nearest neighbors with d = Ri − Rj are all the same in a cubic lattice and we may write instead the expression in brackets as 1 ik·d hωk = 2JνS(1 − γk ) + gμB Hext , with γk = ¯ e . (6.38) ν d

Thus, we come to the important result, that the whole spin configuration of the lattice (but also the individual localized spin) performs an oscillatory motion with the frequency ωk , if small deviations from the ferromagnetic ground state are considered. This collective mode, the spin wave characterized by the wave vector k (depicted in Fig. 6.5), corresponds to a coherent precession of the

λ Fig. 6.5. Schematic of a ferromagnetic spin wave on a linear chain

6.3 Spin Waves in Ferromagnets

173

individual spins around the direction of the ferromagnetic orientation. It is completely analogous to the lattice modes in Chap. 3. Its quantized form is called magnon as it is related to magnetic properties of the solid. For methodical reasons, it is worth presenting as an alternative to the Bloch equation, which essentially is a classical concept, a formulation based on the occupation number representation and adapted to the quantized form of elementary excitations. In the previous formulation of excitations out of the ferromagnetic ground state, the emphasis was on the transverse components of the local spins to quantify the precession of the spin vectors around the z direction. The alternative is to quantify the deviation of Sjz from its maximum value S. This is achieved with a transformation that replaces the ladder operators Sj± by boson operators a†j , aj with [ai , a†j ] = δij according to 1/2 1/2 √ √ 1 † 1 † † + − a aj a aj aj , Sj = 2Saj 1 − (6.39) Sj = 2S 1 − 2S j 2S j without changing the commutation relations among the three angular momentum operators Sj+ , Sj− , Sjz . It is the Holstein4 –Primakoﬀ 5 transformation [185]. The z component of the spin operator follows (with f (a†j aj ) = {1 − a†j aj / 2S}1/2 ) from [Sj+ , Sj− ] = 2S f (a†j aj )aj a†j f (a†j aj ) − a†j f 2 (a†j aj )aj = 2(S − a†j aj )

such that

Sjz = S − a†j aj

(6.40)

(6.41)

describes the deviation from the maximum value S by the number operator a†j aj which will turn out to be that of quantized excitations out of the ferromagnetic ground state. Let us assume again excitations of low energy which can be quantified by a†j aj ≪ S and allow us to expand the operator function f (a†j aj ). Then ' √ √ √ & 1 † aj aj + . . . aj ≃ 2Saj , and Sj+ ≃ 2Sa†j , (6.42) Sj+ = 2S 1 − 4S

and we have a one-to-one correspondence between the ladder operators Sj± and the boson operators aj and a†j . The factor f (a†j aj ) is required here only to obtain (6.41). Under the condition of low-energy excitations the Heisenberg Hamiltonian can now be written & † ' S ai aj + a†i aj + S 2 − S ai a†i + a†j aj Hspin ≃ −J n.n.i,j

≃ −E0 + JS 4 5

' & † ai ai + a†j aj − ai a†j − a†i aj

n.n.i,j

Theodor D. Holstein, 1915–1985. Henry Primakoﬀ, 1914–1983.

(6.43)

174

6 Spin Waves: Magnons

where higher order terms in the occupation number operator (corresponding to magnon–magnon coupling analogous to lattice dynamics beyond the harmonic approximation) have been omitted and E0 is the ground state energy (see (6.30)). The last two terms represent a coupling between nearest neighbor lattice sites which can be removed by transforming to normal coordinates (or switching from the Wannier to the Bloch representation) with 1 −ik·Rj aj = √ bk , e N k

1 ik·Rj † a†j = √ bk . e N k

(6.44)

The individual contributions to (6.43) can be expressed in the new boson operators for the collective excitation as

a†j aj = ν

a†i aj

n.n.ij

† 1 ′ bk bk e−i(k−k )·Rj b†k′ bk = ν N j k kk′ δkk′

n.n.ij

1 ′ eik·Ri e−ik ·Rj b†k bk′ = N n.n.ij kk′ 1 ′ ei(k−k )·Rj eik·d b†k bk′ = ν γk b†k bk (6.45) = N ′ j d

kk

k

and the Hamiltonian takes the approximate form Hspin ≃ E0 + 2JνS (1 − γk )b†k bk = E0 + ¯hωk b†k bk . k

(6.46)

k

It is the Hamiltonian for spin waves in ferromagnets. The meaning of the boson operators b†k and bk can be read from the relation † Sjz = N S − bk bk , (6.47) j

k

where the first term N S gives the maximum value of the z components of all spin operators in the ferromagnetic ground state and the second term counts the number of quantized collective excitations, the magnons. Because we have assumed exchange interaction only between nearest neighbors, the magnon dispersion ωk , depending on the configuration of the magnetic ions, can be expressed in terms of cosine functions as is typical for the tight-binding approximation (see Sect. 5.4). The width of the magnon band, determined by the number of nearest neighbors and the exchange constant, is in the range of a few to some tens meV, which is about the same as for phonons. For |k| ≪ 2π/a we may expand 1 − γk ≃ 1 − (1 −

1 |k · d|2 ) ∼ k 2 2ν d

(6.48)

6.4 Spin Waves in Anti-Ferromagnets

175

and find a quadratic dependence on the wave vector around the minimum at k = 0. This quadratic dispersion has been measured by inelastic neutron scattering, e.g. in the 3d transition metals [186, 187], but will be shown only when we discuss itinerant magnetism in Sect. 6.6. For a more recent introduction into the concepts of neutron scattering in solids with examples of magnetic excitations see [96, 97]. Similar to phonons, magnons also can be thermally excited and contribute to the specific heat. This contribution can be calculated from the thermal expectation value of (6.46) hωk b†k bk ¯ E(T ) = E0 + k

= E0 +

V (2π)3

hω k ¯

1 d3 k. eβ¯hωk − 1

(6.49)

As we are interested in the low-temperature behavior of the specific heat, which is determined by the quantum nature of the excitations, the dispersion can be approximated as h ¯ ωk = αk 2 , where for a simple cubic lattice with lattice constant a one has α = 2JSa2 . With this isotropic quadratic dispersion the integral can be evaluated in polar coordinates and reads after substituting x = βαk 2 !5/2 xmax 1 kB T α dx. (6.50) E(T ) = E0 + 2 x3/2 x 2π α e −1 0 We may compare here with E(T ) of (3.64), obtained for the acoustic phonons in the Debye model. For low-temperatures the upper limit of the integral can be shifted to infinity and we obtain one of the Bose integrals (see Appendix), which can be expressed here as Γ(5/2)ζ(5/2; 1). Finally, we find for the magnon contribution to the thermal energy E(T ) = E0 +

0.45 5/2 (kB T ) π 2 α3/2

(6.51)

and for the specific heat cV (T ) =

dE(T ) "" = 0.113 kB " dT V =const.

kB T α

!3/2

.

(6.52)

This T 3/2 dependence is characteristic of the magnon contribution to the specific heat and makes it distinct from the contributions of the acoustic phonons (∼T 3 ) or free electrons (∼T ).

6.4 Spin Waves in Anti-Ferromagnets In the last section, we have assumed a ferromagnetic ground state, which required a positive exchange integral between nearest neighbors. An antiferromagnetic order, with the spins on neighboring lattice sites being oriented

176

6 Spin Waves: Magnons

Fig. 6.6. Linear chain with anti-ferromagnetic order, the Wigner–Seitz cell contains two ions with opposite spin

× ×

Mn lattice 1 ×

Mn lattice 2 × × F ×

×

Fig. 6.7. Structure of the anti-ferromagnetic insulator MnF2

anti-parallel to each other, would mean an enlarged unit cell with a basis consisting of (at least) two atoms with a negative exchange integral between nearest neighbors, as depicted in Fig. 6.6. As a real three-dimensional system we show in Fig. 6.7, the crystal structure of the anti-ferromagnetic insulator MnF2 . The configuration of the fluorine ions around the Mn ions on the corners of the cube is diﬀerent from that around the Mn ion in the center, thus the structure is that of two interpenetrating simple cubic lattices each with a basis of one Mn and two F ions. The Hamiltonian for this system (without external magnetic field) can be written Hspin = Ja S 1i · S 2j n.n.i,j

= Ja

n.n.ij

! 1 + − − + z z S S + S1i S2j + S1i S2j . 2 1i 2j

(6.53)

Here S 1i , S 2j are the spin vector operators on the two sublattices and the negative sign of the exchange integral is absorbed in the positive constant Ja . In order to derive a spin wave Hamiltonian similar to (6.46) we employ the Holstein–Primakoﬀ transformation, introduced in the previous section, to replace in a first step, the spin operators on each sublattice by boson operators √ √ + − z S1i ≃ 2Sa1i , S1i = +S − a†1i a1i (sublattice 1) ≃ 2Sa†1i , S1i √ √ − + z S2i ≃ 2Sa†2i , S2i = −S + a†2i a2i (sublattice 2) ≃ 2Sa2i , S2i (6.54)

and find the approximate Heisenberg Hamiltonian for low-energy excitations out of the anti-ferromagnetic state, i.e. for a†1i a1i , a†2i a2i ≪ S,

6.4 Spin Waves in Anti-Ferromagnets

Hspin ≃ Ja

(

n.n.ij

) − S 2 + S a†1i a1i + a†2i a2i + a1i a2j + a†1i a†2j .

177

(6.55)

The first term ∼ −S 2 on the right hand side gives the energy of the antiferromagnetic configuration Ea = −2Ja νN S 2 . It will be shown later that this configuration does not represent the ground state. The remaining terms can be formulated again in the Bloch representation for each sublattice and we obtain & † ' Hspin ≃ Ea + 2Ja νS b1k b1k + b†2k b2k + γk b†1k b†2k + b1k b2k . (6.56) k

The first terms under the sum, count the elementary excitations on each sublattice, while the last term describes the coupling between the sublattices which still has to be removed. This is achieved by the Bogoliubov6 transformation αk = uk b1k − vk b†2k ,

βk = uk b2k − vk b†1k ,

α†k = uk b†1k − vk b2k

βk† = uk b†2k − vk b1k ,

(6.57)

where the coeﬃcients uk and vk are real. Also the new operators obey the boson commutation rules (Problem 6.6) [αk , α†k′ ] = [βk , βk† ′ ] = δkk′ , [αk , βk′ ] = [α†k , βk† ′ ] = [αk , βk† ′ ] = 0.

(6.58)

which leads to the constraint u2k − vk2 = 1. Note, that the boson operators for diﬀerent sublattices commute with each other. The Bogoliubov transformation is an important concept to exactly eliminate a bilinear coupling between two boson systems, of those represented by the spin wave operators on the two sublattices. It can also be applied to phonon–photon or exciton–photon coupling (leading to polaritons, see Chap. 10) or to plasmon–phonon coupling (to yield the coupled plasmon– phonon modes (see Sect. 4.6)). A further example is the magnon–phonon coupling to be treated in Problem 6.7. An analogous transformation for bilinear fermion coupling, the Bogoliubov–Valatin transformation is used in the theory of superconductivity [19]. This transformation is applied here in its inverted form according to which b1k = uk αk + vk βk† , b2k = uk βk + vk α†k

(6.59)

and corresponding expressions for Hermitian adjoint operators, to find (Problem 6.8) & Hspin = Ea + 2Ja νS − 1 + u2k + vk2 + 2uk vk α†k αk + βk† βk + 1 k

6

' + 2uk vk + γk (u2k + vk2 ) α†k βk† + αk βk .

Nikolai Nikolaevich Bogoliubov 1909–1992.

(6.60)

178

6 Spin Waves: Magnons

The last term on the rhs , which is bilinear in the α and β operators, vanishes for 2uk vk + γk (u2k + vk2 ) = 0, (6.61) which together with u2k − vk2 = 1 leads to u2k + vk2 + 2γk uk vk = and u2k

1 = 2

1

vk2

# +1 , 1 − γk2

8 1 − γk2

1 = 2

(6.62)

1

# −1 . 1 − γk2

(6.63)

Thus, we arrive at the Hamiltonian for anti-ferromagnetic spin waves Hspin = −2Ja νN S(S + 1) + (6.64) hωk α†k αk + βk† βk + 1 . ¯ k

This Hamiltonian diﬀers from that of the ferromagnetic magnons (6.46) by containing number operators of two elementary excitations (α†k αk and βk† βk ) with the same energy and in addition, a zero point contribution. The magnon energy is h ¯ ωk = 2Ja νS(1 − γk2 )1/2 with γk as defined before. In contrast with the dispersion of the ferromagnetic magnons, we have now for k ≪ π/a with 1 − γk2 ∼ k 2 a linear dependence of the magnon frequency with k. An example for this linear dispersion measured by inelastic neutron scattering in MnF2 is shown in Fig. 6.8. Note that the dispersion does not exactly follow the linear dependence ∼ k, but exhibits a small gap at k = 0. It is due to an anisotropy field, which removes the degeneracy of the two diﬀerent magnons and will be the subject of Problem 6.9.

100

hωk [°K]

<001> direction <100> direction

50

0 0

0.5 k/kmax

1.0

Fig. 6.8. Magnon dispersion in MnF2 measured by inelastic neutron scattering for two diﬀerent directions in k space after [188]. Note the small gap at k = 0 due to a small anisotropy field (see Problem 6.9)

6.5 Molecular Field Approximation

179

The energy of the ground state (the magnon vacuum, here denoted by |Ψ0 ) ! 8 2 2 E0 = Ψ0 |Hspin |Ψ0 = −2Ja νN S − 2Ja νS (6.65) 1 − 1 − γk k

is composed of the energy of the perfect anti-ferromagnetic configuration (first #term) and the zero-point contribution, which is negative because of 1 − 1 − γk2 > 0 (second term). This indicates that the ground state deviates from the exact anti-ferromagnetic order as can be seen more clearly by looking at the departure from the maximum value of the z component of the spin, N S − Ψ0 | j Sjz |Ψ0 , in the sublattices (which would vanish for the perfect spin alignment). We find for the sublattice 1 (and similar for the sublattice 2) with † Sjz = N S − b1k b1k j

k

= NS − = NS −

k

k

uk α†k + vk βk

uk αk + vk βk†

u2k α†k αk + vk2 βk βk† + uk vk α†k βk† + αk βk (6.66)

and by making use of the commutation rules for the β operators after taking the expectation value with |Ψ0 N S − Ψ0 |

j

Sjz |Ψ0 =

k

vk2 =

1 (1 − γk2 )−1/2 − 1 = 0. 2

(6.67)

k

Thus, in the ground state, the magnon vacuum, the spins in the individual sublattices are not perfectly aligned but slightly disordered.

6.5 Molecular Field Approximation The theory of spin waves outlined in Sects. 6.3 and 6.4 has provided some insight into the low-energy excitations, which means a departure from the state with ferromagnetic or anti-ferromagnetic order. It is known, that magnetic order exists only below a critical temperature at which a phase transition takes place. For ferromagnets, this critical temperature is the Curie 7 temperature TC at which a transition to the paramagnetic phase takes place. The transition temperature for anti-ferromagnets is the N´eel 8 temperature TN . 7

8

Pierre Curie 1859–1906, shared the Noble prize in physics 1903 with his wife Marie Curie. Louis N´eel 1904–2000, Noble prize in physics 1970.

180

6 Spin Waves: Magnons

The quantity to be studied here is the magnetization M (T ) and its temperature dependence up to and beyond TC . Assuming its orientation (which can be defined by a weak external magnetic field) in z direction we may write M (T ) = (0, 0, M (T )) with z (6.68) Sj , M (T ) = gμB j

i.e., the magnetization is determined by the thermal expectation value of the z components of the localized spin operators. For temperatures below the Curie temperature TC , it describes the spontaneous magnetization which in terms of ferromagnetic spin waves reads † M (T ) = gμB N S − (6.69) bk bk . k

The first term M (0) = gμB N S is the saturation magnetization and the second term accounts for the departure from M (0) by thermal excitation of magnons which follows a temperature dependence given by T 3/2 (Problem 6.10). When approaching TC the magnetization vanishes, which in the language of magnons would require b†k bk to approach N S in contrast with the condition b†k bk ≪ N S for the spin–wave theory, or else, this theory is valid only for T ≪ TC . Let us look, therefore, again at the Heisenberg Hamiltonian with an external magnetic field (to fix the orientation of the spontaneous magnetization) written in the form ' & J S i + gμB H ext · S j . (6.70) Hspin = − j

n.n.i=j

It suggests to interpret the content of the curly bracket as an eﬀective magnetic field acting on the spin S j . This interpretation requires that the spin operator S i is replaced by its thermal expectation value S i . To be more explicit, by splitting the spin operator S i = S i + δS i into its thermal expectation (or mean) value and deviations from it, called spin fluctuations, the spin–spin interaction can be rewritten S i · S j = S i · S j + S i · δS j + δS i · S j + δS i · δS j .

(6.71)

Neglecting the last term quadratic in the fluctuations, which are assumed to be small, the Heisenberg Hamiltonian can be cast into the form ' & Hspin = − 2J S i + gμB H ext · S j + JS i · S j , j

n.n.i=j

n.n.i=j

(6.72)

6.5 Molecular Field Approximation

181

where the last term is a constant. The first term contains besides the Zeeman term the molecular field, defined by 2J S i = gμB H M (6.73) n.n.i=j

and the spin Hamiltonian becomes {H M + H ext } · S j + J S i · S j . Hspin ≃ −gμB j

(6.74)

n.n.i=j

The molecular field H M (originally introduced by P. Weiss9 ) accounts for the interaction of the spin S j with all the other spins S i , i = j replaced by S i , but can be quantified, only when the thermal expectation values of all theses spins are known. Note that this seemingly simple concept is quite general as it can be applied to any system of interacting particles as e.g. in the Hartree– Fock approximation introduced in Sect. 4.4. In this more general context, the molecular field is also called mean field and the approximation denoted mean field approximation. According to the translational symmetry of the system, the individual localized spins contribute equally to the magnetization. Therefore, it can be written N M (T ) = gμB S j = gμB N S j (6.75) j

and the molecular field can be expressed as HM =

2J νJ S i = λM , λ = 2 2 . N gμB g μB

(6.76)

n.n.i=j

It is determined by the magnetization M (T ) of the system and a constant ∼νJ, the Weiss constant, which is stronger the larger the exchange integral J and the number ν of nearest neighbors in the lattice. Now, we focus on the temperature dependence of the magnetization. Let the external magnetic field and the magnetization point in the z direction to define an eﬀective field Heﬀ = Hext + λM . Then the spin Hamiltonian takes the form (up to a constant) Hspin = −gμB Siz Heﬀ (6.77) i

and the magnetization is to be calculated as thermal expectation value 1 e−βHspin Siz with M (T ) = gμB Tr Z 9

Pierre Ernest Weiss 1865–1940.

β = 1/kB T.

(6.78)

182

6 Spin Waves: Magnons

The trace can be evaluated with the eigenstates of (6.77) which are * |SMS i , MS = −S, −S + 1, . . . S − 1, S.

(6.79)

i

The eigenvalues of Siz are independent of the site

Siz |SMS i = MS |SMS i

(6.80)

and we can write M (T ) =

1 Z i

+S

2μB MS egµB βMS Heff

(6.81)

MS =−S

which gives M (T ) = gμB SBS (gμB βSHeﬀ ),

(6.82)

with the Brillouin function (with y = gμB βSHeﬀ ) ! 1 y 2S + 1 2S + 1 coth y − coth . BS (y) = 2S 2S 2S 2S

(6.83)

In the limit of low temperatures, y → ∞, the Brillouin function BS (y) → 1 and we find T →0:

M (T ) → M (0) = gμB S

(6.84)

i.e., the magnetization correctly approaches the saturation value. Actually, according to (6.82) M is a function of the variable x = Heﬀ /T in which it shows the saturation behavior depicted in Fig. 6.9. Because of its definition, the eﬀective field is itself a function of the magnetization. Thus, (6.82) represents an implicit equation for M (T ). It can be solved by considering besides (6.82) the second expression of M (x) obtained from the eﬀective field for the case of vanishing external field

M(T) M(0) 1

T=Tc

T/Tc <1

x Fig. 6.9. Saturation behavior of the spontaneous magnetization and graphical solution for M (T ) with the phase transition for a ferromagnet (after [12])

6.5 Molecular Field Approximation

183

T x. (6.85) λ For suﬃciently low temperature, the graph of (6.85), which is a straight line (see Fig. 6.9), cuts the saturation curve (6.82) always at a finite value of M (T ). With increasing temperature, this crossing point moves to the left until at the critical temperature, the graph of (6.85) becomes the tangent of (6.82) at M (T ) = 0. This signifies the transition from the ferromagnetic to the paramagnetic phase. The critical temperature can be obtained from the derivative of the magnetization with respect to x, which is found from the high temperature expansion (y ≪ 1) of the Brillouin function M (x) =

BS (y) =

S + 1 y (2S + 1)4 − 1 y 3 − + ... S 3 (2S)4 45

(6.86)

according to which the leading terms of the magnetization in the absence of the external magnetic field (Hext → 0) are M (T ) M (T )3 , −A T T3 where we have identified the Curie temperature with M (T ) = TC

TC =

JνS(S + 1) 3kB

(6.87)

(6.88)

and A is another constant determined by the system parameters. After dividing by M (T ) (6.87) becomes a quadratic equation which gives the qualitative relation M (T → TC ) ∼ (TC − T )1/2 ,

T < TC .

(6.89)

Thus, the mean field theory allows us to describe the expected vanishing of the magnetization, when approaching the critical temperature from below, and it gives also the temperature dependence with the critical exponent 1/2. This behavior is typical for a second order phase transition, which is characterized here by the magnetization M (T ) as order parameter. The usual plot of M (T )/M (0) versus T /TC (Fig. 6.10) is universal for a second order phase transition and does not depend on the ferromagnetic material. Thus, the data points for Ni and Fe fall onto the same curve, which is well described by the Brillouin function with S = 1/2. The validity of the mean-field approach has been confirmed also by ab-initio calculations with the dynamical mean-field theory.[189] Above the critical temperature the magnetization in the presence of an external magnetic field is M (T > TC ) = g 2 μ2B N S(S + 1) =

1 (Hext + λM ) 3kB T

TC C Hext + M (T ) T T

(6.90)

184

6 Spin Waves: Magnons

S=1/2

M(T) / M(0)

1.0 0.8

S=1

0.6

S=∞

0.4

Ni Fe

0.2 0

0.2 0.4 0.6 0.8 1.0 T/ Tc

Fig. 6.10. Dependence of the reduced saturation magnetization on the reduced temperature. Symbols are experimental data for Ni and Fe, solid lines are calculated from the Brillouin function with diﬀerent values for the total spin S (after [30])

Table 6.1. Curie temperatures (TC in K) and saturation magnetization (M (0) in Gauss) for some ferromagnets and N´eel temperatures (TN in K) for some antiferromagnets Ferromagnets Fe Co Ni EuO EuS MnAs

TC

M (0)

Anti-Ferromagnets

1043 1388 627 77 16.5 318

1752 1446 510 1910 1184 870

MnO FeO CoO NiO MnF2 CoF2

TN 122 198 291 600 67.34 37.7

and it follows the Curie–Weiss law M (T > TC ) =

C Hext , T − TC

T > TC .

(6.91)

Similar considerations lead to the phase transition between the antiferromagnetic and paramagnetic phases characterized by the N´eel temperature as critical temperature. In Table 6.1 some systems with magnetic order are given together with their critical temperatures and (for the ferromagnetic systems) their saturation magnetization at T = 0 K. As an example of a ferrimagnet we refer to Fe3 O4 (magnetite) with TC = 858 K and M (T = 0) = 510 G. All data are taken from [164]. For further reading about phase transitions and critical exponents we refer to [190–192].

6.6 Itinerant Electron Magnetism

185

6.6 Itinerant Electron Magnetism While in the previous sections we have assumed a dominating exchange interaction and strongly localized electrons, we want to study now the opposite situation of Bloch or itinerant electrons with weak exchange interaction. The appropriate Hamiltonian for this case is [166, 171] H=

kσ

1 † c c† ′ ǫk c†kσ ckσ + U ′ ck′ σ ′ ckσ . 2 kk′ q k+qσ k −qσ

(6.92)

σσ′

It describes electrons in a band ǫk with an interaction independent of k, which 2 can be understood as a screened Coulomb interaction e2 /V ε0 (k 2 + kFT ) with neglect of the dependence on k, i.e. it corresponds to the replacements vk =

e2 e2 e2 → → = U. 2 ) 2 ε0 V k 2 ε0 V (k 2 + kFT ε0 V kFT

(6.93)

We apply the Hartree–Fock approximation by considering only contributions of the interaction with k′ = k + q and σ = σ ′ , which allow us to write the interaction as † 1 1 Hint = − U ck1 σ ck1 σ c†k2 σ ck2 σ = − U n k1 σ n k 2 σ . (6.94) 2 2 k1 k2 σ

k1 k2 σ

This can be formulated also with the numbers Nσ = and spin-down (σ = −) electrons as Hint = −

U 2

k

nkσ of spin-up (σ = +)

! 1 1 (N+ − N− )2 + N 2 , 2 2

(6.95)

where N = N+ + N− is the total number of electrons. The HF Hamiltonian, known also as the Stoner10 model then reads ! U ∆ † (N+ − N− )2 − N 2 . ckσ ckσ + HHF = ǫk − σ (6.96) 2 4 k,σ

The energy spectrum εkσ = εk − σ∆/2 is depicted in Fig. 6.11, where for simplicity, a parabolic dispersion as for free electrons is assumed. The up and down spin electrons diﬀer in their energies by ∆ = U (N+ − N− ). The number diﬀerence between both kinds of electrons determines the magnetization thus indicating the relation with magnetism. We recognize that the exchange interaction leads to a similar result as the Zeeman term in the discussion of the Pauli spin paramagnetism in Sect. 4.2. However, the k independent shift of the spin-up and spin-down band applies to any dispersion relation ǫk as e.g. in Fig. 6.1. 10

Edmund Clifton Stoner 1889–1973

186

6 Spin Waves: Magnons

εkσ

–

+ EF

∆ –

+

kF kF Fig. 6.11. Schematic dispersion for spin-split energy bands

The Stoner model is the starting point for investigating the dependence of the total energy on the degree of spin polarization ζ = (N+ − N− )/N . This will lead us to a refinement of the estimate given at the beginning of this Chapter with respect to the existence of ferromagnetic order in a system of itinerant electrons. The ground state energy in HF approximation follows from (6.96) as ! U U HF fk+ ǫk − (N+ − N− ) + fk− ǫk + (N+ − N− ) E0 = 2 2 k

+

U (N+ − N− )2 − n2 4

(6.97)

with the Fermi distribution function fk± . The sums over the band energies are carried out by assuming the free electron dispersion ǫk = h ¯ 2 k 2 /2m and T = 0 to give 3 fk± ǫk = N± EF (N± ). (6.98) 5 k

Using the corresponding results from Sect. 4.4 but with EF (N± ) = 22/3 EF (N ), because of single occupancy of the states in k space with spin aligned electrons, the total energy in HF approximation is E0HF (N+ , N− ) =

3 N+ EF (N+ ) + N− EF (N− ) 5 U (N+ − N− )2 + N 2 . − 4

(6.99)

We are interested in the dependence on the degree of spin polarization ζ = (N+ − N− )/N , and replace N± according to N± =

1 N (1 ± ζ), 2

which gives for the mean energy per particle

(6.100)

6.6 Itinerant Electron Magnetism

187

E0HF (N+ ) + E0HF (N− ) N UN 2 3 EF (1 + ζ)5/3 + (1 − ζ)5/3 − ζ − 1 . (6.101) = 5 2 4

ǫHF (N, ζ) =

A minimum of ǫHF (N, ζ) at finite ζ would indicate the existence of a stable ferromagnetic state. This condition leads to the relation UN ζ = (1 + ζ)2/3 − (1 − ζ)2/3 , EF

0 ≤ ζ ≤ 1.

(6.102)

The rhs is a monotonous function, starting for ζ = 0 at zero with a slope of 4/3 and reaching for ζ = 1 the value 22/3 with infinite slope. With respect to (6.102) three situations are possible: (a) U N/EF < 4/3, the relation has no solution for finite ζ, which leads to a stable paramagnetic state, (b) 4/3 < U N/EF < 22/3 , there exists a solution for 0 < ζ < 1 representing a stable ferromagnetic state, with partial spin polarization, (c) 22/3 < U N/EF , there is always a solution with ζ = 1 representing a ferromagnet with perfect alignment of all spins. The ratio N/EF is proportional to the density of states at the Fermi energy D(EF ), thus the existence of ferromagnetism in a system of itinerant electrons is ruled by the competition between D(EF ) (which is determined by the dispersion) and the strength of the exchange energy U . If, for the given U , the width of the energy band increases, i.e. D(EF ) decreases, the criterion (c) or even (b) will be missed and ferromagnetism will not be realized. Thus, the Stoner model provides with the criterion (a), the Stoner condition, a clear answer with respect to the existence of ferromagnetism for itinerant electrons. Besides the existence of magnetic order the other basic property of a spin system is the excitation spectrum out of the ground state. In Chap. 4 it was the dielectric function which has led us to the excitation spectrum of the free interacting electrons consisting of single-particle (or electron–hole) and collective excitations (the plasmons). For the latter, it was necessary to employ the random phase approximation. We remember that the inverse dielectric function is a density–density correlation function. Here, we want to put the emphasis on spin-flip excitations. The corresponding response function is the magnetic or spin susceptibility, which is a correlation function between components of the magnetic dipole density. The observable of interest here is the magnetization related to spins S l of electrons at r l M (r) =

m(r) = gμB S l δ(r − rl ) V

(6.103)

l

as introduced already in Sect. 2.5. This quantity can be addressed by a magnetic field with the interaction term Vext (t) = −m · B(t). Spin flips are caused

188

6 Spin Waves: Magnons

by the ladder operators S ± , which for a spin-1/2 system are expressed in terms auf Pauli spin matrices as ! ! 00 01 . (6.104) , S− = (σx − iσy )/2 = S+ = (σx + iσy )/2 = 10 00 In the following we assume Bz = 0 and choose the appropriate decomposition of the external perturbation Vext (t) = −(m+ B− e−iωt + m− B+ eiωt )

(6.105)

which describes the interaction with a rotating magnetic field in the xy plane. The observable we are interested in here is M+ = m+ /V . It is related to the susceptibility responsible for spin-flip processes i +∞ iωτ χM e Θ(τ )[M+ (q, τ ), M− (−q, 0)]0 dτ. (6.106) (q, ω) = +− h −∞ ¯ This correlation function, which has the same structure as (2.61), can be evaluated in the occupation number representation. In terms of fermion operators for the Bloch electrons we have † † M+ (q) = gμB ck+q↑ ck↓ , M− (−q) = gμB ck−q↓ ck↑ , (6.107) k

k

which are spin-density fluctuations (multiplied by gμB ). We evaluate the expectation value under the integral for T = 0 with the eigenstates of the HF Hamiltonian (6.96) analogous to the calculation of the inverse dielectric function in Sect. 4.5 but consider the spin splitting of the HF single-particle energies ǫk± = ǫk ∓∆/2 and obtain as the HF result for the spin susceptibility (Problem 6.11) χ0+− (q, ω) = lim g 2 μ2B δ→0

k

fk↓ − fk+q↑ . hω + ǫk+q↑ − ǫk↓ + iδ ¯

(6.108)

It resembles the polarization function π0 (q, ω) in the HF result for the inverse dielectric function (see (4.122)). Its poles (in the lower half of the complex energy plane) mark the particle–hole excitations with spin flip, the Stoner continuum depicted in Fig. 6.12. For q = 0, excitations are possible only for hω = ∆, because the two bands are shifted against each other by the exchange ¯ energy (see Fig. 6.11). With increasing q, spin-flip excitations become possible for a continuum with increasing width. In Fig. 6.12, two situations are shown. If ∆ is smaller than EF the continuum reaches down to vanishing excitation energies, for which spin-flip excitations are possible in the interval kF+ − kF− < q < kF+ +kF− as can be checked also with Fig. 6.11. This situation is called weak ferromagnetic case. In the strong ferromagnetic case, EF < ∆, single-particle excitations with spin flip are possible only for finite excitation energy.

6.6 Itinerant Electron Magnetism

hωq

hωq

(weak)

∆

Stoner continuum

∆

+

–

kF –kF

+

–

kF +kF

189

(strong) Stoner continuum

q

q

Fig. 6.12. Spectra of single-particle (Stoner continuum) and collective excitations with spin-flip for a weak (left) and a strong (right) ferromagnet

These results can be compared with those of the dielectric function for the free electron system in Sect. 4.5. There we have obtained the single-particle or electron–hole excitations (without spin flip) in the HF approximation. As we know from the discussion of the dielectric function, this approximation does not yield the collective charge-density excitations (or plasmons) for which we had to go beyond HF to the RPA. It consisted in replacing the free polarization function π0 (q, ω) by π0 (q, ω)/(1−vq π0 (q, ω)). The corresponding replacement is possible here by identifying π0 (q, ω) with χ ˜+− (q, ω) = χ0+− (q, ω)/g 2 μ2B and vq with the exchange interaction U . Thus the RPA result for the spin susceptibility is easily obtained as 2 2 χM +− (q, ω) = g μB

χ ˜0+− (q, ω) . 1 − Uχ ˜0+− (q, ω)

(6.109)

Besides the poles of single-particle excitations of χ0+− (q, ω), the RPA susceptibility has an additional pole due to the vanishing denominator, 1 − Uχ ˜0+− (q, ω) = 0,

(6.110)

giving the collective excitations of the spin system, the magnons. In order to find the magnon dispersion, we have, in a first step, to evaluate χ ˜0+− (q, ω) under the conditions of collective excitations outside of the Stoner continuum, which we expect to occur for h ¯ ω ≪ ∆ and at small q. Then we have, using the solution of Problem 6.11 !2 5 1 ¯h2 q V kF+ +3 0 χ ˜+− (q, ω) = kF + 6π 2 ¯ hω − ∆ + ǫ q 5m ¯hω − ∆ + ǫq !2 2 +5 k 1 q h ¯ V 3 kF+ + F − 2 6π ¯ hω − ∆ − ǫ q 5m ¯hω − ∆ − ǫq (6.111)

190

6 Spin Waves: Magnons

with ǫq = h ¯ 2 q 2 /2m. We substitute this result in (6.110) and find + 3 3 kF− kF+ UV − 1− 6π 2 ¯ hω − ∆ + ǫq hω − ∆ − ǫ q ¯ , 5 5 kF− kF+ h4 q 2 ¯ = 0. − + 5m2 (¯hω − ∆ + ǫq )3 (¯hω − ∆ − ǫq )3

(6.112)

Being interested in the leading order term in q, we expand the denominators of the first two terms in the bracket with ∆ ≪ ¯hω ± ǫq but neglect the small terms in the dominators of the last terms, which are already proportional to q 2 . Making use of the relation 6π 2 1 +3 3 (6.113) kF − kF− = ∆ UV

the condition for the magnon pole simplifies and can be solved for the magnon energy , + 3 5 5 3 2¯h2 kF+ − kF− h2 q 2 kF+ + kF− ¯ hω q ≃ ¯ . (6.114) − 2m k + 3 − k − 3 5m∆ k + 3 − k − 3 F

With

3 kF±

=

F

F

F

kF3 (1

± ζ) and ∆ = U N ζ this reduces to 2EF (1 + ζ)5/3 − (1 − ζ)5/3 h2 q 2 ¯ 1− . hω q ≃ ¯ 2mζ 5U N ζ

(6.115)

We find the characteristic q 2 dependence giving a dispersion of the ferromagnetic magnons (see Sect. 6.3) separated from the Stoner continuum as indicated in Fig. 6.12. The magnon dispersion of Fe, measured with inelastic neutron scattering, is shown in Fig. 6.13. It follows more or less that this q 2 law is almost independent of the direction of propagation. With increasing q the magnon dispersion becomes degenerate with the Stoner continuum and decays into single-particle excitations with spin-flip. This situation resembles that of the collective charge density excitations (plasmons, see Sect. 4.6) when they become degenerate with the particle-hole continuum of the free electron gas and decay into particle–hole excitations. The curvature of the magnon dispersion depends on the relation between density of states (or Fermi energy over N ) and interaction strength U . For, complete spin polarization ζ = 1, (6.115) simplifies to 2EF 5/3 h2 q 2 ¯ 2 1− (6.116) hωq (ζ = 1) = ¯ 2m 5U N while for small spin polarization ζ ≪ 1 we have 4EF h2 q 2 ¯ hωq (ζ ≪ 1) = ¯ 1− . 2mζ 3U N

(6.117)

Problems

191

Fig. 6.13. Magnon dispersion of ferromagnetic Fe from inelastic neutron scattering after [186]. The dashed line represents the quadratic dispersion

In this limit ωq is positive if UN 4 > . EF 3

(6.118)

Thus, we arrive at the consistent result that collective spin excitations are possible if according to the Stoner condition ferromagnetism and finite spin polarization exist.

Problems 6.1 Derive the expression for the matrix element of the electron–electron interaction Vαββ ′ α′ in the general case! Make use of the Fourier expansion of products of the periodic parts of the Bloch functions. Simplify to the single-band case and use only the leading term of this Fourier expansion. 6.2 Make use of the orthogonality and normalization of the Bloch functions ψnkσ (r) to show that 1 −ik·R e ψnkσ (r) (6.119) φnσ (r − R) = √ N k

192

6.3

6.4

6.5 6.6

6.7

6 Spin Waves: Magnons

are normalized and that they are orthogonal when centered around diﬀerent sites R (Wannier representation)! Derive the commutation relations of cnRσ , c†n′ R′ σ′ from those of the Bloch states! Express the interaction matrix element of Problem 6.1 in the Wannier representation with localized atomic orbitals for the single-band case to obtain (6.10)! Show that the set of operators c†↑ c↓ , c†↓ c↑ , and c†↑ c↑ − c†↓ c↓ fulfills the same commutation relations as the components of the spin-1/2 operators. Make use of the fermion commutation rules for c†σ , cσ ! Derive the equation of motion (6.32) for the spin operator S j determined by the Heisenberg Hamiltonian (6.23)! Given are the commutation rules for the boson operators b†ik , bik , i = 1, 2 for the magnons in the two sublattices of an anti-ferromagnet. Verify the commutation relations (6.58) for the Boson operators αk , βk obtained by the Bogoliubov transformation (6.57)! The Hamiltonian of a solid, in which the simultaneous presence of phonons and magnons shall be studied, can be written as & p † ' Hp−m = h ω k ak ak + h ¯ (6.120) ¯ ωkm b†k bk + ck ak b†k + a†k bk , k

where ωkp (ωkm ) is the phonon(magnon) dispersion and a, a† (b, b† ) are the boson operators for phonons(magnons). The coupling between both elementary excitations is described by the bilinear term with the coupling constant ck . Make use of the following variant of the Bogoliubov transformation to eliminate the coupling (with real Θk ): ak = αk cos Θk + βk sin Θk , bk = βk cos Θk − αk sin Θk

(6.121)

and express Θk (or a function of Θk ) by the coupling constant and the frequencies ωkp (ωkm )! Discuss the solution, especially for ωkp = ωkm , and give the eigenfrequencies of the new collective magnon–phonon modes and how their boson operators are composed of the original phonon and magnon operators! 6.8 Replace in (6.56) the spin wave operators for the sublattices with the help of the Bogoliubov transformation to get the Hamiltonian (6.60)! 6.9 In order to remove the degeneracy of the anti-ferromagnetic magnons, introduce a weak anisotropy field Ha in the Hamiltonian (6.53) acting with opposite sign on the two sublattices, and an external magnetic field. How does this change the dispersion relation for the anti-ferromagnetic magnons? Discuss the frequency in the long wavelength limit k → 0 and sketch the result to show the eﬀect of the anisotropy field! 6.10 Calculate the temperature dependence of M (T ) − M (0) for T ≪ TC in the spin–wave approximation and verify Bloch’s T 3/2 law!

Problems

193

6.11 Derive the expression (6.108) of the spin susceptibility in HF approximation χ0+− (q, ω) at T = 0 ! The calculation is strictly analogous to that performed in Sect. 4.5 to obtain the HF result for the inverse dielectric function. Evaluate the formula by assuming a single-particle dispersion of the form ǫkσ = h ¯ 2 k 2 /2m − σ∆/2 with ∆ = U (N+ − N− )! Make use of the fact that ∆ is the dominating quantity, i.e. ∆ ≫ ¯hω, ¯h2 q 2 /2m!

7 Correlated Electrons

In Sect. 4.7, we have studied in some detail the correlation eﬀects for the homogeneous free electron system and figured out their dependence on the electron density. Correlation has been considered in the eﬀective single-particle potential for crystal electrons within the density-functional theory (DFT), which has led to the independent particle description of the electronic band structure in Chap. 5. In Chap. 6, we have stressed the importance of the exchange interaction and of the electron spin for magnetic properties. The aim of this chapter is to introduce concepts of treating correlation eﬀects for electrons in a crystalline surrounding beyond the independent particle picture in a more general sense. One aspect will be to describe model systems, which allow to demonstrate correlation eﬀects [122, 193–196]. The motivation comes from the observation that some group of solids, namely those with the Fermi energy within narrow bands deriving from d or f electrons, exhibit properties which cannot be understood within the single-particle band structure. Among those are besides the magnetic properties (Chap. 6), the insulating behavior of some transition metal oxides, and the heavy fermion eﬀects in compounds of lanthanides and actinides. Another aspect is the quasi-particle concept in the context of the Fermi liquid theory and the deviations from Fermi liquid behavior in systems with reduced dimension [197–201]. Finally, we would like to address also correlation in two-dimensional electron systems in high magnetic fields, known as the fractional quantum Hall regime [202–204]. As in Chap. 6, we start from the N electron Hamiltonian (see (6.5)) H=

α

ǫα c†α cα +

1 Vαββ ′ α′ c†α c†β cα′ cβ ′ , 2 ′ ′

(7.1)

αβα β

where α, β, α′ , and β ′ are complete sets of single-particle quantum numbers, α = nkσ, and ǫα the corresponding energy values. The matrix elements of the electron–electron interaction e2 ψβ ′ (x)ψα′ (x′ ) Vαββ ′ α′ = dx dx′ ψα† (x)ψβ† (x′ ) (7.2) 4πε0 |x − x′ | U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 7,

196

7 Correlated Electrons

can be expressed in terms of the wave functions ψα (x), where x stands for space and spin coordinates. We shall separate the Hamiltonian (7.1) according to H = Hsp + Hint into the single-particle and the interaction part. In the context of this chapter, the latter accounts for electrons in partially filled bands only, while the interaction of all other electrons is considered as before in the eﬀective single-particle potential leading to the energy values ǫα .

7.1 Retarded Green Function for Electrons In this section, we introduce the concept of the Green function for electrons, which is advantageous in describing electronic properties of the interacting electrons [64, 205, 206]. In general, a retarded Green function is an object defined by (compare with (2.82)) i ′ ′ ˆ ˆ ′ Gret AB (t, t ) = − θ(t − t )[A(t), B(t )]± , h ¯

(7.3)

where . . . denotes the thermal expectation value as defined in Sect. 2.32.3 and ˆ ˆ ′ )]± = A(t) ˆ B(t ˆ ′ ) ± B(t ˆ ′ )A(t). ˆ [A(t), B(t (7.4)

ˆ ˆ ′ ) are fermion(boson) The +(−) sign applies if the operators A(t) and B(t operators. In Sect. 2.6, we have already identified the response function, which is a correlation function for observables, as a special type of retarded Green function. In the context of this chapter, the choice of the operators, ˆ = cα (t) A(t)

ˆ ′ ) = c† (t′ ), and B(t α

(7.5)

will be a creation and an annihilation operator of an electron in a singleparticle state with quantum numbers α. Thus, the Green function (we use as before {. . . , . . . } for the anti-commutator [. . . , . . . ]+ ) i G(α; t, t′ ) = − θ(t − t′ ){cα (t), c†α (t′ )} h ¯

(7.6)

is the probability amplitude to find the particle created at time t′ in the state α at a later time t > t′ still in the same state. It describes the propagation of an electron in the state α from t′ to t and is, therefore, called a propagator. In order to calculate the Green function (7.6), we formulate the equation of motion by taking the derivative with respect to the time argument t. It can be written as i¯ h

∂G(α; t, t′ ) ∂θ(t − t′ ) = {cα (t), c†α (t′ )} ∂t ∂t i ∂cα (t) † ′ , cα (t ) − θ(t − t′ ) i¯ h . h ¯ ∂t

(7.7)

7.1 Retarded Green Function for Electrons

197

With

∂θ(t − t′ ) = δ(t − t′ ) , ∂t and the equation of motion for cα (t) i¯ h

{cα (t), c†α (t)} = 1,

∂cα (t) = [cα (t), H] ∂t

(7.8)

(7.9)

(here [. . . , . . . ] denotes the commutator [. . . , . . . ]− ) this can be cast into the form i¯ h

∂G(α; t, t′ ) i = δ(t − t′ ) − θ(t − t′ ){[cα (t), H], c†α (t′ )}. ∂t h ¯

(7.10)

The second term on the rhs has again the form (7.3) of a retarded Green function, but it has a more complex form. As it will turn out, it contains in general more than two fermion operators due to the electron–electron interaction. One could also formulate, for this higher order Green function, an equation of motion with a similar structure but with an even more complex Green function, and so forth. This is a generic hierarchy problem typical for interacting systems. Concepts of many-particle physics focus on finding for this problem approximate solutions, essentially by truncating the hierarchy at some level. To become acquainted with the concept of Green functions, let us first look for the simplest solution of (7.10) which can be given for the noninteracting electrons, i.e., by considering the Hamiltonian Hsp = (7.11) ǫα′ c†α′ cα′ . α′

Evaluation of the commutator yields i i [cα (t), Hsp ] = e h¯ Hsp t ǫα′ [cα , c†α′ cα′ ]e− h¯ Hsp t = ǫα cα (t).

(7.12)

α′

Here, we have used the fact that the time argument of all operators is the same and determined by Hsp and that the commutation relation for the fermion operators yields [cα , c†α′ cα′ ] = cα′ δα,α′ . The equation of motion for the Green function G(0) (α; t, t′ ) of the noninteracting electron ! ∂ i¯ h − ǫα G(0) (α; t, t′ ) = δ(t − t′ ) (7.13) ∂t can be easily integrated to obtain ′ i i G(0) (α; t, t′ ) = − θ(t − t′ )e− h¯ ǫα (t−t ) . h ¯

(7.14)

After carrying out the Fourier transformation (note the time dependence t − t′ = τ )

198

7 Correlated Electrons

G(0) (α, ω) =

+∞

i

e h¯ (¯hω+iδ)τ G(0) (α, τ )dτ,

(7.15)

−∞

we find the characteristic form of the Green function (with E = h ¯ ω) G(0) (α, E) = lim

δ→0

1 , E − ǫα + iδ

(7.16)

where the small δ has been introduced to yield, in the back transform again, the retarded Green function. It has a pole in the lower complex energy (or frequency) plane close to the real axis at the single-particle energy ǫα . (Compare with the response function for the optical phonons in Sect. 3.5 or with the inverse dielectric function in Sect. 4.5 which exhibit similar structures.) Making use of the relation ! 1 1 lim =P − iπδ(E − ǫα ), (7.17) δ→0 E − ǫα + iδ E − ǫα one finds the density of states 1 D(E) = Tr ImG(0) (α, E) = δ(E − ǫα ). π α

(7.18)

Let us include now the electron–electron interaction for the special case of electrons in a partially filled energy band. For this case, the band index can be dropped, and α is specified by the wave vector k and the spin σ of the electron (see Problem 6.1). The commutator of ckσ with the interaction term Hint =

1 V (k, k′ , q)c†k+qσ c†k′ −qσ′ ck′ σ′ ckσ 2 kk′ q

(7.19)

σσ′

can be calculated as Problem 7.1 and yields V (k′ , k + q, q)c†k′ +qσ′ ck′ σ′ ck+qσ . [ckσ , Hint ] =

(7.20)

k′ qσ′

The equation of motion for the Green function is now ! ∂ i¯ h − ǫkσ G(kσ; t, t′ ) = δ(t − t′ ) + V (k′ , k + q, q)Γ(kσ, k′ σ ′ ; t, t′ ) ∂t k′ q σ

(7.21) where i Γ(kσ, k′ σ ′ ; t, t′ ) = − θ(t − t′ ){c†k′ +qσ′ (t)ck′ σ′ (t)ck+qσ (t), c†kσ (t′ )} h ¯ (7.22)

7.1 Retarded Green Function for Electrons

199

is a higher order Green function. It is convenient to introduce, by means of V (k′ , k + q, q)Γ(kσ, k ′ σ ′ ; t, t′ ) = Σ(kσ; t, t′′ )G(kσ; t′′ , t′ )dt′′ , (7.23) k′ q σ′

the single-particle self-energy Σ(kσ; t, t′′ ). If the Hamiltonian does not depend on t, we can exploit the homogeneity of the time, because the Green functions depend only on the time diﬀerence and perform the Fourier transformation to obtain G(α, E) = lim

δ→0

1 . E − ǫα − Σ(α, E) + iδ

(7.24)

This is the same structure as for the Green function without interaction, but now in addition to the single-particle energy there is a self-energy correction, which in general depends on E and changes the pole structure. Thus, (7.24) indicates the structure of the Green function. However, the problem of including the electron–electron interaction is still to be accomplished by calculating the self-energy and to find the poles of the Green function. For the concepts, how this can be done in a systematic way, we refer to the literature [64]. The Green function of the noninteracting particle (7.16) is related to that of the full Green function (7.24) by the Dyson1 equation G = G(0) + G(0) Σ G,

(7.25)

known from scattering theory. It can be obtained also by taking the Fourier transform of (7.21) together with (7.23) (Problem 7.2). We may look for approximate calculations of the self-energy and do this for T = 0, for which the thermal expectation value reduces to the expectation value in the system ground state (see Problem 2.2). The Green function (7.22) can be approximated with the following replacements c†k′ +qσ′ (t)ck′ σ′ (t)ck+qσ (t) → c†k′ +qσ′ (t)ck′ σ′ (t)ck+qσ (t)

−c†k′ +qσ′ (t)ck+qσ (t)ck′ σ′ (t),

(7.26)

which is a factorization of the three-fermion term into an expectation value and a single-fermion term. The former yields the occupation number nk′ σ′ if q = 0 and nk+qσ provided k′ = k, and σ ′ = σ, while the latter combines with the rest of (7.22) to the Green function G(α; t, t′ ). The corresponding expressions for the self-energy, which do not depend on the time arguments, are ΣH (kσ) = (7.27) V (k, k′ , 0)nk′ σ′ , k′ σ ′

1

Freeman John Dyson *1923.

200

7 Correlated Electrons

k’

<

>

>

< k

k’

k

Fig. 7.1. Graphs of the Hartree and Hartree–Fock self-energies

which is the Hartree self-energy, and V (k, k′ , q)nk′ +qσ , Σ HF (kσ) = −

(7.28)

k′ q

which is the Hartree–Fock self-energy. There is a quite intuitive graphical representation of these self-energies (Fig. 7.1): Both graphs are obtained from the interaction graph (see Fig. 4.10) by connecting incoming and outgoing fermion lines, which is possible in just two distinct ways, leading to the Hartree and Hartree–Fock self-energy. For free electrons, the potential matrix element simplifies according to Vkk′ q → vq , where vq is the Fourier transform of the bare Coulomb potential and we may compare with the results obtained in the jellium model of Chap. 4. In this case, the Hartree self-energy vanishes, because the contribution for q = 0 is exempt from the summation in the interaction term due to compensation with the jellium term and the Hartree–Fock self-energy becomes the one calculated in Sect. 4.4. Let us finally reformulate the denominator of the Green function for interacting particles close to the chemical potential with E ≃ μ. Making use of the expansion ∂Σ "" Σ(α, E) ≃ Σ(α, μ) + (E − μ), (7.29) " ∂E E=µ

the denominator of the retarded Green function (7.24) takes the form ! ∂Σ "" E − ǫα − Σ(α, E) ≃ (E − μ) 1 − − ǫα − μ + Σ(α, μ) (7.30) " ∂E E=µ and the Green function can be written as G(α, E) =

Z(α) . E − μ − ǫ˜α − iγα

(7.31)

Here, we introduce ∂Σ "" " ∂E E=µ ǫα = Z(α) (ǫα − μ + ReΣ(α, μ)) ˜

Z −1 (α) = 1 −

γα ≃ Z(α)ImΣ(α, ǫ˜α ).

(7.32) (7.33) (7.34)

7.2 The Hubbard Model

201

Due to the interaction, the pole of the Green function is now at the energy ǫ˜α + iγα and has the spectral weight Z(α), which equals 1 if the self-energy does not depend on the energy, as is the case for the Hartree and Hartree– Fock approximation. The pole is in the lower (upper) half plane of complex energies depending on the sign of E − μ and corresponds to a quasi-particle (quasi-hole). We shall come back to this result in Sect. 7.3. Here, we note only that the Fourier back transformation leads to a time-dependent propagator of the form (see [122], Sect. 10) ′ i G(α; t − t′ ) = − Z(α)θ(t − t′ )e−(i˜ǫα +γα )(t−t )/¯h . h ¯

(7.35)

The imaginary part of the self-energy indicates a finite lifetime of the quasiparticle or quasi-hole due to particle–particle interaction. A closer inspection shows a decrease of the lifetime with increasing distance from the chemical potential μ (see Sect. 7.3).

7.2 The Hubbard Model After having gained some insight into the Green function concept and the calculation of self-energies, we turn now to a special model, for which the correlation eﬀects can be evaluated in closed form. This model has been designed to explain the observation that transition metal oxides, which according to their electronic configuration should be metals due to partially filled single-particle energy bands, turn out to be insulators. Take as an example CoO, which crystallizes in the rocksalt structure: the configuration of the valence electrons, Co 4s2 3d7 and O 2s2 2p4 , tells us that there is an odd number of electrons for each lattice point (of the fcc lattice), which in the independent particle picture leads to half-filling of the topmost band and means metallic behavior. This topmost band is a narrow d band (see Sect. 5.4), which is even narrower here than for Co (Fig. 5.14), because due to the larger spacing between neighboring Co atoms in CoO the overlap of the d orbitals is smaller. Under this condition the electronic correlation leads to a significant modification of the electronic structure, which – as we shall see – allows one to understand the experimental fact that CoO is not a metal but an insulator [169]. Let us consider electrons in an energy band deriving from one atomic (d or f ) orbital. The corresponding Bloch function can be expressed (see (6.7)) by 1 ik·R e φσ (r − R), ψkσ (r) = √ N R

(7.36)

where we have dropped the band index and kept only the spin index. We may switch with 1 ik·R ckσ = √ e cRσ (7.37) N R

202

7 Correlated Electrons

φ(r)

r R–d

R

R+d

Fig. 7.2. Schematic of atomic orbitals at lattice points of a linear chain and their overlap

to a formulation with fermion operators (see Problem 6.2) and write for the single-particle term as in Sect. 6.1 Hsp = ǫk c†kσ ckσ kσ

=

1 ′ ǫk eik·(R−R ) c†R′ σ cRσ = tRR′ c†R′ σ cRσ . (7.38) N RR′ σ k RR′ σ tRR′

The dispersion of the energy band ǫk is now expressed in terms of the hopping or transfer matrix elements tRR′ , R = R′ , while the term with R = R′ gives the atomic energy level from which the band derives. For small overlap of the atomic orbitals, as is the case here and illustrated in Fig. 7.2, the tight-binding approximation applies and we may write † † ǫk c†kσ ckσ ≃ ǫ0 cRσ cRσ + t cR+dσ cRσ (7.39) Rσ

kσ

Rdσ

with the atomic level energy ǫ0 and the hopping or transfer matrix element between nearest neighbors R − R′ = d (both assumed not to depend on σ) t= φ∗ (r)v(r)φ(r − d)d3 r. (7.40) Likewise, the electron–electron interaction takes the form VR1 R2 R′1 R′2 c†R1 σ c†R′ σ′ cR′2 σ′ cR2 σ H1 = 1

(7.41)

R1 R2 σ R′1 R′2 σ′

with the interaction matrix element now being the four-center integral: 1 d3 r d3 r′ φ∗ (r − R1 )φ∗ (r ′ − R′1 ) VR1 R2 R′1 R′2 = 2 e2 φ(r ′ − R′2 )φ(r − R2 ). × (7.42) 4πε0 |r − r ′ |

7.2 The Hubbard Model

203

With the same arguments as used in Sect. 5.4, it is conceivable that the dominating contributions come from two-center terms with R1 = R2 = R and R′1 = R′2 = R′ , which give the approximate form VRR′ c†Rσ c†R′ σ′ cR′ σ′ cRσ = (7.43) VRR′ nRσ nR′ σ′ . Hint ≃ RR′ σσ′

RR′ σσ′

The last expression results by identifying the number operator nRσ = c†Rσ cRσ and applying the anticommutation rules for (Rσ) = (R′ σ ′ ). If only the term with R = R′ is considered, which describes double occupancy of the site R with electrons of opposite spins, we arrive at the single-band Hubbard 2 Hamiltonian [207, 208] † † H = ǫ0 cRσ cRσ + t cR+dσ cRσ + U nR↑ nR↓ . (7.44) Rσ

Rdσ

R

Here U = VRR , called the Hubbard–U, weighs the strength of the correlation in case of two electrons at the same site. This Hamiltonian contains two competing mechanisms: the hopping term is responsible for band formation and delocalization, while the correlation term, with its energy increase for double occupancy of sites with electrons of opposite spin, favors a ground state with localized magnetic moments of uncompensated spins at each site. In the following, we replace R by the site index i and consider the Hubbard Hamiltonian (7.44) in the form H = H0 + H1 =

ijσ

1 niσ ni−σ , tij c†iσ cjσ + U 2 iσ

(7.45)

where tij is the hopping matrix element between two sites (which need not be the nearest neighbors). In order to calculate the spectrum of H, we calculate the Green function describing the propagation of an electron with spin σ from site j at time t′ to site i at time t i G(ijσ; t, t′ ) = − θ(t − t′ ){ciσ (t), c†jσ (t′ )}. h ¯

(7.46)

This is done by solving the equation of motion i¯ h

∂ i G(ijσ; t, t′ ) = δ(t − t′ )δij − θ(t − t′ ){[ciσ (t), H], c†jσ (t′ )}. ∂t ¯ h

(7.47)

The commutator of ciσ with H can be evaluated with [ciσ , c†jσ′ cmσ′ ] = cmσ δij δσσ′ and 2

John Hubbard 1931–1980.

(7.48)

204

7 Correlated Electrons

[ciσ , njσ′ nj−σ′ ] = δij ciσ ni−σ (δσσ′ + δσ−σ′ ) giving [ciσ , H] =

tim cmσ + U ni−σ niσ .

(7.49) (7.50)

m

Thus, the equation of motion (7.47) reads i¯ h

∂ G(ijσ; t, t′ ) = δ(t − t′ )δij + tim G(mjσ; t, t′ ) + U Γ (iiijσ; t, t′ ) (7.51) ∂t m

with a higher order Green function of the general form i Γ(ilmjσ; t, t′ ) = − θ(t − t′ ){c†i−σ (t)cl−σ (t)cmσ (t), c†jσ (t′ )}. h ¯

(7.52)

Instead of truncating the hierarchy by factorizing the higher order Green function in (7.51)(which would be the Hartree–Fock approximation to be considered later), we continue by writing its equation of motion as i¯ h

∂ Γ(iiij; t, t′ ) = ni−σ δij δ(t − t′ ) ∂t i − θ(t − t′ ){[ni−σ (t)ciσ (t), H], c†jσ (t′ )} . (7.53) h ¯

The commutator with H gives the two contributions (Problem 7.3) tim ni−σ cmσ + c†i−σ cm−σ ciσ − c†m−σ ci−σ ciσ [ni−σ ciσ , H0 ] =

(7.54)

m

and [ni−σ ciσ , H1 ] = U ni−σ ciσ

(7.55)

which leads to

i¯ h

∂ − U Γ(iiijσ; t, t′ ) = ni−σ δij δ(t − t′ + tim Γ(iimjσ; t, t′ ) ∂t m +Γ(imijσ; t, t′ ) − Γ(miijσ; t, t′ ) . (7.56)

This equation of motion simplifies for the case tim = ǫ0 δim , i.e., for a band without dispersion, known also as the atomic limit, to ! ∂ i¯ h − ǫ0 − U Γ(iiijσ; t, t′ ) = ni−σ δij δ(t − t′ ). (7.57) ∂t After Fourier transformation, it yields the solution Γ(iiijσ; E) =

n−σ δij , E − ǫ0 − U

(7.58)

7.2 The Hubbard Model

205

E

ε0+U

HF l –

ε0

Dσ(E) Fig. 7.3. Density of states for the Hubbard model in the atomic limit and comparison with the HF result

with n−σ = ni−σ independent of the site index i due to the translation invariance of the system. For the same situation, (7.51) takes a form, which after Fourier transformation leads to G(iiσ; E) =

E − ǫ0 − U (1 − n−σ ) (E − ǫ0 )(E − ǫ0 − U )

(7.59)

and can be written in the form G(iiσ; E) = lim

δ→0

n−σ 1 − n−σ + E − ǫ0 + iδ E − ǫ0 − U + iδ

!

.

(7.60)

The surprising result is that G(iiσ; E) has two poles, one at E = ǫ0 − iδ with the weight Z = 1 − n−σ and one at E = ǫ0 + U − iδ with the weight Z = n−σ (Fig. 7.3). This pole structure is a consequence of the energy dependence of the self-energy. The second term vanishes for n−σ → 0 and the remaining first term reduces for U = 0 to the Green function of the noninteracting electron G(0) (iiσ; E) = lim

δ→0

1 . E − ǫ0 + iδ

(7.61)

The eﬀect of considering the correlation U is the splitting of the single-particle level into two spin independent quasi-particle levels (separated by U ) whose weights depend on the occupation. This is clearly seen also in the density of states 1 Tr ImG(iiσ; E) π = 1 − n−σ δ(E − ǫ0 ) + n−σ δ(E − ǫ0 − U )

Dσ (E) =

(7.62)

shown in Fig. 7.3. The new quality of the result can be emphasized also by comparing with the Hartree–Fock approximation in (7.51), which means to replace Γ(iiijσ) by n−σ G(ijσ). The result for the case tim = ǫ0 δim is

206

7 Correlated Electrons

GHF (iiσ; E) =

1 , E − ǫ0 − U n−σ + iδ

(7.63)

which has a single pole only with weight Z = 1 at E = ǫ0 + U n−σ as indicated in Fig. 7.3. Let us now consider the band formation by relaxing the approximation tim ∼ δim . Then, (7.56) can be factorized with (for i = m) Γ(iimjσ) → n−σ G(mjσ)

Γ(imijσ) → c†i−σ cm−σ G(ijσ)

Γ(miijσ) → c†m−σ ci−σ G(ijσ),

where the last two terms compensate each other in (7.56) due to tim = tmi and one obtains with τ = t − t′ ! ∂ − ǫ0 − U Γ(iiijσ; τ ) = n−σ δij δ(τ ) + i¯ h tim G(mjσ; τ ) . (7.64) ∂τ m=i

The Fourier transform of this equation Γ(iiijσ; E) =

n−σ δij + tim G(mjσ; E) E − ǫ0 − U

(7.65)

m=i

can be used to write, instead of (7.51), the equation (E − ǫ0 )G(ijσ; E) = δij + 1+ tim G(mjσ; E) m=i

U n−σ E − ǫ0 − U

!

. (7.66)

Finally, we transform into the Bloch representation, which for the onedimensional model with lattice constant a means 1 −ik(i−j)a G(kσ; E) = e G(ijσ; E), (7.67) N ij to find (E − ǫ0 )G(kσ; E) = 1 − ǫk G(kσ; E) 1 +

with ǫk =

n=m tim

U n−σ E − ǫ0 − U

!

(7.68)

exp (−ik(n − m)a). This can be solved to yield G(kσ; E) =

1 E − ǫk − Σ(σ; E)

(7.69)

E − ǫ0 . E − ǫ0 − U (1 − n−σ )

(7.70)

with the self-energy Σ(σ; E) = U n−σ

7.2 The Hubbard Model

207

E

ε0+U EF ε0

insulator

metal

t

Fig. 7.4. Spectrum of the Hubbard model in dependence on the hopping matrix element t to indicate the insulator–metal transition

Due to the energy dependence of the self-energy, the Green function (7.69) has two separate poles with diﬀerent spectral weights (as for the atomic limit (Problem 7.4)) for each k. The spectrum, obtained by projecting the eigenvalues onto the energy axes, exhibits two bands whose widths are determined by that of the single-particle energy dispersion ǫk (note that the self-energy (7.70) does not depend on k). In tight-binding approximation, the dispersion of the energy band becomes ǫk = ǫ0 + 2t cos ka with t being the hopping matrix element between nearest neighbors. In Fig. 7.4, the spectrum of the Hubbard model is depicted for fixed U in dependence on t. Two bands evolve with increasing t from the two levels of Fig. 7.3, the upper and lower Hubbard band: they are separated by a gap as long as the band width 4t is smaller than U , but they overlap for 4t > U . For a system with one electron per lattice site, a single band can be completely filled. Thus for 4t > U the overlapping Hubbard bands are partially filled (the Fermi energy is in a region with finite density of states) and the model describes a metal, while for 4t < U the Hubbard bands are separated by a gap, the lower (upper) band being completely filled (empty), and the model describes an insulator. This metal–insulator transition is the important result of the Hubbard model. Systems showing this behavior are called Mott3 – Hubbard insulators. For the case of half-filling, the Hubbard Hamiltonian can be transformed into a Hamiltonian of the Heisenberg type (Problem 7.5), that allows one to describe the magnetic properties of these systems. However, the Hubbard model is not correct for small values of U , because it always gives the two separate eigenvalues for each k, while for small correlation we expect the single band solution. 3

Sir Nevill Mott 1905–1996, Nobel prize in Physics 1977, together with Philip W. Anderson and John H. Van Vleck.

208

7 Correlated Electrons

For more realistic descriptions, there exist extensions of the Hubbard model which also include transfer to next nearest neighbors, more than one band, and correlation terms including diﬀerent sites. One of these extensions is the subject of Problem 7.6. These models are used to investigate material systems whose electronic properties are determined by correlation [209–213]. Among those are the high-Tc cuprate superconductors. Their crystal structure contains layers with Cu and O atoms in a quadratic lattice. The overlapping Cu 3d and O 2p orbitals form a narrow p-d hybrid band. Due to the other constituents of the crystal structure, this band can be partially depopulated (or filled with holes). Thus, by comparing photoemission data for diﬀerent filling, the spectral weight of the evolving Hubbard bands becomes evident. The results of this section have uncovered a deficiency of the DFT-LDA concept. Making use of an eﬀective single-particle potential, which is the same for all electrons, irrespective of their orbital, this approach does not account for the correlation eﬀect outlined in presenting the Hubbard model, i.e., it is not capable of describing occupation-dependent energy bands. This deficiency is not problematic for energy bands deriving from s or p orbitals, for which the nearly free electron or pseudo-potential approach applies as well as the LCAO method because the orbital character is diminished due to delocalization. In contrast, for the localized d and f orbitals, this is not the case . In order to overcome this deficiency, the LDA+U concept has been developed [195]. Its essential idea is to replace the LDA Coulomb energy U N/(N − 1)/2 for d − d interaction, which is assumed to be part of the LDA energy functional ELDA [n], by the Hubbard correlation by writing 1 1 ni nj , E[n] = ELDA [n] − U N (N − 1) + U 2 2

(7.71)

i=j

where N = i ni is the number of electrons and ni the orbital occupancy. The orbital energies ! ∂E[n] 1 − n ǫi = (7.72) = ǫLDA + U i i ∂ni 2 are changed according to their occupancy with respect to the LDA value. A more recent extension of this concept is the dynamical mean-field theory (DMFT) [210, 214].

7.3 Fermi Liquids Classical liquids are known to exist due to particle–particle interaction, when the average interaction energy cannot be neglected in comparison with the average thermal energy kB T , and condensation, i.e., the phase transition from the gaseous to the liquid phase, takes place. In this condensed phase, the thermal motion is comparable with the mean particle separation. By reducing the

7.3 Fermi Liquids

209

temperature, the thermal motion can get so small that the interaction dominates the kinetic energy. In this situation, the phase transition from the liquid to the solid state, takes place. This systematic, which is based on classical arguments, does not account, however, for quantum eﬀects as has become apparent for He due to its small mass. At suﬃciently low temperatures, liquid He does not condense into the solid state. Instead, when the thermal deBroglie wavelength λT = (¯ h2 /2M kB T )1/2 , a quantum mechanical length scale, becomes comparable to the average particle separation (while at the same time the energy of the zero-point motion is much larger than the interaction energy), it enters a state known as quantum liquid. This phase transition is ruled by quantum statistics and leads to a Fermi liquid for 3 He but to a Bose liquid for 4 He. As the interacting particles are neutral He atoms, these two phases are jointly denoted as neutral quantum liquids. Especially for 3 He, Landau developed a theory of Fermi liquids [197], which comprises the interplay between Fermi statistics and particle interaction. As it turned out later on, this theory applies as well to interacting electrons in metals and doped semiconductors, which can be classified as charged Fermi liquids. We note in passing that Landau’s Fermi liquid theory is used also for neutron stars. In this section, a brief outline is given of this theory, which is closely connected with the concept of quasi-particles, and we refer for further reading of the literature [64, 194, 197–199]. Fermion systems without interaction, as treated in Chap. 4 in the Sommerfeld model or in Chap. 5 in the independent particle model of electronic band structure, can be characterized by their ground state and low-energy or elementary excitations out of the ground state. The former is defined for T = 0 K by filled states up to the Fermi energy, the latter are particle–hole excitations across the Fermi surface. Here, particle(hole) means an electron(missing electron) above(below) the Fermi energy. We have noticed already that the eﬀect of the interaction is to modify the single-particle energy ǫkσ by a self-energy, which incorporates interaction eﬀects (to an extent depending on the applied approximation, see e.g., (7.27) and (7.28)) into the independent particle picture thus leading to the concept of quasi-particles, which we have addressed already in Sects. 4.4, 5.2, and 7.1. Looking at the quantum numbers, momentum p = h ¯ k and spin σ (the band index is suppressed here), we have found a one-to-one relation between an independent particle and the corresponding quasi-particle, although via the self-energy, the quasi-particle energy ǫkσ [nkσ ] becomes a functional of the occupation numbers. Thereby, the spin is not changed and the quasi-particles remain fermions. The self-energy quantifies virtual excitations of electron–hole pairs, which represent charge or spin density waves. Consequently, a quasiparticle is the bare particle of the noninteracting system dressed by a cloud of virtually excited density waves. The lifetime of a quasi-particle, defined by the imaginary part of its self-energy, is determined by the scattering processes which take place under the constraints of energy and momentum conservation. At T = 0, this leads to an infinite lifetime for quasi-particles at the Fermi

210

7 Correlated Electrons

energy, because the available phase space for scattering shrinks to zero, while the scattering rate increases quadratically with the energy separation from the Fermi energy (or at finite T from the chemical potential) [122, 200]. Thus, quasi-particles are well defined for low-energy excitations. This leads to the concept of Fermi liquids: it is based on the assumption that the excitation spectrum of the interacting Fermi system is similar to that of the noninteracting Fermi system and that the particles (or states) of the latter evolve one-to-one into the quasi-particles (or states) of the former without changing the quantum numbers when the interaction is adiabatically switched on. This concept is supported by the observation that in a certain temperature range some properties (specific heat, spin susceptibility) of many metals (like those of 3 He) correspond, in their temperature dependence, to those of the noninteracting Fermi system, however, with changed kinematic properties such as the particle mass. Elementary excitations can be described as changing the occupation of states around the chemical potential (or the Fermi energy) with respect to the ground state occupation n0kσ +1 |k| > kF δnkσ = nkσ − n0kσ = (7.73) −1 |k| ≤ kF . The total energy of the system is a functional of the occupation numbers nkσ E = E[nkσ ] but E = ǫkσ nkσ , (7.74) kσ

because the quasi-particle energy ǫkσ = δE/δnkσ depends on the occupation due to the self-energy and has in general a nonvanishing variational derivative δǫkσ δ2E = =: f (kσ; k′ σ ′ ) = 0. δnk′ σ′ δnkσ δnk′ σ′

(7.75)

We may express the total energy as a Taylor series with respect to the elementary excitations (i.e., the changes in the occupation numbers) about the ground state energy E0 E[nkσ ] = E0 + ǫkσ [n0kσ ]δnkσ kσ

1 + f (kσ; k′ σ ′ )δnkσ δnk′ σ′ + O(δn3 ). 2 kσ

(7.76)

k′ σ′

Denoting the quasi-particle energies for the ground state distribution ǫkσ [n0kσ ] by ǫ0kσ , the quasi-particle energies can be expressed as ǫkσ =

δE ≃ ǫ0kσ + f (kσ; k′ σ ′ )δnk′ σ′ . δnkσ ′ ′ kσ

(7.77)

7.3 Fermi Liquids

211

The adequate thermodynamic potential for the grand-canonical ensemble with varying occupation N = kσ nkσ is the free energy F = E − μN which under elementary excitations with N − N0 = kσ δnkσ , where N0 is the total occupation in the ground state, changes by F − F0 = E − E0 − μ(N − N0 ) 1 = f (kσ; k′ σ ′ )δnk′ σ′ δnkσ . ǫ0kσ − μ δnkσ + 2 ′ ′

(7.78)

kσ

kσ

Note that we consider a situation where the quasi-particle energies are close to the chemical potential μ and that ǫ0kσ − μ = 0 only for kσ with δnkσ = 0; thus, the first term is of the order (δnkσ )2 . As for the noninteracting particles, the free energy is stationary for the equilibrium distribution function −1

nkσ = [1 + exp (β(ǫkσ − μ))]

,

(7.79)

with the dispersion of the independent particles replaced by that of the quasi-particles. It has the form of the Fermi–Dirac distribution function but is an implicit equation for nkσ due to the functional dependence of the quasi-particle energies ǫkσ on the occupation. In contrast to a microscopic theory, which aims at a calculation of the quasi-particle energies, the Fermi liquid theory replaces the interaction by parameters and relies on the one-to-one correspondence between independent (or bare) particles and quasi-particles including their statistics. This is outlined in the following by, assuming for better transparency, an isotropic Fermi liquid and a spin degenerate dispersion. The Fermi velocity is defined by vk,F =

" ¯hkF 1 "" ∇k ǫkσ |k=kF " =: ∗ , h ¯ m

(7.80)

where m∗ denotes the eﬀective mass of the quasi-particle at the Fermi energy. Let us assume the quasi-particle dispersion ǫkσ to be a suﬃciently smooth function in the vicinity of the Fermi energy EF (or the chemical potential μ). Then, we can write " " ¯ 2 kF h ǫkσ − μ ≃ "∇k ǫkσ |k=kF "(k − kF ) = (k − kF ) m∗

(7.81)

as depicted in Fig. 7.5. The same relation holds for the independent particle, however, with a mass m instead of m∗ . This diﬀerence in the masses is due to the fact that the quasi-particle consists of the bare particle and a cloud of density fluctuations around it, which moves along with the particle and reduces its mobility, i.e., m∗ > m. One immediate consequence is the enhancement of the density of states at the Fermi energy D(EF ) =

m∗ kF 1 δ(ǫkσ − EF ) = 2 2 , V π h ¯ kσ

(7.82)

212

7 Correlated Electrons

εkσ µ

kF

k

Fig. 7.5. Quasi-particle dispersion (solid) and linear approximation around the chemical potential (dashed)

which implies an enhancement of all quantities which are proportional to the D(EF ) as e.g., the particle contribution to the specific heat. The eﬀect of the interaction is considered here in the parameter m∗ , which is accessible by measuring the Sommerfeld coeﬃcient (see Sect. 4.2). Let us now take into account the interaction between quasi-particles represented by f (kσ; k′ σ ′ ). For the isotropic Fermi liquid and dominating exchange interaction (as in Sect. 6.2), this can be written [64] as f (kσ; k′ σ ′ ) = f s (k, k′ ) + σ · σ ′ f a (k, k′ ).

(7.83)

Being close to the Fermi energy, we have |k|, |k′ | ≃ kF and because of the isotropy, the f λ (k, k′ ) depend only on the angle θ between k and k′ . Therefore, we may expand these functions in terms of Legendre polynomials (normalized for practical reasons by the density of states D(EF )) ∞

f λ (k, k′ ) =

1 λ Fl Pl (cos θ). D(EF )

(7.84)

l

The coeﬃcients Flλ are the phenomenological Fermi liquid parameters which have to be determined by comparison with experimental data. As it turns out [64, 122], the quasi-particle eﬀective mass is given by m∗ = m (1 + F1s /3) .

(7.85)

Thus, F1s could be determined from the low-temperature behavior of the specific heat (provided the system is isotropic). The Pauli spin susceptibility is enhanced due to interactions and can be expressed as χspin =

m∗ χ0 , m(1 + F0a ) spin

(7.86)

with the spin susceptibility χ0spin of free electrons. Its experimental value provides the parameter F0a , if the eﬀective mass is already known from specific heat data. Fermi liquid parameters are reported so far only for He [64, 199, 215].

7.4 Luttinger Liquids

213

nkσ 1 < <

Z

kF

k

Fig. 7.6. Particle (dashed) and quasi-particle (solid) distribution at T = 0 K. The discontinuity at kF represents the weight Z

Finally, we look at the momentum distribution nkσ , which can be obtained from the Green function with the help of the dissipation–fluctuation theorem +∞ ImG(kσ; ǫ) 1 † nkσ = ckσ ckσ = − dǫ, (7.87) hπ −∞ 1 + e−βǫ ¯ with G(kσ; ǫ) from (7.31). The denominator simplifies for T = 0 K and the integration can be extended to a contour in the upper complex plane. This includes only the quasi-particle poles for k < kF (which are in the upper half plane) but not those for k > kF (which are in the lower half plane) and results in a jump of nk at k = kF , which equals the weight Z(kF ) of the quasi-particle pole (Fig. 7.6). The calculation of this jump is the subject of Problem 7.7. As we have seen by these considerations, the Fermi surface introduced in Chap. 4 for the noninteracting electrons exists also in the presence of the electron–electron interaction, as can be seen from the discontinuity in the momentum distribution (Fig. 7.6). It is the signature of existing quasi-particles. However, when replacing the independent particles by quasiparticles, the kinematic properties and consequently also physical quantities change as compared with the independent particle result. This can be exploited to determine the parameters of the Fermi liquid theory. The incorporation of particle interaction in the quasi-particle depends essentially on the phase space available for interaction processes close to the Fermi energy. While in three and two dimensions, this phase space is a sphere or a circle, respectively, it shrinks to two points for a one-dimensional system. The dimensionality eﬀect can be demonstrated for the Lindhard function (Problem 7.8). We shall see in the next section how this will dramatically alter the situation when we consider one-dimensional electron systems.

7.4 Luttinger Liquids The dimensionality and its influence on solid state properties have been mentioned already in several sections. Obviously the two-dimensional systems are the surface of a solid (Sect. 3.6) and the interface between two diﬀerent

214

7 Correlated Electrons

solids (a heterostructure). We have learnt about semiconductor heterostructures that they can accommodate a two-dimensional electron gas (Sect. 5.6). In this section the one-dimensional electron systems will be the focus of interest. There are several realizations, which have stimulated the investigation of such systems. Among those are special molecular crystal structures such as inorganic and organic linear chain compounds that allow band formation by overlapping atomic orbitals in one spatial direction only [201]. Another example is conducting polymers, for which energy bands arise from repeated conjugated bonds along the strand [111, 112]. But one may start also from the 2D electron systems in heterostructures to prepare by etching, cleaved edge overgrowth, or depletion via top gates a 1D channel (a quantum wire) along which electrons can move freely [216–218]. The youngest child in this family is carbon nanotubes, which can be understood as a graphite monolayer rolled up to a cylinder with a diameter of a few nm [21, 113, 114]. All these systems have been and are still under investigation due to their peculiar properties determined by the low dimensionality which do not fit into the framework of Fermi liquid theory. The essential point here is the breakdown of the quasi-particle concept [22, 122, 193, 219–222]. Free electrons in one dimension would be characterized by a quadratic dispersion relation which cuts the Fermi energy at k = ±kF as depicted in Fig. 7.7. Being interested in elementary excitations around the Fermi energy, it is advantageous to linearize the dispersion relation as in the previous section by writing (relative to the chemical potential) ǫ0±,k ≃

¯ 2 kF h (±k − kF ) = h ¯ vF (±k − kF ) , m

(7.88)

with the Fermi velocity vF . This spectrum consists of two branches (linear in k) that correspond to electrons traveling left and right along the extension of the 1D system. We can immediately write down the corresponding singleparticle part of the Hamiltonian in terms of fermion operators ({ckα , c†k′ α′ } = δα,α′ δk,k′ )

ε°k

EF

–k F

kF

k

Fig. 7.7. Dispersion relation for a 1D fermion system with linear approximation around the Fermi energy

7.4 Luttinger Liquids

Hsp = h ¯ vF

k,α=±

(αk − kF ) c†kα ckα − c†kα ckα 0 .

215

(7.89)

Here, the ground state expectation value c†kα ckα 0 of the number operator is subtracted to prevent a divergence of the ground state energy due to occupation of states with negative energy. In the following, we assume a system length L and apply periodic boundary conditions with the consequence of discretizing k in multiples of 2π/L. Number fluctuations would be described here by † nqα = (7.90) ck+qα ckα − δq,0 c†kα ckα 0 , kα

which according to their commutation rule [nqα , n−q′ α′ ] = αδα,α′ δq,q′

qL 2π

(7.91)

are boson operators. Moreover, we have hvF qnqα , [Hsp , nqα ] = α¯

(7.92)

indicating that the number fluctuations created by nqα are eigenstates of Hsp with the eigenvalue α¯ hvF q. This leads to an alternative formulation of Hsp ⎞ ⎛ π¯ hvF ⎝ Hsp = nqα n−qα + n20α ⎠ , (7.93) L α q=0,α

now in terms of boson operators. A closer inspection shows that the existence of these two equivalent formulations of H0 , (7.89) and (7.93) is characteristic for the 1D case and the linearized dispersion relation. Electron–electron interaction can take place within each branch of the spectrum or between the two diﬀerent branches. For small momentum transfer, corresponding to forward scattering, this can be written in terms of boson operators as 1 Hint = vq (nqα n−qα, + nqα n−q−α ) , (7.94) 2L q,α where vq quantifies the strength of this scattering. The system Hamiltonian H = Hsp + Hint is a bilinear expression in the boson operators and can be diagonalized by a Bogoliubov transformation n ˜ qα = nqα cosh(φ(q)) + nq−α sinh(φ(q)),

(7.95)

with the interaction parameter e

2φ(q)

=

vq 1+ π¯ hvF

!−1/2

=: K(q).

(7.96)

216

7 Correlated Electrons

hωq

2kF

q

Fig. 7.8. Particle–hole excitations in the 1D fermion system. Note the missing of low-energy excitations for 0 < |k| < 2kF

The diagonal form of H reads π π vN N 2 + vJ J 2 , H= vq n ˜ qα n ˜ −qα + L 2L

(7.97)

q =0

¯ vF /K(q), vN = v02 /¯ hvF , and vJ = h ¯ vF , while, N = n0+ + where vq = h n0− and J = n0+ − n0− describe charge and current excitations respectively. Ultimately, by introducing normalized boson operators 9 2π † (θ(q)˜ nq+ + θ(−q)˜ bq = nq− ) , (7.98) L|q| with the step function θ(q), the Hamiltonian takes the form π vN N 2 + vJ J 2 H= hωq b†q bq + ¯ 2L

(7.99)

q=0

with the dispersion h ¯ ω(q) = vq |q|. It is the model Hamiltonian for interacting 1D electrons named after Tomonaga4 [223] and Luttinger5 [224]. The remarkable property of this Hamiltonian is that it is expressed in terms of collective excitations, which are the low-energy excitations of the system similar to the vibrations of a string. This means that low-energy electron–hole pair excitations are absent here, which indicates also the absence of quasiparticles for interacting electrons in 1D. This finding can be confirmed by looking at the excitation spectrum in Fig. 7.8. At low energies, excitations are possible only between states close to the two Fermi points (see Fig. 7.7), but there are no particle–hole excitations and, therefore, no quasi-particles for 0 < |k| < 2kF . This nonexistence of quasi-particles can be made explicit by 4 5

Sin-itiro Tomonaga 1906–1979. Joaquin Luttinger 1923–1997.

7.5 Heavy Fermion Systems

217

Fig. 7.9. Schematic view of charge and spin separation in a linear spin chain with antiferromagnetic order: a charge excitation (hole) is created (upper part) and moves away by hopping leaving a spin excitation behind (second and third line)

calculating the momentum distribution, which for the Tomonaga–Luttinger model does not exhibit the quasi-particle discontinuity of the Fermi liquids at kF (see Fig. 7.6) but instead an infinite slope. Including the spin, the model Hamiltonian H would contain additional terms corresponding to collective spin excitations. As they are additive, the propagation of collective charge and spin excitations takes place with diﬀerent velocities. This separation of charge and spin is another characteristic feature of a Luttinger liquid. It can be visualized for a spin chain with antiferromagnetic order (Fig. 6.6), as described by a 1D Hubbard model with half-filling and negative exchange coupling which is a lattice variant of the Luttinger liquid. A charge excitation corresponds to removing an electron with its spin (or to create a hole). This hole is surrounded by two parallel spins (upper part of Fig. 7.9). Its motion (to the left, see lower parts of Fig. 7.9) is ruled by the transfer matrix element and not connected with spin-flips thus, the moving hole is always surrounded by a pair of up/down spins, while the parallel spins remain at the site, where the hole was created.

7.5 Heavy Fermion Systems Metallic systems with the Fermi energy in a range of the spectrum, where strongly localized states deriving from f electrons hybridize with delocalized states with s, p, or d character, show Fermi liquid behavior, however, with a strong enhancement of the Sommerfeld coeﬃcient γ in the specific heat and also of the Pauli spin susceptibility χPauli (see Sect. 7.3). In an independent particle description, both quantities are proportional to the density of states at the Fermi energy, which for isotropic systems is proportional to the eﬀective mass m∗ of the quasi-particles. Thus, the observed enhancement can be understood as a dramatic increase of the eﬀective mass and these materials have been named, therefore, heavy fermion systems [110, 122, 225].

218

7 Correlated Electrons

CV

γT

CeAl 3

T*

LaAl 3 T

Fig. 7.10. Low-temperature behavior of the specific heat for a heavy fermion system (CeAl3 ) compared with a normal metal (LaAl3 ). The characteristic temperature T ∗ indicates the transition from heavy fermion to normal fermi liquid behavior (after [122])

In general, heavy fermion systems are compounds with ions containing partially filled 4f (lanthanides) of 5f shells (actinides). For further charac2 Pauli terization of their properties, one looks at the ratio R = π 2 kB χ /3μ2eﬀ γ, where μeﬀ is the eﬀective magnetic moment of the quasi-particles. For noninteracting quasi-particles, R = 1, because the eﬀective mass drops out. In contrast, γ and χPauli are modified diﬀerently by quasi-particle interactions which is reflected in R = 1. The enhancement of γ and χPauli occurs only in a small temperature range below a characteristic temperature T ∗ of a few Kelvin. Above T ∗ , it disappears and these systems behave like normal Fermi liquids. This can be seen by comparing the temperature dependence of the specific heat of CeAl3 and LaAl3 (Fig. 7.10). The Ce compound diﬀers from the otherwise identical system with La by the partial occupation of the 4f shell. The appropriate Hamiltonian for modeling the electronic properties of heavy fermions is[122] H=

ǫk c†kσ ckσ +

im

kσ

1 +√ N

k,σ,i,m

† ǫf m fim fim +

U 2

nfim nfim′

i,m=m′

† ik·Ri . e Vkσm c†kσ fim e−ik·Ri + ckσ fim

(7.100)

The first term describes the dispersive energy band ǫk of the delocalized electrons with wave vector k and spin σ. The second term represents the singleparticle energies of the f electrons, localized at lattice sites Ri , where m = 1 . . . νf counts the orbital degeneracy, while the third term is Hubbard’s on-site Coulomb interaction of these electrons (with number operator nfim ). The fourth term accounts for the hybridization of the f electrons with the delocalized Bloch electrons. Equation (7.100) is a variant of the Anderson6 Hamiltonian, which was originally formulated for only one site 6

Philip W. Anderson, *1923, Noble prize in Physics 1977 together with Sir Neville Mott and John H. Van Vleck.

7.5 Heavy Fermion Systems

219

occupied with an ion carrying f electrons (Anderson impurity model). This Hamiltonian combines two aspects of electronic structure: The hybridization (see Sect. 5.4) (which converts the localized f level into a band with finite width) with the correlation of the Hubbard model (which prevents double occupancy of f states at a given site). Without the correlation term, the problem can be diagonalized as in Sect. 5.4. The correlation can be taken into account by a proper modification of the hybridization or hopping matrix element, which ensures that a conduction band electron does not hop to an already singly occupied f state. In fact, according to Fermi’s golden rule, the rate of hopping between the conduction band and an f state is proportional to this matrix element squared. If it is renormalized by a factor 1 − nf , where nf is the occupancy of the f state, the rate will be reduced as desired. For the hopping matrix element, this means the replacement Vkσm → rVkσm = V˜kσm ,

(7.101)

where r2 = 1−nf . In principle, this modification has to be done site dependent and with the number operator of the localized f electrons. By using r in the following as a parameter, a mean-field approach has been adopted, by which these site-dependent number operators are replaced by their mean value nf . The relation between r and nf is to be included as a subsidiary condition with the Lagrange parameter Λ in the Hamiltonian † HMF = fkm ǫ˜f m fkm ǫk c†kσ ckσ + kσ

+

k,σ,m

im

† rVkσm c†kσ fkm + fkm ckm + ΛN (r2 − 1).

(7.102)

Here, the fermion operators of the f electrons are in the Bloch representation and ǫ˜f m = ǫf m + Λ. This mean-field version of the Anderson Hamiltonian depends on the two parameters r and Λ which have to be determined at the end. It is advantageous to simplify the model by considering only a nondegenerate conduction band and an f orbital with degeneracy νf = 2 by dropping the indices m and σ. The diagonal form of the simplified mean-field operator HMF = El (k)b†kl bkl + ΛN (r2 − 1) (7.103) kl

is obtained by replacing the fermion operators in (7.102) according to ykl b†kl , c†k = xkl b†kl , (7.104) fk† = l

l

where the index l refers to the two branches of the hybridized bands. The coeﬃcients are determined by the normalization condition

220

7 Correlated Electrons

|ykl |2 + |xkl |2 = 1

(7.105)

and the coupled linear homogeneous equations ǫ˜f − El (k) ykl + V˜k xkl = 0 V˜k∗ ykl + ǫk − El (k) xkl = 0.

(7.106)

These equations signalize a second order perturbation calculation with respect to the hybridization and by eliminating e.g., xkml we have V˜k V˜k∗ ykl = 0. ǫ˜f − El (k) + (7.107) El (k) − ǫk The coeﬃcient ykl is the probability amplitude for the quasi-particle in the band l to be in the f state. From (7.107), one obtains the two quasi-particle bands E± (k) =

1 ((ǫk + ǫ˜f ) ± W (ǫk )) , 2

with W (ǫk ) = (ǫk − ǫ˜f )2 + 4V˜ 2 .

(7.108)

They are depicted in Fig. 7.11, where a linear dispersion is assumed for the conduction band ǫk . The requirement that the lower branch crosses the chemical potential for k = kF , E− (kF ) = μ, allows one to determine from (7.107) the renormalized energy of the f level ǫ˜f = μ +

V˜ 2 , ǫkF − μ

(7.109)

which for weak hybridization is slightly above μ.

E ~ ε

f

µ

εf 0

1

k/kF

Fig. 7.11. Quasi-particle dispersion for heavy fermion behavior. The linear dispersion of the conduction band hybridizes with the renormalized level at ǫ˜f of the f state (after [122])

7.5 Heavy Fermion Systems

221

The next task is to determine the two parameters r and Λ of the model. One obvious condition is that the ground state expectation value of the mean-field Hamiltonian takes a minimum with respect to the f state occupation nf or the parameter r. This condition can be formulated as ∂/∂rΨ0 (r)|HMF |Ψ0 (r) = 0, which reduces because of the normalization of the ground state, to Ψ0 |∂HMF /∂r|Ψ0 = 0 or with (7.102) 1 ˜ Vk Ψ0 |c†k fk + fk† ck |Ψ0 + 2Λr2 = 0. N

(7.110)

k

The second relation is nf = 1 − r2 , reformulated as 1 Ψ0 |fk† fk |Ψ0 + (r2 − 1) = 0. N

(7.111)

k

The first terms in these relations with the expectation values can be expressed with the help of the expansions (7.104) and the possibility to create the ground state |Ψ0 by applying b†kl for |k| ≤ kF to the fermion vacuum νf 1 D(EF ) (ǫkF + W (ǫkF ) − W (0)) = nf (7.112) Ψ0 |fk† fk |Ψ0 = N 2 k ! 1 ˜ ǫkF − ˜ǫf + W (ǫkF ) . (7.113) V Ψ0 |c†k fk |Ψ0 = −νf D(EF )V˜ 2 ln N −˜ ǫf + W (0) k

This leads in the leading logarithmic approximation to V˜ 2 2 r = 1 − νf D(EF ) (ǫkF − μ) − μ

(7.114)

and 2

Λ = νf D(EF )V ln

(ǫkF − μ)μ V˜ 2

!

.

It is useful to define a characteristic temperature T ∗ by ! Λ , kB T ∗ = μ exp − νf D(EF )V 2

(7.115)

(7.116)

which allows one to express the renormalized energy of the f level as ǫ˜f = μ + kB T ∗

(7.117)

and by eliminating ǫkF to write nf = 1 −

kB T ∗ . νf D(EF )V 2

(7.118)

222

7 Correlated Electrons

Inspection of Fig. 7.11 sheds some light on the physics: Without hybridization the f level at ǫf would be occupied, nf = 1, and also the conduction band states up to kF . With hybridization, the f level is shifted above μ, thereby the occupation is reduced, and the conduction band is flattened out close to the chemical potential. The latter is connected with an increase of the density of states and a gain in energy, because the occupied conduction band states are now at lower energy. This gain in total energy can be quantified by comparing (0) the ground state energies without (E0 ) and with (E0 ) hybridization: (0)

E0 − E0

= −kB T ∗ .

(7.119)

Without giving further details of the calculation [122], one obtains for the Sommerfeld coeﬃcient in the specific heat γ=

1 2 2 nf π kB (2J + 1) 3 ǫ˜f − μ

(7.120)

and for the spin susceptibility χspin = (gJ μB )2

nf 2J + 1 . νf (˜ ǫf − μ) 3

(7.121)

Both quantities show a strong increase with ǫ˜f being close to the chemical potential μ. This explains the heavy fermion behavior.

7.6 Fractional Quantum Hall States In this chapter, we have seen so far that the influence of electron correlation increases with the localization of the electrons due to their orbital motion. We found that systems with electrons at the Fermi energy in atomic d and f orbitals are destined to exhibit eﬀects of the electron–electron interaction. There is quite a diﬀerent system, where the localization is not a genuine property of an atomic orbital, but where it is due the orbital motion enforced by applying a magnetic field. Remember the two-dimensional electron systems of Sect. 5.6 and the discussion of the quantum Hall eﬀects. A magnetic field applied perpendicular to the semiconductor hetero-interface forces the electrons into cyclotron orbits whose radius decreases with increasing magnetic field. This is the localization mechanism which eventually leads to a regime where correlation eﬀects become significant. It is the regime of the fractional quantum Hall eﬀect. This eﬀect has been discovered only a few years after the integer QHE. In 1982 Tsui, Stoermer, and Gossard [226] reported the observation of a QHE not at integer filling factors ν but at the fractional filling ν = 1/3 of the lowest spin-split Landau level in a high-mobility GaAs/AlGaAs heterostructure. In the following years, with increasing perfections of the heterostructure

7.6 Fractional Quantum Hall States

223

2.5 32

N=0

N=1

2.0

Pxy

[h /e2]

n =4 3

2

1

1.5 2/5

1.0

2/3

3 /2

3/7

1 /2

3/5

4/9 4/7 3/4

6/11

4/3

0.5

6/13

7/13

5/3 10/7

Pxx

5/11

5/9

11/9

0

9/7

0

5/7 4/5

8/5 7/5

0

5

10 15 20 MAGNETIC FIELD [T]

25

30

Fig. 7.12. Low-temperature data of the longitudinal (ρxx ) and Hall resistance (ρxy ) of a high-mobility two-dimensional electron system in a GaAs/AlGaAs heterostructure. N indicates the Landau level quantum number and ν the filling factor. After [227]

samples, a whole family of fractional quantum Hall states was discovered preferentially in the same material system (see Fig. 7.12) but also in semiconductor heterostructures with other material combinations. The characteristic features of the QHE, the plateaus of the Hall conductance, and the vanishing of the longitudinal magneto-resistance (in the first observation just a dip at ν = 1/3), appear systematically at filling factors ν=

p 2p ± 1

(7.122)

with integer p. Remarkable is here the odd denominator and the convergence of the two series of fractions for the ± sign from above and below toward the even fractional filling with ν = 1/2, for which no plateau is detected and the longitudinal magneto-resistance does not decrease to zero. Shortly after the discovery of the eﬀect (and before most of the other fractions were discovered), Laughlin came up with the fundamental idea of an incompressible quantum liquid caused by electron correlation and designed the corresponding N electron wave function [228]. Tsui, St¨ ormer, and Laughlin received jointly in 1998 the Nobel prize in physics for their discovery, which together with other contributions to the subject is well documented in several monographs [202–204].

224

7 Correlated Electrons

In order to present Laughlin’s idea, we have to briefly go through the singleparticle description of two-dimensional electrons in a perpendicular magnetic field. Supplement: Two-dimensional electrons in a magnetic field The system Hamiltonian (without Zeeman term) H=

1 (p + eA)2 2m∗

(7.123)

can be written with the symmetric gauge of the vector potential A = B(y, −x, 0)/2 in the form ! ω∗ m∗ ωc∗ 2 2 1 2 2 H= x + y 2 − c Lz . (7.124) p + p + x y ∗ 2m 2 2 2

The eﬀective mass m∗ (and the cyclotron frequency ωc∗ = eB/m∗ ) accounts for the fact that the electrons are in a subband deriving from the conduction band of a semiconductor. The symmetric gauge restores the cylindrical symmetry of the system according to which the z component of the angular momentum Lz commutes with the system Hamiltonian, which immediately can be recognized as that of two harmonic oscillators in the x, y planes, respectively. Accordingly, two sets of oscillator operators with standard commutation rules are introduced ! ! x x 1 1 il il a†x = √ − px , ax = √ + px h ¯ h ¯ 2 l 2 l ! ! 1 1 y y il il a†y = √ − py , ay = √ + py , (7.125) h ¯ h ¯ 2 l 2 l # # (¯ h/m∗ ωc∗ ) = (¯ h/eB) is the magnetic length (see Sect. 4.2). The where l = single-particle Hamiltonian H=

' hωc∗ & † ¯ ax ax + a†y ay + 1 − i ay a†x − ax a†y 2

can be diagonalized by oscillator operators 1 a†+ = √ a†x + ia†y , a+ = 2 1 † † a− = √ ax − ia†y , a− = 2

1 √ (ax − iay ) 2 1 √ (ax + iay ) , 2

(7.126)

(7.127)

which obey the commutation relations [a± , a†± ] = 1. They can be understood as operators for right/left circular rotations around the direction of the magnetic field. With these operators, one finds h hωc∗ † ¯ ¯ ωc∗ † a+ a+ + a†− a− + 1 + a+ a+ − a†− a− 2 2 ! 1 ∗ ˆ+ + =¯ h ωc n 2

H =

(7.128) (7.129)

7.6 Fractional Quantum Hall States

225

with eigenvalues

! 1 . (7.130) 2 Note that the eigenvalues depend only on the quantum number n+ while n− counts the level degeneracy. The eigenstates of H can be created by multiple application of the raising operators onto the oscillator vacuum % n+ n− 1 |00 . (7.131) a† a†− |n+ n− = n+ !n− ! + hωc∗ n+ + ǫn+ n− = ¯

Instead of n± , it is advantageous to use the quantum numbers n = min(n+ , n− ) and the angular momentum quantum number m = n+ − n− , which are related with the Landau level quantum number nL = n + (m + |m|)/2. (Note: in Fig. 7.12 nL is denoted N .) In combining (7.125) and (7.127), it is suggestive to introduce the complex dimensionless variable z = (x − iy)/l and the corresponding derivative ∂/∂z = (∂/∂x + i∂/∂y)l/2 along with the conjugate definitions. This gives the following convenient properties ∂ ∂ z−z =1, ∂z ∂z

∂ ∂ ∗ z − z∗ =0 ∂z ∂z

(7.132)

and similarly for the complex conjugates. The oscillator operators can now be written as ! ! 1 1 ∂ ∂ † ∗ z −2 , a+ = z+2 ∗ a+ = 2 ∂z 2 ∂z ! ! ∂ 1 ∂ 1 z − 2 ∗ , a− = z∗ + 2 (7.133) a†− = 2 ∂z 2 ∂z which lead to the Hamiltonian in dimensionless coordinate representation ! ! ∂ ∂ ∂ 1 ∗ 1 ∗ ∗ ∂ ∗ hωc −4 h ¯ ω − z H= ¯ . z − + zz c 2 ∂z ∂z ∗ 2 ∂z ∂z ∗

(7.134)

The lowest energy eigenfunction can be obtained from the conditions a± |00 = 0, which easily lead to the normalized wave function of the lowest Landau level.

Making use of the complex notation in the (x, y) plane, the normalized wave function for the lowest Landau level (nL = 0) is given by ! 1 −|z|2 /4 r2 1 √ ϕ00 (r) = exp − 2 = √ e . (7.135) 4l 2πl 2πl This wave function with angular momentum m = 0 is degenerate with finite angular momentum wave functions 2 1 ϕ0m (z) = √ z m e−|z| /4 , m 2π2 m!l

(7.136)

√ which have a maximum probability |ϕ0m (r)|2 on a circle with radius 2ml and a spread of the order of the magnetic length l. Thus, a sample of circular

226

7 Correlated Electrons

shape with radius R can accommodate only states with m values fulfilling the condition 2ml2 < R2 . This allows one to count the degeneracy of the Landau level. The maximum value of m is determined by the number of elementary flux quanta threading the sample area. This is the same result as the one obtained in Sect. 4.2 assuming the asymmetric Landau gauge. Considering fractional quantum Hall states, Laughlin constructed a N electron wave function from the single-particle wave functions of the lowest Landau level. Their general form, a linear combination of the ϕ0m , is ϕ(z) = f (z) exp (−|z|2 /4) with a polynomial f (z). The N -electron wave function, expressed as a linear combination of Slater determinants composed of these single-particle wave functions, has the general form N Ψ(z1 , . . . , zN ) = f (z1 , . . . , zN ) exp − |zi |2 /4 . (7.137) i=1

Here, f (z1 , . . . , zN ) is a polynomial in every variable zi and its individual terms are products of zimi indicating that the electron i is in an angular momentum eigenstate with mi . The symmetry of the problem requires that the total angular momentum h ¯M = i ¯ hmi is conserved and the wave function Ψ(z1 , . . . , zN ) should contain only terms with the same M . Thus, the polynomial f has to be homogeneous. The antisymmetry of the Slater determinants makes this polynomial also antisymmetric in the particle coordinates. On top of these symmetry requirements, the N -particle wave function, in order to describe a ground state, should by construction take into account that the electrons try to avoid each other due to the repulsive Coulomb interaction. This can be achieved by writing the polynomial as a product of functions g(zi − zj ) depending on the inter-particle separation. This form, which accounts for two-particle correlations, is known as a Jastrow-type wave function and was used before in atomic physics. Together with the general symmetry considerations g(z) has to be an odd power polynomial. Moreover, Ψ is an eigenfunction of the total angular momentum h ¯ M where M counts the powers of the zi , which are all the same and M = N (N − 1)m/2. Thus, one arrives at the N electron wave function N * 2 m |zi | /4 . (7.138) (zi − zj ) exp − Ψm (z1 , . . . , zN ) = i<j

i=1

This is Laughlin’s wave function. As we have seen, the maximum possible angular momentum of a single particle state is determined by the degeneracy of the Landau level which is the sample area divided by 2πl2 or the number Φ/Φ0 of flux quanta threading the sample. On the other hand, the maximum power (or angular momentum quantum number) of each z is given by m(N − 1) and we can equate (for N ≫ 1) Φ Φ 1 mN = or m = = . (7.139) Φ0 N Φ0 ν

7.6 Fractional Quantum Hall States

227

The last relation connects the angular momentum with the filling factor ν. For fractional filling ν = 1/m with odd m, the Laughlin wave function is an antisymmetric many-body wave function, where each electron position is an m-fold zero with respect to the dependence on all other electron positions. At the same time, m is the number of flux quanta (or vortices) attached to each electron. An interpretation of this many-body wave function is possible by looking at |Ψm (z1 , . . . , zN )|2 = e−Φm (z1 ,...,zN )

(7.140)

where Φm (z1 , . . . , zN ) = −2m

i<j

ln |zi − zj | +

N 1 |zi |2 2l2 i=1

(7.141)

is formally identical with the electrostatic energy of a charge-neutral twodimensional plasma, where the first term accounts for the electron–electron interaction and the second for the interaction with the neutralizing background. The corresponding expression would be obtained by replacing 2m → e2 and 1/2mπl2 by the particle density. Thus, |Ψm (z1 , . . . , zN )|2 is the classical probability distribution of N electrons in a plane with logarithmic interaction. The quality of Laughlin’s wave function has been tested by projection onto ground state wave functions, which for small N were calculated numerically by exact diagonalization. These projections were found to be very close to 1. Another result, that characterizes Laughlin’s wave function is the radial distribution function "2 N (N − 1) gm (|z1 − z2 |) = Ψm (z1 , . . . , zN )" d2 z3 . . . d2 zN . . . . 2 ρm (7.142)

g(x) 1.0

m=3 liquid crystal

0.5

0

0

1

2

3

4

x

Fig. 7.13. Radial distribution function for the Laughlin state with m = 3 (solid line) and the Wigner crystal state in the HF approximation √ (dashed line) for the same density. Here, x is the inter-particle distance in units of 2ml (after [203])

228

7 Correlated Electrons

For m = 3 it is shown in Fig. 7.13 (solid line) together with the radial distribution function for the Wigner crystal (dashed line) calculated in the HF approximation. It has been found that the energy of the Laughlin wave function is always lower than that of the Wigner crystal. The radial distribution function indicates that in contrast with the Wigner crystal, the Laughlin state does not exhibit a long-range order and can be identified as a liquid. A more detailed investigation shows also that the excitation spectrum out of the Laughlin state has an energy gap, classifying the fractional quantum Hall states as incompressible quantum liquids. The excitation spectra uncover several unexpected properties of these states, such as the hierarchy, the fractional charge, and the composite fermion concept, for which the reader is referred to the literature [202–204, 229, 230]. The geometry of a Hall eﬀect measurement is always connected with the finiteness of the sample with the edge of the sample representing a potential barrier for the electrons. Consequently, the Landau levels, which are constant except for fluctuations due to disorder (see Chap. 9), bend upwards towards the edge of the sample where they cross the Fermi energy. Thus, along the sample boundary, each Landau level represents a one-dimensional electron system, the so-called edge channel. In a classical picture, the edge channels correspond to skipping cyclotron orbits along the edge. The quantum Hall eﬀect can be understood as transmission between the diﬀerent probes of the Hall bar along these edge-channels [231, 232]. In the fractional QH regime, the edge-channels are seen also as a realization of one-dimensional systems of interacting electrons to which the Tomonaga–Luttinger model applies [204].

Problems 7.1 Calculate the commutators of ckσ and c†kσ with the interaction part Hint of the Hamiltonian and verify (7.20). 7.2 Derive the Dyson equation (7.25) by making use of the Fourier transform of (7.23) in the equation of motion for the Green function. 7.3 Evaluate the commutators of ni−σ ciσ with the single-particle and with the interaction term of the Hubbard Hamiltonian (7.45) for the onedimensional case to verify (7.54) and (7.55). 7.4 Calculate the spectral weight Z = (1−∂Σ/∂E)−1 for the lower and upper Hubbard band, especially for the center of the band at k = π/2a. 7.5 Consider the Hubbard Hamiltonian for the case of half-filling and weak hopping or strong correlation. Show by applying perturbation theory in second order and by introducing spin operators Siz =

1 zσ niσ , 2 σ

Siσ = c†iσ ci−σ

(7.143)

Problems

229

with z± = ±1, that the Hubbard Hamiltonian can be mapped onto an operator of the Heisenberg type. Identify the sign of the exchange coupling and the magnetic order of the ground state. 7.6 Set up the Hamiltonian for a system with localized electrons and on-site correlation (e.g., d or f electrons) overlapping with delocalized electrons deriving e.g., from atomic s states! This Hamiltonian is known in the literature as the Anderson Hamiltonian. 7.7 Calculate the momentum distribution nk at T = 0 by evaluating (7.87) in order to quantify the quasi-particle jump. 7.8 Evaluate the real part of Lindhard function (4.136) at T = 0 for the onedimensional and the three-dimensional electron systems and discuss the behavior for q → kF at ω = 0. While for d = 1 the calculation can easily be carried out for finite ω and be discussed for ω → 0 it is advantageous to do the calculation for d = 3 from the beginning for ω = 0.

8 Electron–Phonon Interaction

Based on the Born–Oppenheimer (or adiabatic) approximation (Chap. 2), the dynamics of the heavy and light constituents of a solid, the ions and the electrons, respectively, have been presented in the previous chapters as those of independent systems. For the lattice dynamics (Chap. 3), the electrons were considered only by their contribution to the binding forces which determine the dynamical matrix. For the electron systems, their energy spectrum and excitations (Chaps. 4–7), the position of the ions was kept fixed in the periodic configuration of a lattice, while the chemical nature of the ions determined the material specific properties. Releasing the Born–Oppenheimer approximation enables the two subsystems to communicate with each other by exchanging energy. This leads to a variety of eﬀects, which are not restricted to solids but are found in all types of condensed matter including macromolecular systems in chemistry and biology. The electrons experience the moving lattice as a perturbation of the periodic potential, which can be understood as scattering between electrons and phonons. Four scenarios will be considered here: 1. An excited electron gets rid of its excess energy by emitting phonons and at the same time changes its momentum. This leads to a finite lifetime of the carrier in its single-particle state. 2. An electron system in a solid driven by an external field (electric field, temperature gradient) adopts a nonequilibrium state. When the field is switched oﬀ, the electrons emit phonons and the system relaxes into an equilibrium state. This process of electron–lattice relaxation is characterized by a transport relaxation time. (In the same way an excited spin system can transfer its excess energy to the ion system and equilibrate by spin–lattice relaxation.) In thermodynamic terminology, the lattice is a heat bath serving as an energy sink or a reservoir. In exchanging energy with this bath, the relaxation processes change the phase of the individual electron wave function which makes these processes incoherent. By discussing these phenomena, electrical transport will become a

U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 8,

232

8 Electron–Phonon Interaction

topic of this Section. Diﬀerent physical scattering mechanisms, depending on the phonons involved, determine by their specific contributions the temperature-dependence of the electric conductivity. 3. Recalling the eﬀect of electron–electron interaction (Chap. 4), the electron– phonon interaction also can lead to a new ground state of the system. Due to its charge, an electron creates a polarization cloud in a polar lattice, which moves around with the electron and changes its dynamic properties. Thus, electron–phonon interaction leads to a new quasi-particle, the electron and its polarization cloud, called polaron because of the polar electron–phonon coupling causing the interaction. 4. Electron–phonon interaction can result also in an attractive electron– electron interaction if the phonon emitted by one electron is absorbed by another one within their lifetimes (virtual phonon exchange). This phonon-mediated electron–electron interaction favors the formation of pairing of electrons, which is one of the basic mechanisms of superconductivity. The diﬀerent aspects of electron–phonon interaction are subjects of most textbooks in Solid State Theory. For complementary reading, we refer here to [10, 13, 14, 89, 95, 233, 234].

8.1 Preliminaries Let us recall from Sect. 2.1 the separation of the Hamiltonian (2.1) H0 = Hel + Hion + Hel−ion .

(8.1)

The electron–ion interaction with the general form Hel−ion =

Ne l=1 n,τ

v(r l − Rnτ )

(8.2)

has been considered in the electronic energy Eel ({Rnτ }) as a contribution to the adiabatic potential U({Rnτ }) defined in (2.26), which was the starting point for the lattice dynamics in Chap. 3. This energy, obtained as the eigenvalue of the electron problem (2.25) in a static configuration of the ions, was assumed later to be that of the equilibrium configuration {R0nτ }. Here, we have to take into account the moving lattice with time-dependent positions Rnτ (t) = R0nτ + unτ (t) and do this by expanding around the equilibrium positions " Hel−ion = v(r l − R0nτ ) − ∇l v(r l − Rnτ )"R0 · unτ + . . . . (8.3) l,n,τ

l,n,τ

nτ

The first term has become part of the eﬀective periodic single-particle potential of the band structure problem (Chap. 5). In the second term, linear in

8.1 Preliminaries

233

the displacements unτ , ∇l means the derivative with respect to the position vector of the lth electron. This term will be considered in the following as the linear electron–phonon coupling Hel−ph . The higher order terms, indicated by dots, will be neglected. The linear approximation is suﬃcient for displacements that are small compared to the lattice spacing, as will be assumed throughout this chapter. It is advantageous for the illustration and for the evaluation of the interaction to make use of the occupation number representation and write it in terms of creation and annihilation operators. The lattice displacements can be formulated as (see Chap. 3, (3.23, 3.24) together with (3.39)) 0 1 unτ = √ Qs,q esτ (q)eiq·Rn N Mτ s,q 9 0 h ¯ esτ (q)eiq·Rn a†s (−q) + as (q) = 2N Mτ ωs (q) s,q

(8.4)

with the boson operators a†s (q) and as (q) of phonons with frequency ωs (q) and eigenvectors esτ (q). The time-dependence is not indicated here to simplify notation. The gradient of the potential is a single-particle term, which according to (4.76) can be written in terms of fermion operators for Bloch states (Problem 8.1). We want to simplify here the electron–phonon interaction within a single band, for which we adopt the eﬀective mass approximation. For this case, we may write, with the help of the Fourier transform of the potential, " ′ 0 ∇l v(r l − Rnτ )"R0 = i q ′ v(q ′ )eiq ·(rl −Rnτ ) (8.5) nτ

l

and use

l

′

eiq ·rl =

q ′ ,l

c†k+q′ ck

(8.6)

k

with the fermion operators c†k+q′ and ck for free electrons. Making use of (8.4) together with (8.5), we can perform the lattice sum in the second term of (8.3) with ′ 0 ei(q−q )·Rn = N δq−q′ ,G (8.7) n

G

and write the linear electron–phonon interaction in the convenient operator form Hel−ph = Vs (q − G) a†s (−q) + as (q) c†k+q−G ck (8.8) s,k,q,G

with the coupling matrix element 9 N¯ h (q − G) · esτ (q)e−i(q−G)·τ v(q − G), (8.9) Vs (q − G) = −i 2M ω (q) τ s τ

234

8 Electron–Phonon Interaction k– q

>

>

k+q

>

> sq

>

>

sq

k

k

Fig. 8.1. Graphical representation of the electron–phonon interaction: phonon emission (left) and phonon absorption (right)

k+q

k+q k

k

G

k+q–G normal

umklapp

Fig. 8.2. Kinematics of the electron–phonon interaction: normal process (left) and Umklapp process (right), the thin solid lines mark the boundary of the first Brillouin zone

which depends on the coupling mechanism as will be outlined below. The operator part of Hel−ph tells us about the kinematics as determined by the wave vectors. For this, we illustrate the interaction in graphical form (Fig. 8.1) and the kinematics in k space (Fig. 8.2). The graphs contain the fermion operators as straight lines and the boson operator as a wavy line. Hel−ph contains only a single phonon operator: a phonon is created or annihilated, and two electron operators: a creation and an annihilation operator. In the interaction process, the electron is scattered between two Bloch states while a phonon is emitted or absorbed (see Fig. 8.1). The total momentum is conserved in the scattering (the sum of the wave vectors of the creation operators equals that of the annihilation operators, up to a reciprocal lattice vector). We note in passing that (8.8), although derived here for the electron–phonon system, is the standard form of a fermion–boson interaction. It applies as well to the coupling of electrons with photons or magnons and also to similar problems in nuclear physics. Two characteristic scenarios are shown in Fig. 8.2 to demonstrate the kinematics of electron–phonon interacion. If the wave vectors of the electron before and after the scattering are within the 1st Brillouin zone and the momentum transfer is small (k and k + q are almost parallel), the scattering does not strongly change the direction of the moving electron. This is called the normal process. If the same small momentum transfer shifts the wave vector across

8.2 Coupling Mechanisms

235

the Brillouin zone boundary, we have to bring it back by subtracting a reciprocal lattice vector, which almost inverts the direction of the moving electron (k and k + q − G are almost antiparallel). This is called the Umklapp process. It is intuitively clear that the Umklapp processes take a stronger influence on the transport properties of electrons than the normal processes. Formally, the energy balance of the scattering process will be considered by treating the electron–phonon interaction as a time-dependent perturbation, which leads to self-energy corrections changing the energy of the electrons and giving them a finite lifetime. But, it can be made clear from the graphs: the total energy is conserved and the energy of an emitted (absorbed) phonon has to be provided (is carried away) by the electron. The diﬀerent coupling mechanisms, which depend on the lattice properties of the solid and exhibit characteristic dependencies on q, will be presented in the following Section.

8.2 Coupling Mechanisms In Chap. 3, we have demonstrated the physical properties of phonons in different branches. The long-wavelength acoustic phonons were recognized as causing local lattice compression or dilation, while long-wavelength optical phonons have been identified with electrical dipole vibrations. These pictures are helpful in deriving the corresponding mechanisms of electron–phonon coupling. Deformation Potential Coupling (Acoustic Phonons): The local homogeneous compression or dilation ∆(r) caused by acoustic phonons in the long-wavelength limit can be described as a local relative volume change (Fig. 8.3) ∆(r) =

∆V "" " . V r

(8.10)

It is experienced by an electron as a local change of the lattice constant which shifts the single particle energy. If we consider a simple nondegenerate band

V ∆V r Fig. 8.3. Sketch of the homogeneous dilation caused by a longitudinal acoustic phonon

236

8 Electron–Phonon Interaction

(deriving from atomic s states as e.g., the conduction band in a normal metal or semiconductor), the electron–phonon interaction can be written as a local potential seen by the electrons Hel−ph = − D ∆(r l ), (8.11) l

where D is the deformation potential corresponding to the energy shift for a relative volume change ∆V /V = 1. Deformation potentials are of the order of a few eV (see the data collection of Landolt–B¨ ornstein [1]). The relative volume change caused by phonons, being much smaller than 1, is related with the flux of the continuous lattice displacement field u(r) through the area enclosing the volume V as depicted in Fig. 8.3. By Gauss’ theorem, we have ∇ · u dV (8.12) ∆V = u · dA = V

which for a homogeneous lattice distortion in the volume V (assumed to have a small linear extension compared with the wavelength of the phonon) can be written also as V ∇ · u(r). Note that ∇ · u(r) can be expressed as the trace of the strain tensor field, Trǫ(r) (Problem 8.2). The displacement field u(r) is obtained in the limit of long wavelengths from the expression for the displacement unτ (3.23) by replacing R0n → r and using the appropriate expression for the eigenvector esτ (q). As the electron–phonon interaction is determined by ∇·u(r), only the longitudinal acoustic phonons with eLA τ (q) = # Mτ /M q/q contribute, where M is the total mass of the ions in the Wigner– Seitz cell. Thus, we can write Hel−ph = −D

l

∇l · u(r l ) = −iD

QLAq q · q √ eiq·rl q N M l,q

(8.13)

and obtain with QLAq =

9

h ¯ 2ωLA (q)

a†LA (−q) + aLA (q)

and (8.6) the Hamiltonian for the deformation potential coupling † D Hel−ph = Vq aLA (−q) + aLA (q) c†k+q ck .

(8.14)

(8.15)

k,q

The coupling matrix element reads Vq = −iD

%

h ¯ q 1/2 2N M cLA

where the use was made of the dispersion relation ωLA (q) = cLA q.

(8.16)

8.2 Coupling Mechanisms

237

Supplement: Deformation potential coupling of electrons in a p band In deriving (8.15), we have assumed the electrons to be in a simple s band. For energy bands originating from atomic orbitals with higher angular momentum, we have to consider the orbital degeneracy, according to which the Hamiltonian for the electron–phonon interaction becomes a matrix in the Bloch (or angular momentum) representation. Here, we face the same problem as in the eﬀective-mass approximation for the p type valence band in semiconductors (see Sect. 5.5), which resulted in a 3 × 3 matrix Hamiltonian with bilinear expressions in the components of the wave vector as matrix elements and three-material specific parameters (L, M, N ), which define the curvature of the bands in diﬀerent directions in k space. This matrix can be expressed also in terms of angular momentum matrices for I = 1, which in the basis |x , |y , |z read ⎛ ⎞ ⎛ ⎞ ⎛ ⎞ 00 0 00 i 0 −i 0 (8.17) Ix = ⎝ 0 0 −i ⎠, Iy = ⎝ 0 0 0 ⎠, Iz = ⎝ i 0 0 ⎠. 0 i 0 −i 0 0 0 00 As it turns out, the k · p matrix M (5.104) can be decomposed according to ! 1 2 Iα2 − I 2 kα {Iα , Iβ }kα kβ , (8.18) − 2N M = Ak2 1 + B 3 α α<β

where in the first term k2 = kx2 + ky2 + kz2 and 1 is a 3 × 3 unit matrix (which is proportional to Ix2 + Iy2 + Iz2 ) and {Iα , Iβ } = (Iα Iβ + Iβ Iα )/2, while the constants are related by A = (L + 2M )/3 and B = −L + M . The individual terms of M are invariant tensor products under the point group and can be formulated on group theoretical grounds [235, 236]. The symmetry properties of the symmetric second-rank strain tensor with the components ǫαβ are the same as those of the tensor formed by kα kβ . Therefore, in the same degeneracy space (here the 3-fold space of the p states), the Hamiltonian of the deformation potential coupling has the form ! 1 2 2 ǫαα + 2D3 I HD {Iα , Iβ }ǫαβ . (8.19) = D Trǫ + D − I 1 2 el−ph α 3 α α,β It consists of three invariant contributions connected with three deformation potentials. D1 is recognized as the deformation potential of hydrostatic strain (as in (8.11)), while D2 and D3 correspond to shear strain in (001) and (111) direction, respectively. This strain Hamiltonian, originally derived for homogeneous static strain, is used here to describe the deformation potential coupling. The components of the strain tensor field ǫαβ (r), which are symmetrized derivatives of the displacement field (see (3.76)), can be expressed in terms of phonon operators as before and one obtains the matrix Hamiltonian for the deformation potential coupling of electrons in a p band. The complexity of this interaction Hamiltonian allows not only for the coupling to be longitudinal but also to transverse phonons (Problem 8.3). The same concept can be applied to energy bands with other orbital degeneracies and to coupling between diﬀerent bands.

238

8 Electron–Phonon Interaction

Polar Coupling with Optical Phonons or Fr¨ ohlich1 Coupling [237, 238]: In the long-wavelength limit, longitudinal optical phonons cause a macroscopic polarization field M (r) P (r) = (8.20) V with the electric dipole field M (r). A charge density ρ(r) placed in this polarization field gives rise to the interaction energy Eint = − P (r) · E(r)d3 r (8.21) with E(r) = −∇r

ρ(r′ ) d3 r ′ , 4πε0 ε∞ |r − r′ |

(8.22)

where ε∞ is the background (or high frequency) dielectric constant. By partial integration, this can also be written as ρ(r ′ ) 3 (8.23) Eint = − d r ∇r · P (r) d3 r′ 4πε0 ε∞ |r − r ′ |

with the charge distribution given by e iq′ ·r′ † e iq′ ·(r′ −rl ) e =− e ck−q′ ck . ρ(r′ ) = − V V ′ ′ l,q

q

(8.24)

k

The lattice displacement for longitudinal optical phonons in the longwavelength limit reads in operator form

uLO nτ

=

q

%

¯ h † iq·R0n a eLO (q)e (−q) + a (q) . LO LO 2N Mτ ωLO τ

(8.25)

Here we assume a crystal with two ions in the Wigner–Seitz cell with masses M1 , M2 and M = M1 + M2 for which the phonon eigenvectors are (see Problem 3.1) % % M2 q M1 q LO LO e1 (q) = , e2 (q) = − . (8.26) M q M q The moving ions, carrying the dynamical eﬀective charge η1 = −η2 = η (introduced in Sect. 3.5), provide a macroscopic dipole moment 9 h ¯ q iq·R0n † e aLO (−q) + aLO (q) (8.27) ητ un,τ = η M= 2N μωLO q n,τ n,q 1

Herbert Fr¨ ohlich, 1905 – 1991

8.2 Coupling Mechanisms

239

where μ is the reduced ion mass. The continuum limit M (r) is obtained by replacing the discrete lattice vectors R0n by the local position vector r. Thus, we can write 9 N q · q iq·r † h ¯ ∇ · P (r) = iη e aLO (−q) + aLO (q) . (8.28) V q 2N μωLO q The two volume integrals in (8.23) are performed by taking the Fourier transform of 1/|r − r ′ | (see Appendix) to arrive at ′ ′ 1 4π 1 = 2 δq,−q′ . d3 r d3 r′ eiq·r eiq ·r (8.29) ′ V |r − r | q Finally, we take the dynamical eﬀective charge from (3.115) and use the Lyddane–Sachs–Teller relation (3.114) to replace the transverse optical phonon frequency by the longitudinal one η=

μV ε0 1 1 − N ε∞ ε(0)

!1/2

ε∞ ωLO

(8.30)

to write Eint as the electron–phonon interaction for the Fr¨ ohlich coupling † F = Hel−ph (8.31) Vq aLO (−q) + aLO (q) c†k+q ck k,q

with the interaction matrix element !1/2 2 hωLO 1 1 e ¯ 1 . − Vq = i 2V ε0 ε∞ ε(0) q

(8.32)

The strength of the interaction is frequently quantified by the Fr¨ ohlich coupling constant e2 α = 8πε0 ¯ hωLO F

2m∗ ωLO h ¯

!1/2

! 1 1 , − ε∞ ε(0)

(8.33)

with the eﬀective electron mass m∗ . It is defined in analogy with the fine structure constant, which is the corresponding coupling constant for the electron–photon interaction and shall find an obvious meaning when dealing with the polaron in Sect. 8.4. Depending on the polarity, determined by the diﬀerence between ε∞ and ε(0) of the solid, on the eﬀective mass and on the energy of the longitudinal optical phonon, αF takes values between 0.07 in GaAs, 0.39 in CdTe, and 6.6 for RbBr (compared with the fine structure constant of 1/137). For the weak coupling regime (αF < 1) the Fr¨ ohlich interaction can be treated by perturbation theory (see Sect. 8.4), while special concepts have been developed for the strong coupling regime (αF > 1)[239].

240

8 Electron–Phonon Interaction

Piezoelectric Coupling (Acoustic Phonons): In crystals without inversion symmetry, a homogeneous strain causes a dielectric polarization P , known as the piezoelectric eﬀect. It is quantified by the relation (double index summation understood) Pi = eijk ǫjk ,

(8.34)

where eijk is the piezoelectric and ǫjk the strain tensor. Due to symmetry considerations, a third rank tensor has nonvanishing elements only for crystals lacking inversion symmetry [74, 235]. For the particular case of zinc blende, the piezoelectric tensor takes the form eijk = e14 |εijk | with the Levi–Civita symbol εijk and the piezoelectric constant e14 (written in Voigt notation, see Sect. 3.4). The piezoelectric eﬀect, originally related to static strain, applies as well to the dynamical case of strain fields connected with acoustic phonons and gives rise to the piezoelectric electron–phonon coupling. For this case, the classical interaction energy is to be formulated with the polarization field P (r) according to (8.34) with the strain field ǫjk (r) =

i (qk uj + qj uk ) eiq·r 2

(8.35)

connected with the displacement field u(r) = u exp(iq · r). For acoustic phonons, u can be written as % % N M s s s u= Qsq e (q), withe (q) = e (q). (8.36) M s Mτ τ τ Taking the corresponding displacement field P (r), the piezoelectric electron– phonon interaction can be derived in analogy with the Fr¨ ohlich coupling (Problem 8.4). The result is P Hel−ph = (8.37) Vsq a†s (−q) + as (q) c†k+q ck s,k,q

with the coupling matrix element % hN 1 ¯ 2ee14 Vsq = − ε0 ε∞ 2M vs q 1/2

qx qy esz (q) + c.p. q2

!

,

(8.38)

where vs is the sound velocity of the phonon branch s.

8.3 Scattering Processes: Lifetime, Relaxation The electron–phonon interaction represents a link between the electron and phonon systems, which in the previous chapters have been investigated separately. In this Section, we consider it as a perturbation of the electron system by evaluating the scattering processes depicted in Fig. 8.1: phonon emission

8.3 Scattering Processes: Lifetime, Relaxation

241

and phonon absorption by a single electron. According to the time-dependence of the phonon operators, the electron–phonon interaction is periodic in time and Fermi’s Golden Rule applies. In general, the scattering rate between Bloch electron states with energies ǫnk and ǫn′ k′ under a perturbation H ′ is rnk,n′ k′ =

2π 2 |n′ k′ |H ′ |nk| δ(ǫnk − ǫn′ k′ − ∆ǫ), h ¯

(8.39)

where ∆ǫ is the energy change in the case of inelastic scattering. Considering scattering-out from the Bloch state with n, k to all other possible Bloch states n′ , k′ , one finds the inverse single-particle or carrier lifetime 1 τnk

=

n′ ,k′

rnk,n′ k′ (1 − f (n′ k′ )) ,

(8.40)

where f (n′ k′ ) is the distribution function. It vanishes if we consider a single electron in an otherwise empty band. In the language of Green functions, this lifetime is related with the imaginary part of a self-energy contribution h ¯ /τnk , which represents a level broadening. Taking now Hel−ph as the perturbation H ′ , we have to evaluate the matrix element with the electron and phonon states in the occupation number representation. This can formally be done, but the result is written immediately by inspection of the graphs for phonon absorption and emission processes (see Fig. 8.1). In the one-band approximation adopted in Sect. 8.2, we can drop the band index and know from momentum conservation that k′ = k + q for normal processes, to which we can restrict ourselves here. The matrix elements of the phonon operators yield # ns (q) − 1|as (q)|ns (q) = ns (q) for phonon absorption (8.41) # † ns (q) + 1|as (q)|ns (q) = ns (q) + 1 for phonon emission (8.42)

with the phonon occupation numbers ns (q), and the inverse lifetime is expressed by 1 2π 2 |Vsq | (1 − f (k + q)) = τk ¯ s,q h ! 1 1 δ(ǫk − ǫk+q ∓ ¯hωs (q)). × ns (q) + ∓ 2 2

(8.43)

Here, the upper(lower) sign refers to phonon absorption(emission). Phonon absorption is possible only if the occupation factor ns (q) diﬀers from zero. Energy conservation in the scattering process (depicted in Fig. 8.4) is expressed by the δ-function. For conduction electrons in semiconductors with ǫk ≫ hωs (q), the scattering with acoustic phonons is almost elastic, while scatter¯ ing with optical phonons is connected with a substantial change in energy.

242

8 Electron–Phonon Interaction

>

Ck

>

>>

hωLO hωLO

k Fig. 8.4. Scattering of conduction electrons in a semiconductor with acoustic and optical phonons

The emission processes describe energy dissipation from the electrons to the lattice, which serves as a heat sink. The rate of energy transfer between the electron and phonon systems is given for this process by an expression similar to (8.44) but with an additional factor h ¯ ωs (q)/ǫk under the sum. For high energy transfer rates, the phonon system will be heated up and one has to consider phenomena related with hot phonons. The general result of (8.44) can be specified for the diﬀerent mechanisms of the electron–phonon interaction with their particular matrix elements, which diﬀer with respect to their dependence on the wave vector q of the emitted or absorbed phonon: 2

|VLAq | ∼ q 2

|VLOq | ∼ q 2

−2

|Vsq | ∼ q −1

(deformation potential coupling)

(8.44)

(Fr¨ ohlich coupling)

(8.45)

(piezoelectric coupling).

(8.46)

This has a consequence when evaluating the sum over q as an integral with the thermal phonon and electron occupations. This is done by substitutions (see Appendix) which lead to diﬀerent dependencies of these scattering mechanisms, their contribution to electron lifetime and energy dissipation, on the temperature. The dynamical processes connected with electron–phonon scattering can be studied for two diﬀerent scenarios. The first applies to semiconductors and insulators, where electrons in the conduction band can be created by optical excitation from the valence band. These electrons thermalize due to carrier– carrier interaction on a very short time scale and form a hot carrier system (with a temperature higher than the lattice temperature), which relaxes due to phonon emission (energy relaxation) and finally equilibrates with the lattice before the electrons and holes recombine [240–242]. These processes can be investigated in time-resolved spectroscopy [95, 243].

8.3 Scattering Processes: Lifetime, Relaxation

243

The second scenario is that of transport, which applies as well to metals as to doped semiconductors. The ensemble of carriers, which in an external electric field are accelerated, is described by a nonequilibrium distribution function f (k, T ). The carriers dissipate their excess energy by emitting phonons. A stationary situation is achieved if the rate of energy gain of the carriers in the electric field equals the rate of energy dissipation by phonon emission. In a homogeneous system, this situation is accounted for by the Boltzmann equation (or Boltzmann’s stationarity condition) [64, 244–246] ! e ∂f (k, T ) − E · ∇k f (k, T ) = , (8.47) h ¯ ∂t coll where the lhs accounts for the rate of energy gain by the carriers in the field E. The rhs of this relation, the scattering or collision term, can be formulated in terms of the single-particle scattering rates rk,k′ and occupation factors ! ( ∂f (k, T ) = (1 − f (k, T )) f (k′ , T )rk,k′ ∂t coll ′ k ) (8.48) −f (k, T ) (1 − f (k′ , T )) rk′ k ,

where the first term under the sum on the rhs represents the scattering processes into the state with k while the second term represents scattering-out from this state. For isotropic scattering rates, rk,k′ = rk′ ,k , we find ! ∂f (k, T ) = (8.49) rk,k′ (f (k′ , T ) − f (k, T )) . ∂t coll ′ k

Instead of these microscopic expressions, the collision term is frequently treated in the relaxation time approximation ! ∂f (k, T ) f (k, T ) − f0 (k, T ) , (8.50) =− ∂t τtr (k) coll which describes the evolution of the nonequilibrium distribution f (k, T ) into the equilibrium distribution f0 (k, T ) in the characteristic transport relaxation time τtr (k). Equations (8.47) and (8.50) can be solved by iteration e τtr (k)E · ∇k f (k, T ) ¯ h e = f0 (k, T ) + τtr (k)E · ∇k f0 (k, T ) + . . . . h ¯

f (k, T ) = f0 (k, T ) +

(8.51)

For small deviations from the equilibrium distribution, it is suﬃcient to take into account only the lowest order correction (linear in E), which defines the regime of linear transport, while the higher order terms describe the nonlinear transport [247].

244

8 Electron–Phonon Interaction

The physical observable to quantify the carrier transport, the electric current density, can be expressed as e j=− v(k)f (k, T ). (8.52) V k

Here, the carrier velocity v(k) is the group velocity of electrons in an energy band ǫk 1 (8.53) v(k) = ∇k ǫk , h ¯ which for a simple parabolic band equals h ¯ k/m∗ with the eﬀective mass m∗ . Consider now one component of the current density with the nonequilibrium distribution function from (8.51) ⎧ ⎫ ⎨ e eτtr (k) ∂f0 (k, T ) ⎬ jα = − vα (k) f0 (k, T ) + Eβ . (8.54) ⎩ ⎭ V h ¯ ∂kβ β

k

The first term on the rhs vanishes because the system in equilibrium does not carry a current. By writing the distribution function as f (ǫk , T ), the derivative with respect to kβ in the second term becomes ∂f0 (k, T ) ∂f0 (ǫ, T ) ∂ǫk ∂f0 (ǫ, T ) ¯hvβ (k), = = ∂kβ ∂ǫ ∂kβ ∂ǫ

(8.55)

and we find Ohm’s law in the form jα =

σαβ Eβ

(8.56)

e2 ∂f0 (ǫ, T ) vα (k)vβ (k). τtr (k) V ∂ǫ

(8.57)

β

with the electric conductivity σαβ = −

k

This second rank tensor resembles the form derived in Chap. 2 within the concepts of linear response (see Problem 2.4), and we may recognize the correlation between the two components of the velocity. On the other hand, here the appearance of the transport relaxation time is a new aspect. It accounts for the dissipation of energy in collisions, which is essential for obtaining a finite conductivity. Comparing (8.49) and (8.50) we may identify f (k, T ) − f0 (k, T ) = rk,k′ (f (k, T ) − f (k′ , T )) . τtr (k) ′

(8.58)

k

and see that in general the transport relaxation time depends on the distribution function, which limits the validity of the relaxation time approximation.

8.3 Scattering Processes: Lifetime, Relaxation

245

However, this is not the case for an isotropic dispersion ǫk = ǫk and elastic scattering for which we may write f (k) = f0 (k) + k · EC(k)

(8.59)

and similar for f (k′ ) with the electric field E and a scalar function C(k) = C(k ′ ). In polar coordinates with respect to the direction of E, one has k · E = kE cos ϑ

and k · k′ = k 2 cos ϑ′ ,

(8.60)

where ϑ′ is the angle between k and k′ , and k′ · E = kE(cos ϑ cos ϑ′ + sin ϑ sin ϑ′ cos ϕ′ ). As a consequence, the diﬀerence f (k) − f (k′ ) can be expressed as f (k) − f (k′ ) = kEC(k) cos ϑ(1 − cos ϑ′ ) − sin ϑ sin ϑ′ cos ϕ′

(8.61)

(8.62)

where the second term on the rhs vanishes by integration over ϕ′ (note that for isotropic scattering rk,k′ depends on k − k′ ) which gives with dΩ′ = dϕ′ sin ϑ′ dϑ′ dΩ′ f (k) − f (k′ ) = kEC(k) cos ϑ dΩ′ (1 − cos ϑ′ ) . (8.63) With (8.59), we identify the factor in front of the integral on the rhs as f (k) − f0 (k) and obtain, using (8.39), for the relaxation rate 1 1 = 2 |k|H ′ |k′ |2 δ(ǫk − ǫk′ )(1 − cos ϑ′ )d3 k ′ (8.64) τtr (k) 4π ¯ h

which is independent of the distribution function. The factor 1 − cos ϑ′ under the integral tells us that collisions by which the propagation direction of the charge carrier is reversed (ϑ′ ≃ π, back scattering) contribute much to the relaxation rate, while the eﬀect of forward scattering with ϑ′ ≪ π is only small. For further evaluation of σαβ (8.57), let us assume the electric field in the z direction. Then, we have σzz as the only nonvanishing component of the conductivity tensor. The sum over k can be performed in spherical polar coordinates. By making use of the dispersion relation ǫk = h ¯ 2 k 2 /2m∗ , we can write with vz (k) = h ¯ kz /m∗ and kz = k cos θ vz2 (k) =

2 ǫ cos2 θ m∗

(8.65)

and formulate the remaining integration over k as an energy integral σzz

e2 =− ∗ 2 3m π

2m∗ h2 ¯

!3/2

ǫ3/2 τtr (ǫ)

∂f0 (ǫ, T ) dǫ . ∂ǫ

(8.66)

246

8 Electron–Phonon Interaction

The transport relaxation time is a power function of the carrier energy ǫ. Integrals of this type can be evaluated by using the fact that ∂f0 (ǫ, T )/∂ǫ is a symmetric function with a pronounced maximum at ǫ = μ (see Appendix). Depending on the temperature, two limiting cases shall be discussed here. For a degenerate electron system with EF ≫ kB T , the derivative of the Fermi distribution function can be written as ∂f0 /∂ǫ ≃ −δ(ǫ − EF ) thus giving σzz =

e2 nτ (EF ) m∗

(8.67)

with the degenerate carrier density n=

2m∗ h2 ¯

1 3π 2

!3/2

3/2

EF .

(8.68)

Formally, this result is identical with that obtained by the classical Drude theory [12, 29, 31]. However, here, the quantum statistical nature of the degenerate electron system becomes relevant in the dependence on the Fermi energy and the transport relaxation time can be traced back to the microscopic scattering processes. For the nondegenerate carrier system, kB T ≫ ǫ, the derivative of the distribution function takes the form ∂f0 /∂ǫ ≃ − exp (−ǫ/kB T )/kB T and we may write σzz =

e2 nτ (ǫ) m∗

with the nondegenerate carrier density !3/2 2m∗ 1 dǫ ǫ1/2 e−ǫ/kB T n= 2π 2 h2 ¯ and the averaged transport relaxation time 3 dǫ ǫ3/2 τ (ǫ)e−ǫ/kB T 2 3 . τ (ǫ) = 3kB T dǫ ǫ1/2 e−ǫ/kB T

(8.69)

(8.70)

(8.71)

The eﬀects of the band structure (here the eﬀective mass) and of the scattering processes are comprised in the mobility μ=

eτ m∗

(8.72)

with the temperature-dependent transport relaxation time τ . Thus, the conductivity is written in the form σ = enμ. The mobility is frequently measured in magneto-transport experiments making use of the Hall eﬀect and then called Hall mobility. An example is shown in Fig. 8.5 for n-doped GaAs. The measured values are compared with the calculated mobilities for diﬀerent scattering mechanisms (Problem 8.5). Besides scattering with phonons, which

8.4 The Fr¨ ohlich Polaron

Half mobility mH [cm2V–1s–1]

108

GaAs

247

neutral impurity

107 ionized impurity

106

piezoelectric deformation potential

105

polar

104 103 1

10

103 102 Temperature T [K]

104

Fig. 8.5. Hall mobility versus temperature for n-doped GaAs, showing experimental data together with the contributions from diﬀerent scattering mechanisms, after [89]

limits the mobility at high temperatures, scattering also with impurities (see Chap. 9) is considered. The latter dominates the mobility at low temperature, when the phonons are frozen out. For degenerate electron systems, the relaxation time can be converted into a mean free path lmfp = vF τ , which is the average distance between collisions covered by a carrier with Fermi velocity vF . Note that the momentum selection rule represents a geometrical constraint depending on the system dimension. Thus, compared with the threedimensional case, electron–phonon interaction is modified in electron systems with reduced dimensionality (e.g., for electrons confined at semiconductor heterostructures or in quantum wells or quantum wires). Moreover, by remote doping, the ionized impurities can be separated from the mobile carriers to suppress the scattering. This leads to an enormous increase of the mobility at low temperatures. Thus, carrier confinement takes influence on the lifetime [242, 243, 248]. At the same time, the mean-free-path, the distance between successive scattering events, can become comparable or larger than the system dimension, as is typical for mesoscopic systems. In this case, the carrier passes through the sample without being scattered and the transport is called ballistic.

8.4 The Fr¨ ohlich Polaron In contrast with the single-particle approximation in Chaps. 4 and 5, where a free electron or a Bloch state was an eigenstate of the electron Hamiltonian, these states are now perturbed by the electron–phonon interaction. We

248

8 Electron–Phonon Interaction

encounter a similar situation as for interacting electrons: As the electron– electron interaction gives rise to formation of quasi-particles consisting of the bare particle and a cloud of virtual collective excitations, so does the electron– phonon interaction. But now, instead of charge or spin density excitations, the electron (or hole) will be dressed by a cloud of virtual phonons. This composite particle is called polaron with particular reference to the lattice deformation in a polar semiconductor or insulator by a charge carrier, which repels the ions with the same charge but attracts those with the opposite charge thus generating a distortion typical for optical phonons. The interaction to be considered in this case is the Fr¨ ohlich coupling. We follow here [4, 14] and adopt the concept of lowest order perturbation theory to calculate the change of the single-electron state |k and its energy by the coupling with longitudinal optical phonons. We consider the state |k, 0q as the eigenstate of Hel + Hph characterizing an electron in the undistorted harmonic lattice. Due to electron–phonon interaction, this state gets modified by contributions from states with phonons. The lowest order correction is due to one-phonon contributions |k, 0q (1) = |k, 0q +

q

|k − q, 1q

k − q, 1q |Hel−ph |k, 0q . ǫk − ǫk−q − ¯hωLO

(8.73)

The lattice distortion caused by the carrier can be quantified by the expectation value of the phonon number operator which due to the first order correction is |k − q, 1q |Hel−ph|k, 0q |2 n ¯= . (8.74) (ǫk − ǫk−q − ¯ hωLO )2 q This expression is still not specified to one of the coupling mechanisms and yields diﬀerent results for the sum over q according to the q dependence of the interaction matrix element. For the Fr¨ ohlich interaction, we have h ¯ ωq → ¯hωLO and may write with ǫk − ǫk−q − ¯ hωLO =

¯2 h 2 2k · q − q 2 − qLO , ∗ 2m

(8.75)

2 where qLO = 2m∗ ωLO /¯ h, the sum over q as an integral over the Brillouin zone !2 V 1 2m∗ 1 2 d3 q, (8.76) n ¯ = |C| 2 3 2 2 )2 2π q h ¯ (2k · q − q 2 − qLO

where |C|2 = e2 ¯hωLO /2V ε0 (1/ε∞ − 1/ε(0)). Let us consider the state with k = 0, representing an electron at rest. Then the integral can be performed in spherical polar coordinates over a sphere with radius qBZ equal to the volume of the Brillouin zone !2 qBZ dq 2m∗ 3 . . . d q = 4π . (8.77) 2 2 + q 2 )2 (q h ¯ 0 LO

8.4 The Fr¨ ohlich Polaron

The indefinite integral gives dq 1 2 )2 = 2q (q 2 + qLO LO

q q 1 arctan + 2 q 2 + qLO qLO qLO

!

.

249

(8.78)

For qLO ≪ qBZ , because the wave vector of an electron at the energy of the LO phonon is much smaller than the Brillouin zone, the upper limit 3 can be extended to infinity and the integral simplifies to π/4qLO . Taking the interaction constant C from (8.33), we find n ¯=

αF , 2

(8.79)

i.e., in this lowest order perturbation, the composite particle polaron consists, of the electron and a number of phonons which is determined by the coupling constant. A similar calculation can be performed also for the other coupling mechanisms (Problem 8.6). For polar semiconductors, this number is much smaller than one, which a posteriori justifies the perturbation treatment. In systems with much larger coupling constants, more elaborate concepts have to be applied [239]. The first order perturbation correction to the free particle energy ǫk = h2 k 2 /2m∗ is given by ¯ (1)

∆ǫk =

|k − q, 1q |Hel−ph |k, 0q |2 q

(ǫk − ǫk−q − ¯hωLO )

.

As before, the sum over q can be written as an integral to have ∗ 1 1 V (1) 2 2m ∆ǫk = |C| d3 q. 2 2π 3 2 2k · q − q 2 − q 2 q h ¯ LO

(8.80)

(8.81)

Being interested now in the dispersion relation for the composite particle, we expand the integrand for small k and find for the leading terms !2 1 2k · q −1 2k · q 1+ 2 = 2 + + ... . 2 2 2 2 2k · q − q 2 − qLO q + qLO q + qLO q 2 + qLO (8.82) When integrating in spherical polar coordinates, the second term on the rhs vanishes while the first and third terms can be evaluated with the integrals dq q 1 arctan and (8.83) = 2 q 2 + qLO qLO qLO ! 2 ) q (q 2 − qLO 1 1 q q2 d q . (8.84) = + arctan 2 2 )2 2 )3 8qLO (q 2 + qLO qLO qLO (q 2 + qLO

250

8 Electron–Phonon Interaction k electron polaron

αF hωLO

k

Fig. 8.6. Dispersion of conduction band electron and Fr¨ ohlich polaron. In order to show energy and mass renormalization a coupling constant of about 1 has been assumed

By taking the limits from 0 to ∞ as before, these integrals reduce to π/2qLO 3 and π/16qLO , respectively, and give for the first order energy correction (1)

hωLO − ∆ǫk = −αF ¯

αF h ¯ 2 k2 , 6 2m∗

(8.85)

which again is linear in the coupling constant. It is conceivable that higher order perturbation theory would lead to terms with higher powers in the coupling constant, i.e., the results obtained here are valid only for αF < 1. The polaron energy dispersion can now be written as ! αF h2 k 2 ¯ P F ǫk = −α ¯ . (8.86) 1− hωLO + 2m∗ 6 The result is shown in Fig. 8.6. There is an energy reduction independent of k and a change of the particle mass, similar to our finding for the Hartree–Fock quasi-particle in Chap. 4. The mass m∗ /(1 − αF /6) ≃ m∗ (1 + αF /6) is the polaron mass. The intuitive physical picture of the polaron is that the carrier has to move around together with the lattice polarization created due to its charge. The coupling constant αF defines the energy renormalization in units of h ¯ ωLO and the average number of virtual phonons in the polaron.

8.5 Eﬀective Electron–Electron Interaction In 1911, the Dutch physicist Kamerlingh Onnes2 discovered the surprising behavior of Hg that below a critical temperature (Tc = 4.2 K) the resistivity dropped suddenly to a value close to zero (see Fig. 8.7), i.e., below this temperature the material can carry an electric current without a voltage drop or dissipation of energy [249]. This behavior, called superconductivity, which was found in the following for many other metals, has become and still is one 2

Heike Kamerlingh Onnes 1853–1926, Nobel prize in physics 1916

8.5 Eﬀective Electron–Electron Interaction

251

ρ(T) normal supra

Tc

T

Fig. 8.7. Schematic dependence of the resistivity on the temperature for a normal metal (dashed) and a superconductor (solid)

of the most challenging problems in solid state physics. Superconductivity has become a standard subject in textbooks on Solid State Physics, but is also well documented in special monographs [250–254]. An indication that the lattices (or precisely the ions) are involved in the mechanism causing superconductivity, was the discovery of the isotope eﬀect, i.e., a dependence of the critical temperature on the ion mass, according to which the critical temperature depends for a given element on the ion mass such that M 1/2 Tc = const. But, it took almost half a century before in 1957 a microscopic explanation of the phenomenon was given by Bardeen, Cooper, and Schrieﬀer3 [255]. According to their theory, known as the BCS theory, electron–phonon coupling can mediate an attractive electron–electron interaction which below Tc gives rise to a new correlated ground state with paired electrons. In the perspective of promising technical applications, eﬀorts have been made in finding materials with higher critical temperatures which, however, until 1985 remained below 25 K. In 1986, Bednorz and M¨ uller4 discovered a new class of superconducting materials, the doped ceramic cuprates, with significantly higher critical temperatures [256]. Within this material class, critical temperatures of up to about 120 K, well above the temperature of liquid nitrogen, were found in the following years. This discovery boosted the research in the field of superconductivity towards both superconducting devices and in theoretical concepts to describe the eﬀect, which is still not completely understood [257]. One of the most recent Nobel prizes in physics

3

4

John Bardeen, 1908–1991, Leon N. Cooper, *1930, J. Robert Schrieﬀer, *1931, shared the Nobel prize in physics 1972 Johannes Georg Bednorz, *1950, Karl Alex M¨ uller, *1927, shared the Nobel prize in physics 1987

252

8 Electron–Phonon Interaction

k′–q

>

>

k+q

> >

>

q k

k′

Fig. 8.8. Graphical representation of the eﬀective electron–electron interaction mediated by virtual phonon exchange

has been awarded to Abrikosov, Ginzburg, and Leggett5 in recognition of their contributions to this field. In the context of this chapter, the BCS theory although only the first concept to explain superconductivity - will be presented as an important outcome of electron–phonon interaction. The system of electrons and phonons, including their interaction, can be described in simplest form by the Hamiltonian introduced in Sect. 8.1 H = Hel + Hph + Hel−ph = ǫk c†k ck + hω(q)a† (q)a(q) ¯ k

q

+

Vq c†k+q ck

k,q

a† (−q) + a(q) .

(8.87)

Here, we assume the electrons in a simple energy band with dispersion ǫk and the phonons from a branch with dispersion ω(q). The matrix element of the electron–phonon interaction depends only on the momentum transfer q and ∗ obeys the relation Vq = V−q . Considering Hel−ph as a small perturbation, one can try to eliminate this interaction at least to lowest order in the phonon operators. As will be shown in the following, this leads to an eﬀective electron– electron interaction mediated by virtual exchange of phonons. Its pictorial representation in Fig. 8.8, which can be understood as a composition of the diagrams shown in Fig. 8.1, resembles the corresponding graph for the direct Coulomb interaction (see Fig. 4.10). The elimination of the terms linear in the phonon operators a† (−q), a(q) can be achieved by a unitary transformation, the Schrieﬀer–Wolﬀ transformation ! ! 1 1 H ′ = e−S HeS = 1 − S + S 2 . . . H 1 + S + S 2 . . . 2 2 1 (8.88) = H + [H, S] + [[H, S], S] + . . . , 2 5

Alexei A. Abrikosov, *1928, Vitaly L. Ginzburg, *1916, Anthony J. Leggett, *1938, shared the Nobel Prize in physics 2003

8.5 Eﬀective Electron–Electron Interaction

253

with a still to be determined operator S. This transformation represents a systematic perturbation expansion. Applied to the Hamiltonian (8.87) the leading terms of the series read H′ = H0 + Hel−ph + [H0 , S] + [Hel−ph , S] +

1 [[H0 , S], S] + . . . , 2

(8.89)

where H0 = Hel + Hph is the dominating term and all other terms, containing diﬀerent powers of Hel−ph or S, represent the perturbation. This expansion can be exploited to eliminate Hel−ph (in principal to any order) by a proper choice of the operator S. For the lowest order, let us assume the form S= (8.90) Vq α a† (−q) + β a(q) c†k+q ck k,q

which is similar to Hel−ph but contains α, β as free parameters. The commutator of S with H0 yields the two contributions (8.91) (ǫk+q − ǫk )Vq α a† (−q) + β a(q) c†k+q ck [Hel , S] = k,q

and [Hph , S] =

k,q

hω(q)Vq α a† (−q) + β a(q) c†k+q ck , ¯

(8.92)

where use was made of ω(−q) = ω(q). We recognize that the second and third terms of the perturbation expansion (8.89) are of first order in the electron– phonon interaction, while the fourth and fifth terms are of second order. The condition Hel−ph + [H0 , S] = 0 for eliminating the first order terms can be fulfilled with the choice α = − (ǫk+q − ǫk − ¯ hω(q))

−1

(8.93)

β = − (ǫk+q − ǫk + h ¯ ω(q))−1

(8.94)

and

and the transformed Hamiltonian takes the form H′ = H0 + [Hel−ph , S] +

1 [[H0 , S], S] + . . . . 2

(8.95)

For this choice of S, the sum of the second and third term is [Hel−ph , S]/2 with 2 [Hel−ph , S] ≃ (8.96) |Vq | (−α + β)c†k+q c†k′ −q ck′ ck . k,k′ ,q

254

8 Electron–Phonon Interaction

Note that the detailed evaluation (Problem 8.7) of the commutator yields an additional term which is neglected here. Finally, we may write the transformed Hamiltonian by considering all terms up to second order in the electron– phonon coupling as H′ = H0 +

k,k′ ,q

2

|Vq | ¯ hω(q) 2

(ǫk − ǫk+q ) −

c†k+q c†k′ −q ck′ ck . 2 (¯hω(q))

(8.97)

The second term of this Hamiltonian describes an electron–electron interaction caused by a virtual exchange of phonons as depicted by the diagram of Fig. 8.8. It has the same structure as the graph for the Coulomb interaction, but now the interaction line represents the phonon mechanism. It is attractive for |ǫk − ǫk+q | < ¯ hω(q) and favors the formation of bound electron pairs. Starting from the filled Fermi sphere |Ψ0 , which represents the electronic ground state of a normal metal, we may ask if an electron pair formed at the Fermi energy due to the eﬀective electron–electron interaction would be stable. Such a pair can be described by applying two creation operators with k1 , k2 from outside the Fermi sphere and superposition to a wave packet |Ψ12 =

′

k1 ,k2

F (k1 , k2 )c†k1 c†k2 |Ψ0 ,

(8.98)

′ indicates that the sum is restricted to ki , i = 1, 2 in the spherical where shell with EF ≤ ǫki ≤ EF + h ¯ ω(q), i = 1, 2. The center of mass momentum, K = k1 + k2 , being a good quantum number, limits the states contributing to the wave packet, as depicted in Fig. 8.9. This figure suggests that the condition of forming a stable pair improves by increasing the number of contributing states with attractive interaction. The most favorite situation corresponds to K = k1 + k2 = 0, for which we can write |Ψ12 =

′ k

F (k)c†k c†−k |Ψ0 .

k1 k 2 K

(8.99)

EF +hωq EF

Fig. 8.9. Single particle states at the Fermi energy contributing to a pair state for fixed center-of-mass momentum

8.6 Cooper Pairs and the Gap

If, in addition, the attractive interaction is simplified by writing V 2 |Vq | ¯ hω(q) |ǫk − ǫk′ | ≤ ¯hωD − 2eff δk′ ,−k ≃ , 2 2 0 otherwise (ǫk − ǫk′ ) − (¯hω(q))

255

(8.100)

where the Debye frequency ωD introduced in Chap. 3 is taken as a representative cut-oﬀ frequency of the phonon spectrum, we obtain the Hamiltonian Heﬀ =

k

ǫk c†k ck −

Veﬀ ′ † ck+q c†−k−q c−k ck 2

(8.101)

k,q

for electrons in the band with ǫk coupled with the phonon-mediated attractive electron–electron interaction.

8.6 Cooper Pairs and the Gap As mentioned before, this interaction can give rise to the formation of bound electron pairs |Ψ12 at the Fermi surface, the Cooper pairs. The condition for this to happen is given by E12 = Ψ12 |Heﬀ |Ψ12 < 2EF ,

(8.102)

which means an instability of the Fermi sphere against the formation of such pairs. The expectation value of (8.101) for the pair state (8.99) turns out to be (Problem 8.8) E12 = 2

′ k

ǫk |F (k)|2 − Veﬀ

′

F ∗ (k + q)F (k)

(8.103)

k,q

and is to be minimized with respect to F (k) under the normalization condition 2 k |F (k)| = 1. Thus, we write + , ′ ∂ 2 E12 − λ |F (k)| − 1 =0 (8.104) ∂F ∗ (k′ ) k

or alternatively, after taking the derivatives, {2ǫk′ − λ} F (k′ ) = Veﬀ By setting

′

k

′

F (k).

(8.105)

k

F (k) = C, we may express F (k) =

Veﬀ C 2ǫk − λ

or

C=

where in the last equation C can be dropped.

′ Veﬀ C , 2ǫk − λ k

(8.106)

256

8 Electron–Phonon Interaction

An expression for λ is obtained by multiplying (8.105) with F ∗ (k′ ) and taking the restricted sum over k′ ′ k′

{2ǫk′ − λ} |F (k′ )|2 = Veﬀ

′

F (k)F ∗ (k′ ).

(8.107)

k,k′

By making use on the lhs of the normalization condition and replacing the rhs k′ by k + q with the corresponding change of the summation, this can be written as ′ ′ 2 ǫk′ |F (k′ )|2 − Veﬀ F ∗ (k + q)F (k) = λ. (8.108) k′

k,q

Comparing with (8.103), we identify λ = E12 and find the relation ′ k

Veﬀ =1 2ǫk − E12

(8.109)

from which the energy E12 of the Cooper pair is to be calculated. In (8.109), the sum over k can be formulated as an integral over the energy 1 = Veﬀ

EF +¯ hωD

EF

D(E) dE, 2E − E12

(8.110)

with the density of states D(E). Because the integral is to be taken over a narrow interval at the Fermi energy (¯ hωD ≪ EF ), we may extract the density of states as a factor D(EF ) and perform the integration to find 1=

¯ ωD ) − E12 1 2(EF + h Veﬀ D(EF )ln . 2 2EF − E12

(8.111)

After separating the logarithm to take the exponential, this can be solved for E12 and yields E12 = 2EF − 2¯hωD

exp (−2/Veﬀ D(EF )) . 1 − exp (−2/Veﬀ D(EF ))

(8.112)

Under the condition of a weak attractive electron–electron interaction, quantified by the relation 2/Veﬀ D(EF ) ≫ 1, (8.112) is simplified by writing E12 ≃ 2EF − 2¯hωD exp (−2/Veﬀ D(EF )).

(8.113)

This relation implies that the energy of the Cooper pair is smaller than the energy of the two electrons (without interaction) taken from the Fermi surface. The particular dependence on the interaction in the exponential function, which cannot be represented in a power series, leads to stable bound pairs even for very small Veﬀ . The formation of Cooper pairs means, for the whole electron system, that the Fermi sphere (or in general, the Fermi surface) is

8.6 Cooper Pairs and the Gap

257

unstable and cannot represent the system ground state. As we shall see, this leads to a new state of the electron system, the superconducting state. After having confirmed the possibility of pair formation due to the eﬀective electron–electron interaction, we now have to find the ground state of the system. The pairs, consisting of electrons close to EF with wave vectors k and −k and (as we shall assume here) up and down spin to form spin-singlet pairs, are bosons. Their creation or annihilation changes the number of electrons in the system. This can be accounted for by considering the so-called BCS Hamiltonian HBCS =

′ k

′ † † (ǫk − EF ) c†k↑ ck↑ + c†−k↓ c−k↓ − Veﬀ ck′ ↑ c−k′ ↓ c−k↓ ck↑ , k,k′

(8.114) † † which is Heﬀ − μN , with μ ≃ EF and N = k ck↑ ck↑ + c−k↓ c−k↓ . As the energy of the electrons does not diﬀer much from the Fermi energy, it is reasonable to introduce the notation E(k) = ǫk − EF . Subtracting this energy diﬀerence summed over all k leads to ′ ′ † † ¯ =− H E(k) 1 − c†k↑ ck↑ − c†−k↓ c−k↓ − Veﬀ ck′ ↑ c−k′ ↓ c−k↓ ck↑ . k

k,k′

(8.115) The operators appearing in this Hamiltonian have an obvious meaning when applied to pair states of electrons. Using the occupation number representation, |0k↑ 0−k↓ and |1k↑ 1−k↓ denote the states without and with an electron pair, respectively. The following relations hold: 1. 1 − c†k↑ ck↑ − c†−k↓ c−k↓ |0k↑ 0−k↓ = |0k↑ 0−k↓ 2. 1 − c†k↑ ck↑ − c†−k↓ c−k↓ |1k↑ 1−k↓ = −|1k↑ 1−k↓ 3. c†k↑ c†−k↓ |1k↑ 1−k↓ = 0

4. c†k↑ c†−k↓ |0k↑ 0−k↓ = |1k↑ 1−k↓ 5. ck↑ c−k↓ |1k↑ 1−k↓ = |0k↑ 0−k↓ 6. ck↑ c−k↓ |0k↑ 0−k↓ = 0

For each k, they can be interpreted as those of spin operators which in the Pauli representation read ! 10 1./2. : 1 − c†k↑ ck↑ − c†−k↓ c−k↓ = = σkz 0 −1 k ! 1 00 = (σkx − iσky ) 3./4. : c†k↑ c†−k↓ = 10 k 2 ! 1 01 (8.116) = (σkx + iσky ) 5./6. : ck↑ c−k↓ = 00 k 2

258

8 Electron–Phonon Interaction

and act on corresponding pseudo-spin states ! 1 |0k↑ 0−k↓ = and |1k↑ 1−k↓ = 0 k

0 1

!

.

(8.117)

k

¯ can be written as In this notation, the Hamiltonian H ¯=− H

′ k

E(k)σkz −

Veﬀ ′ (σk′ x σkx + σk′ y σky ) 4 ′

(8.118)

k,k

and turns out to be of the same structure as the Heisenberg spin Hamiltonian in Chap. 6. Therefore, the interaction (second term) can be treated in the same way, i.e., by applying the mean-field approximation. This is done by introducing the fictitious magnetic field (note that it depends on k and its components represent energies) giving H k = E(k)ˆ z+

Veﬀ ′ σk′ x ˆ x + σk′ y ˆ y , 2 ′

(8.119)

k

with Cartesian unit vectors x ˆ, yˆ, zˆ, and leads to the compact form ′ ¯=− H H k · σk .

(8.120)

k

In spite of its simple form, this mean-field Hamiltonian has a remarkable spectrum, determined by the modulus of the fictitious field. The energetically best arrangement requires all pseudo-spins σ k to be aligned to the corresponding field H k . Let us discuss first the case without interaction (Veﬀ = 0)

IHkI

Veff =0 EF

Ek

z

Fig. 8.10. Mean field (upper part) and eigenstates (lower part) in dependence on the single-particle energy around EF without eﬀective interaction

8.6 Cooper Pairs and the Gap

259

IHkI ∆ Veff ≠0 Ek

EF z

occupied

empty

Fig. 8.11. Mean field (upper part) and eigenstates (lower part) in dependence on the single-particle energy around EF with eﬀective interaction. Note the opening of a gap at EF

for which |H k | depends linearly on the energy E(k), which is the energy difference of the single particle energy and the Fermi energy. This situation is shown in Fig. 8.10 (upper part). The field points in the positive or negative z direction depending on the sign of E(k). The corresponding eigenstates are pseudo-spin up (down) or occupied (empty) pair states below (above) EF , i.e., for negative (positive) E(k) (see lower part of Fig. 8.10). This picture changes for finite Veﬀ (Fig. 8.11). Now, at EF or E(k) = 0 the fictitious field is finite, pointing somewhere in the xy plane. Its strength depends on the expectation values of the pseudo-spin operators σk′ x and σk′ y . But, away from the Fermi energy, this field is turned into the z direction and approaches the linear dependence as for the noninteracting case (upper part of Fig. 8.11). The corresponding eigenstates are pseudo-spin vectors aligned to this field and turn around with H k . Their projection onto the z direction is shown in the lower part of Fig. 8.11. In order to quantify the discussion, let us assume σk′ y = 0, i.e., all pseudo-spins together with the fictitious ′ fields are in the xz plane, and introduce the energy parameter ∆ = Veﬀ k′ σk′ x /2. The energy for exciting an electron pair out of the correlated ground state is given by twice the modulus of the field |H k | = {E(k)2 + ∆2 }1/2 ,

(8.121)

where ∆ is the gap energy at EF (see Fig. 8.11). Due to the alignment of the pseudo-spin along the field H k (Fig. 8.12), the ratios of their x and z components are equal and can be expressed by the angle θk . As, in addition, σk′ x = sin θk′ we may write

260

8 Electron–Phonon Interaction

z Hk <σ>

θk

k

x Fig. 8.12. Alignment of pseudo-spin and fictitious field in the xz plane

tan θk =

Veﬀ

′

∆ k′ sin θk′ = . 2E(k) E(k)

(8.122)

This relation can be solved for ∆ by replacing sin θk′ = to obtain ∆=

∆ (∆2

+ E 2 (k′ ))

1/2

′ ∆ 1 Veﬀ . 2 (∆2 + E 2 (k))1/2

(8.123)

(8.124)

k

The summation, limited to the range of the attractive interaction around the Fermi surface, can again be performed as an integral, giving Veﬀ D(EF ) +¯hωD dE 1= 1/2 2 2 −¯ hωD (∆ + E 2 (k)) ! hω D ¯ . (8.125) = Veﬀ D(EF ) sinh−1 ∆ This is easily solved to yield, as the BCS solution, for the gap parameter ∆=

¯ ωD h ≃ 2¯hωD e−1/Veff D(EF ) , sinh(1/Veﬀ D(EF ))

(8.126)

where the last expression is obtained for 1 ≫ Veﬀ D(EF ). According to this formula, the gap energy (or the stability of the Cooper pairs) increases with the cut-oﬀ phonon energy (or the Debye frequency) and the density of single-particle states at the Fermi energy. In order to find an expression for the critical temperature, we have to ask for the temperature dependence of the gap, which enters through the thermal expectation values of the pseudo-spins ! |H k | 1 −β H¯ σk T = Tr e (8.127) σk = tanh Z kB T

8.6 Cooper Pairs and the Gap

and (using again Fig. 8.12) defines ∆(T ) by the relation ′ Veﬀ k′ sin θk′ |H k′ | ∆(T ) tan θk = tanh = . 2E(k) kB T E(k)

261

(8.128)

Replacing sin θk′ = ∆(T )/{E 2 (k) + ∆2 (T )}1/2 gives Veﬀ ′ |H k | 2 {E (k) + ∆2 (T )}−1/2 = 1. tanh 2 kB T

(8.129)

k

This equation can be solved for ∆(T ). For T = Tc the gap vanishes and |H k | = E(k), which simplifies the relation to Veﬀ ′ 1 E(k) tanh = 1. 2 E(k) kB Tc

(8.130)

k

Extending this result, which is valid for the adopted spin model related to pair states, to include single particle excitations, Tc has to be replaced by 2Tc . The factor 2 accounts for the doubling of the entropy due to a doubling of the possible excitations (see [4]). The sum over k is again performed as an energy integral, giving the BCS result for the critical temperature h¯ ωD /2kB Tc 1 tanh(x) = dx. (8.131) Veﬀ D(EF ) x 0 For Tc ≪ ΘD , where ΘD is the Debye temperature, this can be evaluated to yield Tc = 1.14

ΘD −1/Veff D(EF ) e . kB

(8.132)

This result immediately explains the isotope eﬀect, because the Debye temperature is inversely proportional to the square root of the ion mass. With the help of (8.126), we may also write the direct relation between the gap energy and the critical temperature 2∆ = 3.52kB Tc .

(8.133)

The validity of this relation is demonstrated in Table 8.1 for some representative metals. The principal correctness of the mean-field approximation is shown in Fig. 8.13 by plotting the reduced value of the gap ∆(T )/∆(0) as the order parameter for diﬀerent superconductors versus the reduced temperature T /Tc . The experimental data for diﬀerent materials define a universal curve which is well described by the graph from BCS theory. Note the similarity with Fig. 6.10. Similar to the Fr¨ ohlich polaron, a Cooper pair is a composite particle. However, the statistical properties of the polaron are those of the electron

262

8 Electron–Phonon Interaction

Table 8.1. Debye temperature ΘD , critical temperature Tc , and ratio 2∆(0)/kB Tc for diﬀerent metals, from [31]. Metal An Cd Hg Al Sn Pb

ΘD [K]

Tc [K]

2∆(0)/kB Tc

235 164 70 375 195 96

0.9 0.56 4.16 1.2 3.75 7.22

3.2 3.2 4.6 3.4 3.5 4.3

1.0

∆(T) / ∆(0)

0.8 0.6 0.4 0.2 0

0

0.2

0.4 0.6 T/Tc

0.8

1.0

Fig. 8.13. Dependence of the reduced gap energy on the reduced temperature for diﬀerent superconductors (symbols) and the BCS result (solid line) (after [30])

because the phonon cloud does not change the spin, whereas the Cooper pair, which is a particle with integer spin composed of two electrons, is a boson. Due to their bosonic nature, a system of Cooper pairs can undergo Bose– Einstein condensation to reach the superconducting state. Depending on the total spin of the two electrons in a Cooper pair, one distinguishes singlet and triplet superconductors. Bose–Einstein condensation is possible not only for paired electrons in a solid. By cooling atoms confined in electromagnetic traps, degenerate Fermi gases can be realized (see Chap. 4), which consist of atoms. For these Fermi systems, one has, recently, also studied the formation of molecules due to inter-atomic forces. A system of such bosonic molecules can make a phase transformation into the suprafluid state corresponding to the superconducting state of the electrons [258, 259].

Problems

263

Problems 8.1 Derive the expression for the electron–phonon interaction in a Bravais lattice without making use of the single-band approximation. In this case the operator form should be † as (−q) + as (q) c†nk cn′ k′ . (8.134) 8.2

8.3

8.4

8.5

8.6

8.7 8.8

Derive an expression for the interaction matrix element, which leads to a relation between k, k′ and q. Express the divergence of the continuous lattice displacement field u(r) by the components of the strain tensor field ǫ(r) and realize that only longitudinal phonons contribute. Show that charge carriers in an energy band deriving from p states can interact via deformation potential coupling with both longitudinal and transverse phonons. Derive the expression for the piezoelectric electron–phonon coupling (8.37) by performing the same steps as outlined for the polar optical electron– phonon coupling. Which acoustic branches contribute to the coupling? Given the scattering rates 1/τ of an electron with energy ǫ due to deformation potential (∼T ǫ1/2 ) and piezoelectric coupling (∼T ǫ−1/2 ) for large kB T compared with the phonon energies, calculate the temperature dependence of the contributions of both scattering processes to the mobility of electrons. Compare with Fig. 8.5. Calculate the number of phonons n ¯ in a composite particle due to deformation potential coupling with acoustic phonons by using first order perturbation theory. Note that the q dependence of this coupling diﬀers from that of the Fr¨ ohlich coupling and that the q integration changes. Discuss the result in dependence on the material parameters. Estimate n ¯ if the carrier would be an ion (e.g., a proton) moving in the lattice. Evaluate the commutator between the operators Hel−ph and S. Discuss the approximation made in writing the expression given in the text. Calculate the ground state energy of the BCS Hamiltonian (8.120) relative to the normal ground state. Note that only electrons close to EF are relevant.

9 Defects, Disorder, and Localization

The crystalline order of a solid, characterized by the configuration {R0n } of all atoms with their equilibrium positions at lattice points, has been anticipated throughout this book so far. The essential consequence of the lattice periodicity, deriving from this configuration, is that all states be it of phonons, of electrons, or of magnons follow Bloch’s theorem with a modulus that repeats with the lattice period, e.g., |ψnk (r + R0n )| = |ψnk (r)| ,

for all R0n .

(9.1)

Because of this property, Bloch states are extended states. In this chapter, deviations from the crystalline order and their consequences shall be considered for the electrons (but they are essentially valid also for the collective excitations of the solid). This is done first for individual point defects or impurities, which have to be classified, and can lead to discrete bound states localized in the vicinity of the defect. They are described by wave functions that decay (exponentially) away from the defect. A more general deviation from the crystalline structure is compositional or structural disorder, which will be experienced by the electrons as a random potential. The energy spectrum of the electrons can still exhibit bands with continuous density of states but with a modification of the electronic wave functions depending on their energy: Toward the band edges, the electrons are more strongly influenced by the potential fluctuations and get confined in deep local minima. Such localized states have the characteristic form ψ(r) ∼ e−|r|/λ , (9.2) where λ is the localization length. Thus, disorder causes localization and becomes responsible for dramatic changes in the transport properties. They show up at low temperatures at which – as one of the prominent features – a metal–insulator transition (MIT) can take place due to disorder. The experimental investigations of the MIT and the conceptual theoretical work carried out to understand this phenomenon belong to the outstanding achievements made in solid state physics during the last decades.[275–279] U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 9,

266

9 Defects, Disorder, and Localization

Fig. 9.1. Representation of point defects in a binary solid (open and filled symbols represent two diﬀerent atomic species)

9.1 Point Defects Deviations from the periodic structure can be classified according to their dimension. Point defects represent zero-dimensional perturbations of the periodic lattice structure. Diﬀerent cases are depicted in Fig. 9.1 for a solid containing two sublattices with atoms A and B, respectively. In a substitutional impurity CA (or CB ), an atom C diﬀering from those of the host material A (or B) replaces the host atom at a lattice site. An interstitial impurity I means an atom (of the host material, but also diﬀerent from it) taking a position in some free space between the lattice points. A vacancy VA (or VB ) refers to an empty lattice point. A Frenkel1 defect, VA − IA , is the combination of a vacancy with a nearby interstitial. It can be created by displacing a host atom from its lattice site to the interstitial position. In binary solids AB, especially in compound semiconductors, an anti-site defect AB is possible with an atom A at the regular site of a B atom. This anti-site defect is more likely, the closer A and B are in the periodic table. In semiconductors, isolated impurities can change the electronic spectrum by causing states in the fundamental energy gap from which, at elevated temperature carriers are released into the nearby band edges. This is the concept of doping which is of great importance in device applications. Therefore, the theory of impurities within the eﬀective-mass approximation originates from the early days of semiconductor physics [260], but was refined and extended following the progress of research [261–268]. For an isolated substitutional impurity in an otherwise crystalline solid, we may write 1

Yakov Ilyich Frenkel 1894–1954

9.1 Point Defects

267

¯2 h ∆ + Veﬀ (r) + U (r) (9.3) 2m with the periodic eﬀective potential Veﬀ (r) = n v(r − R0n ) introduced in Sect. 5.1 and the impurity potential U (r) = vI (r) − v(r), where the lattice site of the substitution is taken as the origin. Without U (r), the eigenvalue problem of H leads to band structure En (k) and Bloch functions ψnk (r) which represent the solution of the unperturbed problem. Including the impurity potential, the eigenvalue problem H=−

Hφ = Eφ (9.4) can be solved by expanding φ(r) = nk fn (k)ψnk (r) as a wave packet of Bloch waves. The variational principle for the expectation value of the energy leads to the set of coupled linear equations En (k) − E fn (k) + (9.5) nk|U |n′ k′ fn′ (k′ ) = 0 n′ ,k′

for the expansion coeﬃcients fn (k). Using the Fourier transform U (r) = q Uq exp(iq ·r) and expanding the product of the periodic parts of the Bloch functions in a Fourier series, u∗nk (r)un′ k′ (r) = G Cnkn′ k′ (G) exp(−iG · r), the matrix element of the impurity potential U (r) can be expressed as nk|U |n′ k′ = Uq (9.6) Cnkn′ k′ (G)δq,k−k′ +G . q

G

For G = 0, one has Cnkn′ k′ (0) =

u∗nk (r)un′ k′ (r) ≃ δnn′ d3 r,

(which is exact for k = k′ ) and writes En (k) − E fn (k) + Uk−k′ fn (k′ )

(9.7)

k′

+

n′ k′

Uk−k′ +G Cnkn′ k′ (G)fn′ (k′ ) = 0.

(9.8)

G=0

The last term in this equation, describing the interband coupling due to the impurity potential, can be neglected if for G = 0, the relation |Uk−k′ +G | Cnkn′ k′ (G) ≪ 1 |Uk−k′ |

(9.9)

holds, which leads to 3 the one-band approximation. Taking the Fourier transform with fn (k) = exp(−ik · r)fn (r)d3 r/V and by writing En (k →

¯2 h 1 ∇) ≃ En (0) − ∆, i 2m∗

(9.10)

268

9 Defects, Disorder, and Localization

we find the eﬀective-mass equation for shallow impurities in a simple band [260] h2 ¯ − ∆ + U (r) fn (r) = (E − En (0))fn (r). (9.11) 2m∗ The condition (9.9) is fulfilled for |k − k′ | ≪ |G| and applies if the expansion coeﬃcients fn (k) diﬀer from zero only for small k in the vicinity of the band edge. This is the case for charged impurities as, e.g., a substitutional Si (or Ge) at a Ga site in GaAs, for which the impurity potential is essentially a screened Coulomb potential with a short-range correction term Ucc (central-cell correction) to account for the chemical nature of the impurity U (r) = Ucc (r) −

e2 . 4πε0 εr

(9.12)

energy

This leads to a modified hydrogen problem with bound states below the band minimum. The characteristic units of energy and length, the eﬀective Rydberg constant Ry∗ = (m∗ /mε2 )Ry, and the Bohr radius a∗B = (εm/m∗ )aB , respectively, scale according to the material parameters from the atomic values. For GaAs, one has Ry∗ ≃ 10−3 Ry ≃ 5 meV and a∗B ≃ 100 aB ≃ 10 nm. The bound states for hydrogen-like impurities can be classified by the angular momentum quantum numbers. Only the 1s state is essentially modified by the central-cell correction. Figure 9.2 draws attention to these bound states at the band edge in k and in real space. In k space, the 1s wave function is represented by its Fourier transform f1s (k) which extends over a width of the inverse eﬀective Bohr radius around the band minimum. In real space, it is a wave function f1s (r) extending over a distance of about the eﬀective Bohr radius around the impurity site. The real space localization of the state increases with the strength of the impurity potential. Shallow acceptor states have a more complex structure, which derives from the eﬀective-mass Hamiltonian of p-like valence bands [261]. In contrast to (hydrogen-like) shallow impurities, whose spectrum is dominated by the long-range Coulomb interaction with only minor modifications

conduction band

f1s(k)

E1s

< 1 < a*B

k

f1s(r) < a*B <

r

Fig. 9.2. Impurity states at a band edge in k space (left) and in real space (right)

9.1 Point Defects

269

due to the central-cell correction (especially for the ground state), the situation is completely reversed for deep impurities. Here, the energy spectrum is determined by the short-range central-cell potential, and the influence of the long-range Coulomb potential (if present at all) is considered as a correction. Consequently, the deep impurity states are strongly localized to the neighborhood of the impurity site accompanied by lattice distortions. This situation cannot be well described by a superposition of extended Bloch states. Here it is more appropriate to use atomic orbitals or concepts of scattering theory with a localized basis. Starting from the eigenvalue problem of (9.3) with the impurity potential U (r) now dominated by the central-cell correction, we use the Ansatz for the wave function ψ(r) = cα (R)φα (r − R) (9.13) α,R

with atomic orbitals φα (r − R) localized at R. While in Sect. 5.4, when introducing the LCAO method, we have constructed Bloch functions out of the atomic orbitals, we now remain in the localized representation. The variational principle for the energy leads to the set of coupled linear equations (9.14) (HαR,α′ R′ − E SαR,α′ R′ )cα′ (R′ ) = 0 α′ ,R′

for the expansion coeﬃcients cα (R) and energy eigenvalues following from the secular problem HαR,α′ R′ − E SαR,α′ R′ = 0. (9.15)

Here HαR,α′ R′ and SαR,α′ R′ are the matrix elements of the Hamiltonian and of the overlap between atomic orbitals (see Sect. 5.4), respectively. Some aspects of deep impurities, in particular the chemical trends, can be understood in this model as exemplified here for an isoelectronic impurity, e.g., GaP:N (meaning that P in GaP is substituted by N) [261, 269]. In the picture of atomic orbitals, the valence and conduction bands of intrinsic semiconductors with tetrahedral coordination are formed by the bonding and anti-bonding states, respectively, of the s and p orbitals. This is sketched for the host atoms in the left hand side of Fig. 9.3, where the shaded regions indicate the continua of the energy bands. Replacing one host atom P by the impurity N leads to diﬀerent pairs of s and p bonding and anti-bonding states localized at the impurity site (right hand side of Fig. 9.3): in particular, the s-bonding state becomes a resonant impurity level deep in the valence band, while the s-anti-bonding state of the Ga-N pair is lowered with respect to that of the Ga-P pair of the host crystal, thus forming a deep trap in the energy gap. Corresponding impurity states (not shown in the figure) are formed out of the p orbitals. Chemical trends can be discussed by simulating a continuous change of the impurity (X) with a change of its atomic level energy EX . For the bonding

270

9 Defects, Disorder, and Localization Ga

Ep

antibonding

E Ga p

E pp

N

Ep

deep trap

Ga

Ga

Es

Es

p

bonding

Es

hyper – deep E N s trap

Fig. 9.3. Schematic level diagram with valence and conduction band formed from bonding and anti-bonding states of host atoms Ga and P (left) and of deep impurity states formed with s states of a substitutional N impurity (right)

and anti-bonding states of the s orbitals of Ga and of the impurity X, one may write the simplified secular problem " " " EGa − E " U " "=0 (9.16) " U EX − E "

with EGa(X) being the energy of the atomic s orbital of the Ga(X) atom, and U the coupling between these orbitals. The eigenvalues of (9.16) E± =

EGa − EX EGa + EX ± 2 2

1+

4|U |2 (EGa − EX )2

1/2

(9.17)

are depicted as a function of EGa − EX in Fig. 9.4. Depending on the sign of EGa − EX , the anti-bonding or bonding states are shown: they evolve from the valence or conduction band edge, respectively, and converge with increasing |EGa − EX | toward the energy EGa of the atomic Ga-s orbital which acts as a pinning level. The symmetry of these impurity states is A1 due to the s orbital [269]. Similar considerations for the p orbitals lead to deep impurity states with T2 symmetry. The other concept to describe deep impurities is based on scattering theory and employs the Green function of the impurity problem [270] G(E) = lim

δ→0

1 . E − H + iδ

(9.18)

We make use of the separation of the system Hamiltonian H = H0 +U into the Hamiltonian H0 for the unperturbed periodic solid and the impurity potential U (r), now understood as the diﬀerence of the self-consistent DFT-LDA singleparticle potentials with and without impurity (see Sect. 5.1) 0 [n(r), r]. U (r) = Veﬀ [n(r), r] − Veﬀ

(9.19)

9.1 Point Defects

271

Fig. 9.4. Chemical shift of deep traps evolving from p-bonding and s-anti-bonding states

The Green function G(E) of the full problem can be expanded in the Born series G(E) = G0 (E) + G0 (E)U G0 (E) + . . . = (1 − G0 (E)U )−1 G0 (E).

(9.20)

The first factor in the last expression gives rise to additional poles of G(E) caused by the impurity potential U . They exist besides those of the second factor, the Green function G0 (E) of the unperturbed band structure. For the density of states 2 (9.21) D(E) = − Tr Im G(E) π one finds with G(E) = G0 (E) + (G(E) − G0 (E)) D(E) = − π2 Tr Im G0 (E) − π2 Tr Im (G(E) − G0 (E))

(9.22)

= D0 (E) + ∆D(E) where D0 (E) is the density of states of the unperturbed solid and ∆D(E) its change due to the impurity both in the gap and in the continuum of the valence and conduction bands. In a representation with localized functions, e.g., Wannier functions 1 ik·R e ψnk (r) φn (r − R) = √ N k

(9.23)

with

φ∗n (r − R)φn (r − R′ )d3 r = δnn′ δR,R′

(9.24)

272

9 Defects, Disorder, and Localization

the matrix elements (take for simplicity n = n′ ) φ∗n (r − R)U (r)φn (r − R′ )d3 r URR′ =

(9.25)

are diﬀerent from zero only for a small number of R, R′ out of a set of lattice points {RI } around the impurity site. This defines a short range defect matrix ! U{RI } 0 (9.26) (URR′ ) = 0 0 for which the additional poles of G(E) in (9.20) can be calculated from (Problem 9.1) 1 − G0 (E)U = 1 − G0,{RI } (E)U{RI } = 0.

(9.27)

Here G0,{RI } (E) is part of the matrix representation of the Green function connected with the set {RI }. A point defect of particular interest is the anti-site defect, which has been studied intensively in GaAs and found to be responsible for realizing semi-insulating material [271]. Anti-site defects exist also in other compound semiconductors. The schematic level diagram of the AsGa anti-site defect is shown in Fig. 9.5. Besides resonances, deriving from p-anti-bonding and bonding states in the valence and conduction bands, there is a trap in the middle of the gap connected with the s-anti-bonding states of A1 symmetry. Note the degeneracy of the respective states, it is sixfold (with spin) for T2 deriving from p-states and twofold for A1 deriving from s-states. Transition metal atoms in semiconductors form deep impurity states as a consequence of their tightly bound d electrons.[272] Their orbital multiplets are split by the crystal field into states characterized by the point symmetry (e.g., in zinc blende material A1 , A2 , E, T1 and T2 , see Appendix). VGa

AsGa

As Ec

T2

p Ev

A1 s

Fig. 9.5. Level diagram of the AsGa anti-site defect (center) and its composition out of the Ga vacancy (left) and p states of the As atom (right). Full (open) dots indicate occupied (empty) states

9.2 Disorder

273

9.2 Disorder Following the systematics of dimensionality in the deviations from the crystalline order, we mention briefly the dislocations as one-dimensional defects and may address interfaces (or surfaces) as two-dimensional defects. A particular deviation from crystalline configurations is a disorder by composition or structure, which can exist in all three spatial dimensions. Compositional disorder results, e.g., by placing two or more diﬀerent kinds of atoms statistically on the otherwise unperturbed lattice points. This is the case in so-called mixed crystals or alloys, in which the long-range order is destroyed because there is no correlation in the chemical nature of the ions occupying the lattice points. Also statistical occupation of the lattice points with diﬀerent isotopes of one kind of atoms creates compositional disorder. We mentioned already the broadening of phonon resonances due to isotope disorder in Chap. 3. A more dramatic three-dimensional disorder is the structural disorder, typical for amorphous solids, which do not exhibit any crystalline order. Adopting a single-particle description in the sense of Chap. 5, we write the electron Hamiltonian for a system with disorder as H=−

¯2 h ∆+ vj (r − Rj ) 2m

(9.28)

Rj

with the ion configuration {Rj } and the eﬀective single-particle potentials vj (r − Rj ) provided by the specific atom (or ion) at Rj . Disorder can be accounted for by deviations from the crystalline ion configuration {R0n } and/or by varying the atomic species with the potential vj from site to site. We may write the Hamiltonian H also in the form † H= ǫ j cj cj + tjl c†j cl (9.29) j

j,l

by using fermion operators c†j , cj in the site representation. Here the singleparticle energies ǫj (the site energies) account for disorder in the occupation of the sites Rj , while the transfer matrix elements tjl do it for the hopping between sites Rj and Rl . The actual configuration of a disordered material is not known but is not essential for the physical properties. Instead, disorder is considered by means of probability distributions of the site energies ǫj and/or of the transfer matrix elements tjl . For obvious reasons, the former is called diagonal disorder, the latter oﬀ-diagonal disorder. Both are related to probability distributions of the configurations P ({Rj }). A simple case of diagonal disorder is that the single-particle energies scatter around a mean value ǫ0 , and we may write † ǫ 0 cj cj + tjl c†j cl + (ǫj − ǫ0 )c†j cj , (9.30) H= j

j,l

j

274

9 Defects, Disorder, and Localization

where the first two terms are the tight-binding Hamiltonian for a single band (see Sect. 5.4), while the last term represents the diagonal disorder. We may switch to the Bloch representation and write H= ǫk c†k ck + u(q)c†k+q ck = H0 + Hdisorder (9.31) k

k,q

which in a natural way splits into the unperturbed energy band and the perturbation due to disorder. The potential matrix element u(q) is the Fourier transform of the impurity potential U (r − Rj ) and can be separated according to u(q) = j exp (iq · Rj )U (q) = ρ(q)U (q) into structure factor ρ(q) and form factor U (q). For a systematic treatment of the disorder, we study the single-particle Green function (see Sect. 7.1) i G(k, k′ , t − t′ ) = − θ(t − t′ ){ck (t), c†k′ (t′ )}. h ¯

(9.32)

The equation of motion for G(k, k′ , t − t′ ) is easily obtained with H from (9.31) and reads

i¯ h

∂ − ǫk G(k, k′ , t − t′ ) = δkk′ δ(t − t′ ) ∂t −i + u(k − k′′ ) θ(t − t′ ){ck′′ (t), c†k′ (t′ )} . ¯h k′′ G(k′′ ,k′ ,t−t′ )

(9.33)

¯ ω + iδ, the After Fourier transformation with respect to t − t′ with E + = h retarded Green function of the disorder problem is given by 1 δkk′ + G(k, k′ ; E + ) = + u(k − k′′ )G(k′′ , k′ ; E + ) . (9.34) E − ǫk ′′ k

By identifying the first factor on the rhs as the unperturbed retarded Green function G0 (k, E + ) for the Bloch band, this can be written in the form of the Dyson equation

G(k, k′ ; E + ) = G0 (k, E + ) δkk′ +

k′′

u(k − k′′ )G(k ′′ , k′ ; E + ) . (9.35)

By iteration, it leads to the Born series G = G0 + G0 uG0 + G0 uG0 uG0 + . . . and by introducing the t-matrix

(9.36)

9.2 Disorder

t = u (1 + G0 u + G0 uG0 u + . . . ) = u

∞

n=0

it takes the form

G0 u

n

G(k, k′ ; E + ) = G0 (k; E + ) δk,k′ + t(k, k′ ; E + )G0 (k′ ; E + ) .

275

(9.37)

(9.38)

We note that the full Green function depends on the configuration {Rj } of the impurities, which in principle is not known. On the other hand, we do not expect, that the physical properties of a suﬃciently large sample are determined by its specific impurity configuration. Since on macroscopic length scales (large compared e.g. with the mean free path) diﬀerent configurations are realized, a statistical average can be taken which depends only on the impurity concentration. However, this argument does not apply, in the mesoscopic regime with sample sizes comparable to the typical transport lengths. In this regime, the observation of universal conductance fluctuations can be taken as a fingerprint of the distinct impurity configuration realized in the sample [159, 160]. Here we rely on the property of a macroscopic observable A, that it is self-averaging, i.e., for the variance VarA = (A − A)2 it holds that VarA = 0, V →∞ A2 lim

(9.39)

where V is the volume of the system. An example can be treated as Problem 9.2. The averaging over impurity configurations aﬀects only the structure factor and is written as f (q) = ρ(q)conf =

{Rj }

P ({Rj })

N

eiq·Rj ,

(9.40)

j=1

where P ({Rj }) is the probability distribution for the configurations of N impurities. The t-matrix consists of terms with diﬀerent powers of the potential u(q). Thus, the configuration average has to be performed in all orders of ρ(q), and we write for the nth order f (q 1 , . . . , q n ) = ρ(q 1 ) . . . ρ(q n ).

(9.41)

There is no preference for the impurities to be on particular sites, and we can simplify the double sum over configurations and sites by replacing N

{Rj } j=1

··· → c

...

(9.42)

all sites

with the impurity concentration c. When taking the sum over all sites and not only over those occupied by an impurity, we have n exp (iq · Rn ) = N δq,0 and find

276

9 Defects, Disorder, and Localization

f (q 1 , . . . , qn ) = pn (c)δq1 ,0 . . . δqn ,0

(9.43)

with a polynomial pn (c) of degree n in the impurity concentration. The lowest order terms can explicitly be written iq1 ·Rn = N δq,0 f (q1 ) = (9.44) e f (q 1 , q 2 ) = =

n

e

i(q 1 Rn1 +q 2 Rn2 )

n1 ,n2

e

i(q1 +q2 )·Ri

+

e

iq1 ·Rj

j

i=j

e

iq 2 ·Ri

i=j

= N δq1 +q2 ,0 + N δq1 ,0 (N − 1)δq2 .0 .

(9.45)

Whenever a sum is to be taken over the wave vectors, it can be converted into an integral with the consequence that the Kronecker symbol becomes a δ function or f (q 1 ) = cδ(q 1 )

and f (q 1 , q2 ) = cδ(q 1 + q 2 ) + c2 δ(q 1 )δ(q 2 ).

(9.46)

The power of the impurity concentration in these expressions indicates the number of impurities involved in a scattering process. Scattering with single impurities is dominating, while scattering with two or more impurities becomes less likely according to the corresponding power of c < 1. Turning back to the Dyson equation, we now perform the configuration average with the Green function by replacing G(k, k′ ; E + ) → G(k, k′ ; E + ) = G(k, k′ ; E + )conf

(9.47)

and write (suppressing the energy argument E + to simplify notation) V ′ (0) ′′ ′ 3 ′′ ′′ ′ (9.48) Σ(k, k G(k, k ) = G (k) δk,k + , k )d k )G(k (2π)3

with the self-energy Σ(k, k′′ ) due to the disorder, which in general depends on the energy.

9.3 Approximations for Impurity Scattering The configuration averaged Green function, likewise the self-energy, can be evaluated in diﬀerent approximations, which can be classified by making use of a diagrammatic representation [9, 206, 273]. This is achieved by replacing the physical objects of our formulas by graphs: G:

>

,

G(0) :

>

,

ρU :

.

9.3 Approximations for Impurity Scattering

277

The leading terms of the Dyson equation can be arranged according to their number of interaction lines with the impurities and lead to the following graphical representation of the averaged Green function G or the Dyson equation: G=

> k

= k’

+

> k

k

+

> k’

> k

>

>

k

k

+

> k

+

> > > >

+

+

> >

+

> >

> k’

> k

>

> k

+

> > >

> >

. > >

(9.49) >

>

+

> >

We recognize a growing complexity of the individual diagrams. Some of them can be reduced to simpler ones by cutting a free propagator line, while others are irreducible in this sense. This is used in the graphical representation of the self-energy, which contains only the irreducible diagrams Σ(k, k′′ ; E + ) =

+

+

+

>

+

>

>

+ > > >

+

(9.50)

+ >

>

+ > > >

>

> > >

+ > > >

+ > > >

With reference to the graphical form, we may characterize some standard approximations [9, 206, 273]. 1. Virtual Crystal Approximation (VCA): This simplest approximation consists in taking into account the impurity interaction only in lowest order. The self-energy becomes

ΣVCA (k, k′′ ) =

= ρU = cU (0)δ(k − k′′ )

(9.51)

and does not depend on the energy. For this case, the Dyson equation

278

9 Defects, Disorder, and Localization

G

VCA

VCA (k, k′ ; E + ) = G(0) (k; E + ) δk,k′ + cU (0)G (k, k′ ; E + ) (9.52)

can be solved to give VCA

G

(k, k′ ; E + ) =

δk,k′ G(0) (k; E + ) = + E − ǫk − cU (0) 1 − G(0) (k; E + )cU (0)

(9.53)

and finally VCA

G

(k, k′ ; E + ) = G(0) (k, E + − cU (0))δk,k′ .

(9.54)

In the virtual crystal approximation, the single particle energies are simply shifted by cU (0), which is the averaged perturbation by the impurities and does not break the translational symmetry of the solid. This approximation can be understood also as a replacement of the host atoms by virtual atoms, each of them modified by an averaged impurity contribution. 2. Averaged t-matrix Approximation (ATA): In this approximation, multiple scattering with a single impurity is considered in all orders as expressed by the self-energy ΣATA (k, k′′ ; E + ) =

+

+

+

>

>

= cδk,k′′ t(k, k′′ ; E + )

>

(9.55)

with the t-matrix t(k, k′′ ; E + ) = U (k − k′′ ) V U (k − k1 )G(0) (k; E + )U (k1 − k′′ )d3 k1 + . . . . + (2π)3 (9.56) 3. Born Approximation: Here only the leading terms of the ATA, the single and double scattering with an impurity are considered, i.e., the self-energy consists of two terms ΣBA (k, k′′ ; E + ) =

(9.57)

+ >

and the t-matrix takes the form tBA (k, k; E + ) = U (0) +

V (2π)3

|U (k − k1 )|2 G(0) (k1 ; E + )d3 k1 . (9.58)

4. Self-Consistent Born Approximation (SCBA): The Born approximation can be improved by replacing in the second graph the unperturbed propagator G(0) (k1 ; E + ), describing the free propagation of the electron between two scattering events, by the configuration averaged Green function G(k1 ; E + ), or in graphical form

9.3 Approximations for Impurity Scattering

ΣSCBA (k, k′′ ; E + ) =

279

(9.59)

+

>

At the end of this section, we shall apply the SCBA to a system with discrete energy levels in order to demonstrate the mechanism of level broadening due to impurity scattering. 5. Coherent Potential Approximation (CPA): It combines the SCBA with the ATA by taking into account all graphs of the latter but with the replacement of the free Green function by the configuration averaged one as in the former. Thus, the self-energy is represented by ΣCPA (k, k′′ ; E + ) =

+

+

>

(9.60)

+

> >

In all these approximations, the self-energy turns out to be diagonal in the wave vector, which allows one to solve the Dyson equation to obtain the configuration averaged Green function G(k, k′ ; E + ) =

E+

δk,k′ − ǫk − Σ(k; E + )

(9.61)

with the complex self-energy V Σ(k; E ) = cU (0) + c (2π)2 +

|t(k − k1 )|2 3 d k1 . E + − ǫk

(9.62)

Its imaginary part ImΣ(k; E + ) = c

V (2π)2

|t(k − k1 )|2 δ(E + − ǫk )d3 k1 =

¯ h τk

(9.63)

can be expressed by the quasi-particle lifetime τk and quantifies the rate by which the Bloch electron is scattered by disorder out oﬀ its state with wave vector k. Note that the approximations 1–5 consider only scattering from a single impurity, therefore, the self-energy is proportional to the impurity concentration c. In the CPA, all diagrams with up to four interaction lines, except those where these lines cross each other, are included. This point needs further discussion in Sect. 9.5. The modification of the electron energy by impurity scattering becomes particularly evident for systems with a discrete level spectrum, such as the Hubbard bands in the atomic limit (with energies at t0 and t0 + U , see Sect. 7.2) and the Landau levels of a two-dimensional electron system (with ǫn = h ¯ ωc (n + 1/2), see Sect. 5.6). Their density of states has the form D(0) (E) = gn δ(E − ǫn ) (9.64) n

with the degeneracy factor gn . Let us assume the impurity potential to be extremely short-ranged as described by

280

9 Defects, Disorder, and Localization

U (r − Rj ) = U0 δ(r − Rj )

(9.65)

with the Fourier transform U (q) = U0 . In SCBA, the self-energy can be calculated from Σ(E + ) = cU0 + c

n

U02 . E + − ǫn − Σ(E + )

(9.66)

Given that the separation of neighboring energy levels ǫn+1 − ǫn is much larger than the imaginary part of the self-energy ImΣ(E + ), which determines the level broadening due to the impurity scattering, we may approximate the self-energy for E close to the level energy ǫn by Σ(E ∼ ǫn ) = Σn (E) ≃

cU02 , E − ǫn − Σn (E)

(9.67)

which is a quadratic equation in Σn (E) and can be solved to give ΣSCBA (E) = n

1/2 i 2 E − ǫn Γ − (E − ǫn )2 − , 2 2 0

(9.68)

where Γ20 /4 = cU02 . From the corresponding Green function G(E + ) =

n

1 E + − ǫn − Σn (E)

(9.69)

we obtain the density of states DSCBA (E) = −ImG(E + )/π in the form + 2 1/2 2 Γ0 − (E − ǫn )2 for |E − ǫn | < Γ0 . SCBA πΓ20 (9.70) (E) = D 0 otherwise n The result is a broadening due to impurity scattering of the otherwise discrete and highly degenerate levels (Fig. 9.6). The half-elliptic form of the density of states is an artifact of the assumed short-range of the impurity potential. We recognize also the increase of the broadening with the impurity concentration and/or with the strength of the impurity potential. Impurity potentials of some finite range would create states in the gaps between the neighboring peaks thus leading to an overall finite density of states.

9.4 Electric Conductivity In previous chapters, we have already studied several aspects of the electric conductivity, especially in Chap. 8 we have seen that it is determined by a transport relaxation time τtr (k) which accounts for scattering processes with

9.4 Electric Conductivity

281

D(E) c c’>c

εn

εn+1

εn+2

E

Fig. 9.6. Density of states of Landau levels in a two-dimensional electron system in the SCBA. The broadening increases with the impurity concentration c

phonons. Likewise, scattering with impurities contributes to the relaxation of a non-equilibrium distribution (here written as f (k)) which using, Fermi’s Golden Rule, can be formulated as [64] 2π c f (k) − f0 (k) = |t(k − k′ )|2 δ(ǫk − ǫk′ ) f (k) − f (k′ ) d3 k′ (9.71) τtr (k) h (2π)3 ¯ with the t-matrix t(k − k′ ) and the impurity concentration c. The δ-function in the integral indicates that the impurity scattering is elastic. For an isotropic dispersion, the same arguments hold here as for the electron–phonon interaction with respect to forward and backward scattering (see Sect. 8.3). Another and more principal approach to the electric conductivity is the linear response introduced in Sect. 2.5 (see also Problem 2.4). Let us write the Hamiltonian for the electrons in a system with impurities and in the presence of an external field represented by the vector potential A(r, t) 2 & 1 ' H = d3 r Ψ† (r) p + eA(r, t) + vj (r − Rj ) Ψ(r). (9.72) 2m Rj

Here the notation with field operators Ψ† (r), Ψ(r), introduced in Problem 4.8, is used. Without the external field and using the expansion Ψ(r) = k ψk (r)ck in terms of Bloch functions ψk (r), the Hamiltonian H becomes identical with (9.31). Including the vector potential, it takes the form considered in the linear response theory H = H0 + Hdisorder + Vext (t),

(9.73)

and we may identify the external potential by comparing with (9.72) (taking only the terms linear in A(r, t)) as ( ) e Vext (t) = Ψ† (r) p · A(r, t) + A(r, t) · p Ψ(r)d3 r (9.74) 2m

282

9 Defects, Disorder, and Localization

which for the transverse gauge ∇ · A = 0 and with p = −i¯ h∇ can be written ) ie¯ h ( ∇Ψ† Ψ − Ψ† ∇Ψ · A(r, t)d3 r. (9.75) Vext (t) = 2m

The observable connected with the electric conductivity is the current density, which in terms of field operators can be written as e † ˆ Ψ (r) p + eA Ψ(r) + h.c. j(r) = − 2m e2 ie¯ h † Ψ (∇Ψ) − (∇Ψ† )Ψ − Ψ† ΨA . = 2m m ˆ j p (r)

(9.76)

ˆ j d (r)

The two contributions are called paramagnetic (ˆ j p (r)) and diamagnetic (ˆ j d (r)) current density. The diamagnetic current density is already linear in the external field, while we identify the paramagnetic current density as the one appearing in the external perturbation, which thus takes the form Vext (t) = − ˆ j p (r) · A(r, t)d3 r . (9.77) The electric conductivity is the response to an external electric field that derives from the electromagnetic potentials according to E = −∇φ − ∂A/∂t. Let us assume harmonic external fields, e.g., A(r, t) = Aei(q·r−ωt) ,

(9.78)

then, with the gauge ∇·A = 0, we can separate the electric field into a longitudinal (−∇φ) and a transverse (−∂A/∂t) contribution. The longitudinal field has been the subject of Sects. 4.5 and 4.6 in deriving the energy-loss function and studying screening in the random phase approximation (RPA). Here we are dealing with the transverse response (q ⊥ A) withelectric field components Eα = iωAα . Using the Fourier expansion ˆ j(r) = q′ ˆ j q′ exp (iq ′ · r) of the current density operator, we may write the thermal expectation value for one of its Fourier components as ′ ˆjα,q t = iω σαβ (q; t, t′ )Aβ e−iωt dt′ (9.79) with double index summation understood. Assuming homogeneity in the time dependence, this becomes jα,q (ω) = iωσαβ (q; ω)Aβ .

(9.80)

For isotropic systems, the conductivity tensor simplifies to a scalar and the ˆ used indices α, β can be dropped. We may now identify the operators Aˆ and B

9.4 Electric Conductivity

283

in Sect. 2.4, when introducing the concept of linear response, as ˆjq = ˆjqp + ˆjqd and ˆjqp , respectively, with 2

e ˆjqd = − nq A m

(9.81)

giving for the conductivity i σ(q; ω) = − ω

! e2 χ(q, ω) − nq m

with the susceptibility of the transverse response i +∞ iωτ χ(q, ω) = e θ(τ )[ˆjqp (τ ), ˆjqp (0)]0 dτ h −∞ ¯

(9.82)

(9.83)

The current density operator expressed in terms of fermion operators for free particle states reads " pˆ " " " ˆjqp = −e k + q " "k c†k+q ck (9.84) m k

h/m) k kc†k ck and yields for the which for q → 0 simplifies to ˆj p = (e¯ susceptibility h2 ∞ iωτ ′ † e2 ¯ χ(0, ω) = i (9.85) kk [ck (τ )ck (τ ), c†k′ (0)ck′ (0)]0 dτ . e h m2 0 ¯ ′ k,k

In a more general single-particle basis, the current operator is "" pˆ "" † ˆj p = −e i" "f ci cf m

(9.86)

i,f

and with c†i (τ ) = c†i exp (iǫi τ /¯ h) and the Fermi–Dirac distribution function c†i ci = f (ǫi ) the current–current response can be cast into the form (Problem 9.3) χ(0, ω) = −e2

i,f

|i|ˆ v |f |2

f (ǫi ) − f (ǫf ) . hω + ǫi − ǫf + iδ ¯

(9.87)

When decomposing χ(0, ω) for δ → 0 into real and imaginary part, one finds that the real part compensates the contribution ne2 /m of the diamagnetic current density in σ(0, ω) (9.82) [64, 274] and obtains the real conductivity σ(0, ω) =

πe2 |i|ˆ v |f |2 (f (ǫi ) − f (ǫf )) δ(¯hω + ǫi − ǫf ) ω i,f

(9.88)

284

9 Defects, Disorder, and Localization

known as the Kubo–Greenwood formula. For the metal–insulator transition to be discussed in the next section, we need the dc conductivity σdc , which is the static limit σ(0, ω → 0). As for ω → 0, only states close to the Fermi energy contribute to the conductivity, we may finally write

σdc = πe

2

df0 (E) − dE

! i,f

|i|ˆ v |f |2 δ(E − ǫi )δ(E − ǫf )dE.

(9.89)

For further discussion, we may cast the double sum under the integral into the form Tr(ˆ v ImG(E + )ˆ v ImG(E + )) (Problem 9.4). This form, with the two Green functions appearing, derives from the fact that the commutator in (9.85) contains four fermion operators, each two of which would be required for a single-particle Green function. In fact, it is possible to formulate the dc conductivity in terms of a configuration averaged product of a retarded (G+ ) and an advanced Green function (G− ) [64, 274, 276]. Without going into the details of the derivation, for which we refer to the literature, we quote here in short notation the equation from which it derives. It is the Bethe2 –Salpeter3 equation G+ G− = G+ G− + G+ G− V G+ G−

(9.90)

where V , the irreducible vertex operator, plays the same role for the twoparticle Green function as the self-energy does for the single-particle Green function. G+ G− describes the propagation of an electron–hole pair (G− representing the hole) under the influence of impurity scattering, which not only aﬀects the electron or hole propagation separately but can cause also a coupling of electron and hole propagation. The latter is considered in the vertex operator. Similar to the single-particle Green function, the two-particle propagator can also be represented by diagrams. A first group of simple diagrams contains the non-connected graphs for electrons and holes and corresponds to the approximation V = 0 G+ G− ≃ G+ G−

(9.91)

with separate configuration average for the electron and hole propagator. A second group of more complex diagrams (so called ladder diagrams) connects the electron and hole propagator in the simplest way, e.g.,

2 3

G+ :

>

>

G− :

>

>

Hans Albrecht Bethe 1906–2005, Noble prize in physics 1967 Edwin Ernest Salpeter *1924

9.5 Metal–Insulator Transition

285

> >

>>

>>

>

> >

>

Fig. 9.7. Sketch of an electron path (solid) and its time inverted path (dashed) in an impurity configuration leading to quantum interference and weak localization

and accounts for coherent scattering of electron and hole at the same impurity. Summing these two groups of diagrams leads to the Drude result of the dc conductivity σ = ne2 τ /m = e2 kF2 lmfp/3π 2 ¯h ,

(9.92)

where τ is the single particle life-time and lmfp the mean free path. Finally we mention diagrams, which contain crossing interaction lines between the electron and hole propagator, like G+ : G− : which describe the coherent scattering at an impurity configuration taking place for the hole in the reversed order with respect to that of the electron. Considering these so called maximally crossing diagrams is known in the literature as cooperon approximation, as they represent an interaction between electrons on their time-inverted paths along an impurity configuration in analogy to the electron–phonon mediated formation of Cooper pairs. The situation is depicted in Fig. 9.7. The wave functions corresponding to the two time inverted paths through the impurity configuration interfere positively with each other and lead to a localization due to quantum interference. It is known as weak localization and gives a dc conductivity that corresponds to the one found by solving the Boltzmann equation including the eﬀect of the back scattering which is characteristic for the transport relaxation time [274]. Removal of this interference by a magnetic field leads to a decrease of the resistance and can be taken as a signature of weak localization.

9.5 Metal–Insulator Transition Diﬀerent sections of this chapter have provided several aspects of disorder, and the influence it takes on the electron states. As we have seen in Sect. 9.1, single impurities can lead to discrete energy levels in the band gap region

286

9 Defects, Disorder, and Localization

E

V(r), ψ(r)

extended

2λ

localized

r

D(E)

Fig. 9.8. Schematic view of an electron wave function (thin solid line) in a random potential (dashed line, left) and density of states (right) indicating extended and localized states

of semiconductors. These (shallow or deep) bound impurity states are localized, and their wave functions decay exponentially away from the center. But, impurities can modify also the electronic states within a band as resonant states. However, such modifications are expected to be small in regions of high density of extended states. With increasing impurity concentration the bound state wave functions start to overlap and lead to the formation of an impurity band with a continuous density of states, which can eventually merge with that of the nearby energy band. This situation is schematically depicted in Fig. 9.8. Instead of a sharp band edge resulting from the periodic crystal potential as shown in Fig. 9.2, the electrons close to the band edge experience a random potential caused by the impurities. In this potential landscape, states with low energy are localized with a finite localization length λ, while those with suﬃciently high energy are not confined and extend over the whole solid as λ increases to infinity. This scenario implies that there must be a certain critical energy ǫc which separates the extended from the localized states and that it is likely to be in a region with small density of states or close to the band edges. Let us consider also a scenario, which does not start from an energy band with a continuous density of states, but from a discrete spectrum as that of Landau levels for two-dimensional electrons (see Sect. 5.6 and Fig. 9.6). Without impurities, we have a homogeneous system which is invariant under translations in the plane perpendicular to the magnetic field. Therefore, the Landau levels are discrete but highly degenerate, and the electron wave functions are delocalized. A random potential breaks the translational invariance and lifts the degeneracy of the Landau levels leading to their broadening into a narrow band. In Sect. 9.3, we have discussed the eﬀect of short-range impurities within SCBA with the result shown in Fig. 9.6. The states in the center of the Landau band with the large density of states are still extended, while those at the edges are localized due to the potential fluctuations, and we have again the scenario sketched in Fig. 9.8.

9.5 Metal–Insulator Transition

287

These conclusions, with respect to a clear separation between extended and localized states, are not restricted to substitutional disorder caused by impurities but are valid also for structural disorder as in amorphous solids. In fact, amorphous solids, especially amorphous semiconductors, have been the systems for which the concepts of a disorder induced metal–insulator transition have been developed [39, 209, 211, 275, 276]. Consider now electron transport in a solid with disorder. As we have seen in the previous section, the dc conductivity is determined by the states around the Fermi energy and would be carried for T = 0 K alone by those at the Fermi energy. It follows from the discussion of Fig. 9.8 that we can distinguish between solids or experimental conditions for which the Fermi energy falls into a region of extended states and those with EF in a region with localized states. For the former, the dc conductivity σdc at T = 0 K is finite, these are the metals, while it vanishes for the latter, which are the insulators. Instead of the dc conductivity, we may also use the relation σdc = enμ and distinguish these systems by their mobility μ, which at T = 0 K is finite for extended states but drops to zero when entering the region of localized states. This defines the mobility edge that marks the critical energy ǫc for the transition from a metal to an insulator. In spite of this clear distinction between metals and insulators, the criterion is hard to test in experiments, because of the diﬃculties to approach the T = 0 K limit and also due to the fact that experiments are always performed with finite samples. Let us have a look, therefore, on less restrictive conditions. In an insulator, with the Fermi energy below the mobility edge, already a very low but finite temperature would allow the localized electrons to reach by thermal excitation nearby unoccupied localized states. Thus step by step these electrons carry a current by a hopping process (variable range hopping), which leads to the characteristic dependence of the dc conductivity on the inverse temperature of Mott’s T −1/4 law [39, 275]. The existence of a mobility edge in an electron system with disorder is crucial for understanding the vanishing of the longitudinal magneto-resistivity of two-dimensional electron systems for magnetic fields around integer filling factors (see Fig. 5.21). In the original QHE experiments [161], the carrier density and thus the Fermi energy was changed by the external gate voltage while the sample was exposed to a fixed magnetic field. Changing the carrier density shifts the Fermi energy through the spectrum of Landau levels which due to the disorder is broadened. If EF is close to the center of a Landau band, in a region of extended states, the dc conductivity is finite, while away from the center, and between the Landau levels it hits localized states and the dc conductivity drops to zero. In a Quantum Hall experiment with fixed carrier density (or Fermi energy), an increasing magnetic field increases the degeneracy and separation of the Landau levels, which shift one by one through the Fermi energy. As a function of the magnetic field, EF (B) coincides with extended or localized states connected with finite or zero dc conductivity, the latter always connected with integer filling factors. Thus, one has a sequence

288

9 Defects, Disorder, and Localization

of metal–insulator transitions within one electron system by sweeping either the magnetic field for fixed carrier density or vice versa. Actually Figs. 5.21 and 7.12 show resistivities, and we have to relate our conclusions to this quantity by inverting the resistivity tensor ! ρxx ρxy ρ= (9.93) ρyx ρyy giving 1 σ= ρxx ρyy − ρxy ρyx

ρyy −ρyx −ρxy ρxx

!

.

(9.94)

This simplifies for rotational symmetry around the normal to the plane of the two-dimensional electron system with ρxx = ρyy and for finite ρxy to the relation ρxx ∼ σxx . Thus, vanishing longitudinal resistivity means also vanishing longitudinal conductivity. The question of the quantitative connection between disorder and the quantum mechanical states around the metal–insulator transition found an answer in some fundamental contributions, using scaling arguments to account for the influence of the finite size of the sample [280, 281] and resulted in understanding the MIT as a second order phase transition. For a finite sample (e.g., a cube of typical length L), the relevant transport quantity is not the conductivity (which is a material property independent of the sample size and shape) but the conductance G(L), defined as the ratio of the measured current through the sample and the applied voltage. In terms of the fundamental conductance unit e2 /h, we may define also the dimensionless quantity h g(L) = 2 G(L). (9.95) e If instead of the box of size L a larger box of size L′ > L is considered, the scaling argument is applied to find g(L′ ). It is based on the assumption that the ratio between the relative changes of g and of L dg L d ln g = = β(g) g dL d ln L

(9.96)

is ruled by a universal function β(g) depending on the dimensionless conductance g but not separately on the parameters characterizing the sample such as disorder, energy, or sample size L. Solving the diﬀerential equation for the asymptotic limits yields β(g) = 1 −

gc g

β(g) = −egc /g

for large g

(9.97)

for small g.

(9.98)

Assuming a monotonous function β(g), these results can be connected by an interpolation to obtain the qualitative result shown in Fig. 9.9 (Problem 9.5).

9.5 Metal–Insulator Transition

289

β 1

>

0

gc

g

> Fig. 9.9. Universal function β(g) for the scaling behavior of the conductance g(L) in a three-dimensional system. For β > 0 the conductance increases while for β < 0 it decreases with increasing system size L

For β(g) > 0, the conductance increases with increasing sample size, which corresponds to metallic behavior, while for β(g) < 0 it goes to zero for L → ∞, typical for the insulator. The critical value gc with β(gc ) = 0 marks the metal– insulator transition. This one-parameter scaling theory has been successful in describing the critical behavior of the conductivity σdc when approaching the transition from the metallic side (E > ǫc ) and of the localization length λ, when approaching it from the side of the insulator (E < ǫc ). In accordance with the properties of a second order phase transition [277], these quantities follow power laws: σdc ∼ (E − ǫc )s , E > ǫc λ ∼ (ǫc − E)−ν , E < ǫc .

(9.99)

From scaling relations, the critical exponents s and ν are found to be equal. A particular aspect of this theory, which more recently has attracted renewed interest, is its dependence on the dimension d of the system [280]. As was argued by Thouless [281], it is the ratio of the level broadening due to disorder Γ(L) and the level separation ∆E(L) due to confinement in a system of size L which decides about the character of the quantum states at the Fermi energy as being localized or extended. The former can be expressed as Γ(L) = h/tD (L), where tD (L) is the time for the electron to pass through the sample by diﬀusive motion. This time is determined by the diﬀusion constant D according to tD (L) = L2 /D, thus Γ(L) = hD/L2 . For a particle in a box of −1 the size L, the level separation at EF is ∆E(L) = n(EF )Ld with the particle density n(EF ) and the system dimension d. Using the Einstein relation σ = e2 Dn(EF ), the ratio Γ(L)/∆E(L) can be expressed as h Γ(L) = 2 σLd−2 ∆E(L) e

(9.100)

290

9 Defects, Disorder, and Localization

and we identify the dimensionless conductance g(L) = σLd−2 , where the conductivity σ characterizes the material property and does not depend on L. Using this relation, the scaling argument (9.96) can be discussed with respect to its dependence on the system dimension d (see Problem 9.5). As it turns out, the universal function β(g) passes through zero only for a three-dimensional system. For a two-dimensional system, it approaches zero asymptotically for large g and is always negative for d = 1. This means that a disorder-driven metal–insulator transition should be possible only for d > 2. Detailed investigations of two-dimensional electron systems seem to disprove this conclusion [278]. The critical discussion in the light of these data leads to the result, that possibly disorder is not the only mechanism but that electron–electron interaction may contribute, which is not included in the scaling theory.

Problems 9.1 Verify the eigenvalue equation (9.27) for the additional poles of the full Green function G(E) (9.20) by making use of the block-diagonal form of the short-range defect matrix. 9.2 To demonstrate self-averaging, consider the classical resistance R = N i=1 ri of N resistances ri , i = 1 . . . N in a row. Make sure that the average r exists and that R = N r. Calculate the relative variance of R and show that it tends to zero with increasing N or system size! 9.3 Calculate the current–current response χ(0, ω) of (9.83) with the current operator ˆj p of (9.86) to obtain the conductivity σ(0, ω) (9.88). 9.4 Show that the double sum expression under the integral of σdc (9.89) can be cast into the form Tr(ˆ v ImG(E + )ˆ v ImG(E + )). 9.5 Discuss the asymptotics of β(g) for g → ∞ (using σ independent of L) and for g → 0 (assuming g(L) ∼ exp (−L/λ)) in dependence on the system dimension d. Show that β(g) < 0 for all g and d ≤ 2 while it changes sign for d > 2.

10 Light–Matter Interaction

The investigation of condensed matter systems using light as a probe is a versatile and powerfull concept, which provides various kinds of information not only about the structure but also of the electronic and lattice excitations and their dynamics. Besides linear optical processes, like absorption and emission (including photoemission), light-scattering, two- (or three-)photon absorption, photo-luminescence and high-excitation spectroscopy also belong to the methods to gain information about the single-particle and collective excitations of a solid. In semiconductor samples with a proper design, coherent light emission is possible and used in solid-state lasers. These methods, which allow studies in the frequency or time domain, are all based on the interaction of light with matter. In Sect. 3.5 we have already treated the optical excitation of lattice vibrations as an example. In this chapter, we shall present a systematic theoretical description of light–matter interaction, which can be found in one or the other representation in standard textbooks[7, 14, 89, 95, 282, 283].

10.1 Preliminaries Let us start with the microscopic Maxwell equations1 ∇·E =

ρ ∂B , ∇×E = − ε0 ∂t

∇ · B = 0 , ∇ × B = μ0 j +

1 ∂E . c2 ∂t

(10.1)

Here, E(r, t) and B(r, t) are the space and time dependent electric field and magnetic induction, respectively. The latter can be replaced with B = μ0 H by the magnetic field H(r, t). ε0 and μ0 are the vacuum constants and √ c = 1/ ε0 μ0 is the vacuum velocity of light. In this microscopic formulation, ρ(r, t) and j(r, t) are the charge and current densities, respectively, of 1

James Clark Maxwell 1831–1879.

U. R¨ ossler, Solid State Theory, c Springer-Verlag Berlin Heidelberg 2009 DOI 10.1007/978-3-540-92762-4 10,

292

10 Light–Matter Interaction

all charged particles in the system (nuclei and electrons). We could adopt here the reasonable concept of the previous chapters, according to which we understand the solid as composed of ions and valence electrons. In this case ρ(r, t) and j(r, t) would be connected with the latter, while the electrons in closed shells would belong to the ions, which essentially do not contribute to the material properties of the solid. We come back to this point later in this Chapter. For what follows, instead of the fields E(r, t) and B(r, t), it is convenient to introduce the potentials A(r, t) and φ(r, t), which are related to the fields by B = ∇ × A , E = −∇φ −

∂A . ∂t

(10.2)

Replacing E and B in Maxwell’s equations by the potentials A and φ, we may write (using ∇ × ∇ × A = ∇(∇ · A) − ∆A) the inhomogeneous equations in the form ! 1 ∂2A 1 ∂φ ∆A − 2 2 − ∇ ∇ · A + 2 = −μ0 j (10.3) c ∂t c ∂t and ∆φ +

ρ ∂ (∇ · A) = − . ∂t ε0

(10.4)

The potentials are free to a gauge transformation with the scalar gauge field χ(r, t) ∂χ A → A + ∇χ , φ → φ − , (10.5) ∂t which leaves (10.2), i.e., the fields E and B, invariant. Among the commonly used gauges, the transverse or Coulomb gauge, ∇ · A = 0, is the appropriate one for the interaction of light with matter in the form of solids. With this choice, the third term on the lhs of (10.3) vanishes while (10.4) becomes Poisson’s equation. Thus, we obtain the potential form of Maxwell’s equations in the form of an inhomogeneous wave equation for the vector potential 1 ∂2A ∆A − 2 2 = −μ0 j. (10.6) c ∂t The inhomogeneity is connected with the nuclei and electrons (or ions and valence electrons) and vanishes outside the solid, where (10.6) describes the free propagation of electro-magnetic waves in vacuum. Consider now the solid interacting with the electro-magnetic field of the light. In the most general case, it is described by the Schroedinger equation with the Hamiltonian of the solid (2.1), including the electro-magnetic field. After having dealt already with the excitation of optical phonons, which are connected with the moving ions (see Chap. 3.), we would like to consider here only the electrons as movable charged particles in a rigid periodic arrangement

10.1 Preliminaries

293

of nuclei or ions. Thus, we restrict in this chapter to the interaction of light with the electrons of the solid and start accordingly with the Hamiltonian (5.2) for the electrons. To include the electro-magnetic field, we have to take into account its energy and to fulfill the requirement, that the time-dependent Schroedinger equation ∂ΨN HN ΨN = i¯ (10.7) h ∂t has to be invariant under the gauge transformation (10.5), together with a simultaneous gauge transformation of the time-dependent N -electron wave function ΨN (r 1 . . . r N , t) e Ψ(r1 . . . r N , t) → exp −i {χ(r 1 , t) + · · · + χ(r N , t)} Ψ(r1 . . . rN , t) . h ¯ (10.8) This is achieved by the replacement pl → pl + eA(rl , t) .

(10.9)

For the rest of this chapter the following Hamiltonian will be the starting point:

HN =

N e2 1 1 (pl + eA(r l ))2 + + v(r l − R0n ) 2m 2 k,l=1 4πε0 |r k − r l | l=1

+

n,l

k=l

ε0 1 V (R0n − R0m ) + 2 mn 2

d3 r E 2 + c2 B 2 .

(10.10)

In the last term, describing the energy of the radiation field, E and B can be replaced with (10.2) by the vector potential to yield , + !2 ε0 ∂ 2 2 3 A(r, t) + c (∇ × A(r, t)) d r . (10.11) Hrad = 2 ∂t Having in mind linear response theory (Sect. 2.4), we may separate HN into a Hamiltonian for the uncoupled solid and electro-magnetic field, H0 + Hrad , and an interaction term Hel−rad for the light–matter coupling HN = H0 + Hel−rad + Hrad ,

(10.12)

where H0 describes the unperturbed electron system of the solid, (5.2), and (by making use of the transverse gauge) Hel−rad =

N l=1

e e2 A(r l , t) · pl + (A(rl , t))2 m 2m

!

(10.13)

294

10 Light–Matter Interaction

is the time-dependent perturbation by the radiation field, which can be identified as the external potential of (2.37). Using j = −e l pl /m, the first term can also be written as −j · A (see Sects. 2.4 and 9.4).

10.2 Single-Particle Approximation In Chap. 5 we have invoked density-functional theory (DFT) to reduce the many-body problem with H0 to a single-particle problem, by accounting for the electron–electron interaction in an eﬀective single-particle potential Veﬀ (r) (see (5.30)), which is the same for all electrons. This can now be used to write the many-body Hamiltonian HN as a sum of single-particle terms. Leaving aside the energy of the radiation field as a constant, the single-particle Hamiltonian of an electron interacting with the radiation field takes the form H =−

e2 2 ¯2 h e ∆ + Veﬀ (r) + A · p + A . 2m m 2m

(10.14)

The eigenstates of the first term are the Bloch states of Chap. 5. The last two terms represent a periodic time-dependent interaction of Bloch electrons with the radiation field. For suﬃciently weak amplitudes of the vector potential they can be treated as perturbations of first and second order. Sometimes the Ap -term is also expressed in the form −eE · r, where −er is the dipole operator (Problem 10.1). Let us first treat the term linear in A. For monochromatic light, its space and time dependence is given by A(r, t) = A0 e ei(κ·r−ωt) + c.c. ,

(10.15)

with a scalar amplitude A0 and the unit vector e for transverse polarization, e ⊥ κ. The frequency follows the linear dispersion relation ω = c|κ| of light in vacuum. Because of the periodic time-dependence of the perturbation, Fermi’s Golden Rule applies and we can immediately write the rate for the transition probability between Bloch states Wnk,n′ k′ =

2π 2 |nk|H ′ |n′ k′ | δ (En (k) − En′ (k′ ) − ¯hω) , h ¯

(10.16)

where H ′ = eA0 exp(iκ · r)e · p/m is the perturbation without the timedependent exponential. Together with the time-dependent phase factors of the stationary Bloch functions, it results in the δ-function, which expresses the energy conservation: the energy quantum h ¯ ω of the radiation field equals the energy diﬀerence between the Bloch states. Depending on the sign of this diﬀerence, the transition between the Bloch states can be an absorption or emission process.

10.2 Single-Particle Approximation

295

The transition matrix element can be transformed from an integral over the crystal volume Vc e ∗ A0 eiκ·r e · p ψn′ k′ (r)d3 r (10.17) ψnk (r) nk|H ′ |n′ k′ = m Vc to an integral over a Wigner-Seitz cell by shifting r by lattice vectors Rn and extracting the corresponding phase factors from the Bloch functions ′ nk|H ′ |n′ k′ = ... (10.18) ei(κ+k −k)·Rn d3 r . W SC

n

Here the dots stand for the integrand of the previous integral. The lattice sum results in a Kronecker δk,k′ +κ , which represents the momentum conservation of the excitation process. The diameter of the Wigner–Seitz cell is about the lattice constant of a few ˚ A, which is much smaller than the wavelength of the light corresponding to the typical excitation energy h ¯ ω of a few eV . Therefore, the exponential with κ in the integral is almost constant over the cell and we may safely replace it by 1. This is the dipole approximation. Finally, the transition matrix element can be written (see also Problem 10.2) e A0 e · p un′ k′ (r)d3 r δk,k′ nk|H ′ |n′ k′ = u∗nk (r) m W SC e (10.19) = A0 nk|e · p|n′ kδk,k′ . m Thus, one obtains a finite transition probability Wnk,n′ k (ω) =

2π e2 A20 2 |nk|e · p|n′ k| δ (En (k) − En′ (k) − ¯hω) h m2 ¯

(10.20)

only for direct transitions (or vertical transitions), in which the wave vector of the Bloch state is not changed. The absorption of light passing through a solid of thickness d determines the damping of the light intensity according to I(d) = I0 e−αd .

(10.21)

On the other side, the absorption coeﬃcient α(ω) is connected with the real (n1 ) and imaginary (n2 ) parts of the complex index of refraction according to α=

2n1 n2 ω, c

(10.22)

where 2n1 n2 = ε2 is the imaginary part of the complex dielectric function ε(ω). Microscopically, α(ω) is determined here by the power loss or the rate by which energy of the radiation field is converted into excitations between

296

10 Light–Matter Interaction

Bloch states[14, 89]. Thus, α(ω) =

2¯h W (ω) , where W (ω) = Wnk,n′ k (ω), 2 ε0 cωA0 ′

(10.23)

nn k

and the imaginary part of the dielectric function is given by

ε2 (ω) =

4πe2 2 |nk|e · p|n′ k| δ (En (k) − En′ (k) − ¯hω) . (10.24) ε 0 m2 ω 2 ′ nn k

The real part of ε1 (ω) can be obtained using the Kramers–Kronig relation (2.80) formulated for susceptibilities χ(ω). Considering the relation ε(ω) = 1 + χ(ω)/ε0 , we have to use 2 ∞ ω ′ ε2 (ω ′ ) ′ ε1 (ω) = 1 + dω , (10.25) π 0 ω′2 − ω2 leading to ε1 (ω) = 1 +

4e2 fnn′ (k) ε0 m ′ ωnn′ (k)2 − ω 2 nn k

(10.26)

with the oscillator strength 2

fnn′ (k) =

2 |nk|e · p|n′ k| m¯ hωnn′ (k)

(10.27)

for the transition between Bloch states with the energy diﬀerence ¯hωnn′ (k) = En (k) − En′ (k). In cubic solids, the dipole matrix elements do not depend on the polarization direction e and usually their dependence on the wave vector is not very strong. Thus, we can write nk|e · p|n′ k ≃ Pnn′

(10.28)

with Pnn′ independent of k and ε2 (ω) becomes 4π e 2 2 |Pnn′ | Dnn′ (¯hω) , ε0 mω ′

(10.29)

with the combined density of states δ (En (k) − En′ (k) − ¯hω) Dnn′ (¯hω) =

(10.30)

ε2 (ω) =

nn

k

10.2 Single-Particle Approximation

297

Fig. 10.1. (a) Band structure of Ge with vertical transitions at critical points and (b) comparison of measured and calculated ε2 (ω), showing van Hove singularities at these critical points (from Landolt-B¨ ornstein [1])

of the pair of bands n, n′ . As in Sect. 3.3, the k-sum over the Brillouin zone can be converted into an integral Vc Dnn′ (¯hω) = (10.31) δ (En (k) − En′ (k) − ¯hω) d3 k , (2π)3 BZ which, due to the δ-function, reduces to an integral over the surface Sk , for which the energy diﬀerence of the Bloch states equals ¯hω Vc dSk Dnn′ (¯hω) = . (10.32) 3 (2π) Sk |∇k (En (k) − En (k))| The points, at which |∇k (En (k) − En (k))| vanishes (critical points), give rise to van Hove singularities in the combined density of states, by which they can be identified. As an example, we show in Fig. 10.1 the band structure of Ge in the fundamental band gap region with marked direct transitions at critical points and the measured and calculated imaginary part of the dielectric function, ε2 (ω). In spite of discrepancies in the critical point energies, there is a clear correspondence between the experimental and theoretical curves. It turns out, that the optical spectra of semiconductors with tetrahedral coordination resemble each other [89]. Their band structures have the same topology determined by the common crystal structure but diﬀer with respect to their critical point energies E0 , E0′ , . . . . Investigations of the optical spectra have paved the road to the quantitative understanding of semiconductor band structures. Light–matter interaction stood at the beginning of quantum mechanics with Einstein’s explanation of the photo-electric eﬀect, which earned him the Nobel prize in 1922. It is based on the assumption of a quantized radiation field. Photoemission can take place if the photon energy suﬃces to excite an

298

10 Light–Matter Interaction

electron beyond the energy barrier at the surface of the solid so that it can be detected outside the solid as free particle. Einstein found (in accordance with experimental data for metals) the linear dependence of the maximum kinetic energy of the emitted electrons on the photon energy, which is ruled by the ratio h ¯ /e. Later, this eﬀect has been exploited by Siegbahn2 to develop the concept of photo-electron spectroscopy (PES). It is based on analyzing not only the maximum energy but also the energy spectrum of the photo-emitted electrons. The minimum energy required to free an electron from a metal, is the diﬀerence between the vacuum level and the Fermi energy, known as the work function. In a semiconductor, with the Fermi energy somewhere in the gap, this energy would not be suﬃcient for photoemission, because the highest occupied state, in the intrinsic case, is the top of the valence band. Instead it is the ionization energy that defines the threshold for photoemission. The photoemissive yield is roughly proportional to the density of the initial states. PES is applied at diﬀerent photon energies: ultraviolet light (UPS: ultraviolet PES) for the investigation of valence electron structures and X-rays (XPS) for the study of the more tightly bound core electrons[65, 287]. Note, that for the latter case one has to treat the core electrons in the same way as the valence electrons. The photo-emitted electrons can be also analyzed with respect to their momentum parallel to the surface, which does not change when the electron leaves the solid. This is done in angular resolved photoemission (ARPES) experiments, which give information about the energy bands En (k), as shown in Fig. 10.2 for Cu as an example. The experimental data points map out all details of the energy bands, which we already know from Chap. 5. Further variants of photo-electron spectroscopy are spin-polarized UPS (SPUPS) and the inverse photoemission, in which an electron of known energy is injected and the emitted photon is detected. The former method allows to study the bands of minority and majority spins in ferromagnets (see Fig. 6.1), and the latter yields information about the unoccupied states above the Fermi energy.

10.3 Excitons The basic assumption of the single-particle approximation in Sect. 10.2 is the same eﬀective single-particle potential Veﬀ (r) for electrons in all bands. This assumption is correct for the ground state of the solid, which, for a semiconductor, is characterized by filled valence and empty conduction bands. But for the excited state an electron in the valence band is missing, because it is in the conduction band. Intuitively, this is the two-particle problem of an 2

Kai Manne B¨ orje Siegbahn 1918–2007, Nobel prize in Physics 1981, together with N. Bloembergen and A.L. Schawlow.

299

energy (eV)

10.3 Excitons

wave vector

Fig. 10.2. Calculated energy bands of Cu and data from angular resolved photoemission, after [31]

electron–hole pair, if we consider the remaining N − 1 electrons in the valence band as the hole and the missing contribution to the eﬀective potential as the attractive Coulomb interaction between the electron in the conduction band and the hole in the valence band. In eﬀective-mass approximation, the Hamitonian for the electron-hole pair can be written He−h = Eg +

pe ph e2 . + − 2me 2mh κ|re − rh |

(10.33)

Note, that the kinetic energy of the hole increases in the downward bent parabola for the valence band. If we only had the two bands (one valence and one conduction band) in vacuum, then κ = 4πε0 . However, the correct description has to include all the other valence and conduction bands and the possible excitations between these bands, i.e., the two-band model for electron and hole is embedded in a dielectric with material constant ε and, therefore, κ = 4πε0 ε, i.e., the Coulomb interaction is screened. We would like to mention here, that the electron–hole interaction is just the diﬀerence of the Hartree potentials in the excited and ground states, but there should also be exchange and correlation contributions, which we leave aside for the moment. In relative and center-of-mass coordinates, using for pi = h ¯ ki , i = e, h, ke = k +

mh me Q , kh = k − Q, M M

(10.34)

300

10 Light–Matter Interaction

with the total mass M = me + mh , we can write the Hamiltonian for the electron–hole pair as He−h = Eg −

e2 ¯h2 Q2 ¯2 h ∆+ + , 2μ κr 2M

(10.35)

with r = |r e −rh | and the reduced mass μ = (1/me +1/mh )−1 . Except the last term describing the free center-of-mass motion of the electron-hole pair, this is the hydrogen model characterized by the Rydberg energy Rexc = RH μ/ǫ2 and the Bohr radius aexc = aH ε/μ, where RH and aH are the corresponding constants of the hydrogen atom. The eigenvalue problem with He−h for Q = 0 ! e2 h2 ¯ ∆+ φν (r) = Eν φν (r) (10.36) He−h φν (r) = Eg − 2μ κr gives a series of bound states (ν = n = 1, 2, 3 . . . ), the excitons, at En = Eg − Rexc /n2 . For typical semiconductor data ε = 10 and μ = 0.1, the binding energies of the excitons are much smaller than the typical band gap Eg , and aexc extends over many unit cells. Thus the continuum approximation, assuming eﬀective masses and a homogeneous dielectric, is justified. This case of weakly bound excitons is known as the Wannier–Mott exciton, which is similar to the shallow impurities (see Sect. 9.1). Besides these bound states, for energies larger than Eg , there is a continuum of electron–hole pair or band-toband excitations. For a simple two-band model, the single-particle and exciton pictures are presented in Fig. 10.3. It shows in the left part a valence and a conduction band with a hole at kh and an electron at ke , respectivly, while the exciton picture in the right hand part contains the bound states with their center-of-mass parabolas below the electron–hole continuum and the ground state at Q = 0. E

E

cb

>

>

R exc

n=2

>

>

Eg

n=1

Egx

kh

>

>

ke

k +

Q= ke–kh

vb

Fig. 10.3. Two-band model with electron-hole pair excitation (left) and exciton spectrum with bound states and electron–hole continuum (right)

10.3 Excitons

301

Fig. 10.4. Measured absorption spectrum of GaAs close to the fundamental edge, showing excitonic structure. The dashed curve is the calculated absorption in singleparticle approximation (from Landolt-B¨ ornstein [1])

As a representative example, the optical absorption for GaAs close to the fundamental gap is shown in Fig. 10.4. The bound states dominate the spectrum for h ¯ ω < Eg , with discrete narrow lines for n = 1, 2, 3 . . . . (The small satellites are due to impurities.) The calculated absorption in single-particle approximation drawn for energies h ¯ ω > Eg deviates from the experimental data, thus indicating that the electron–hole interaction also modifies the continuum part of the spectrum. In order to understand these exciton features in detail, we consider the many-body picture [289]. The ground state of the semiconductor (or insulator) is characterized by filled valence and empty conduction bands. Let us denote it by |Ψ0 (the same symbol has been used for the filled Fermi sphere in Chap. 4). The simplest excited state is an electron–hole pair |Ψvkh cke = c†cke cvkh |Ψ0 ,

(10.37)

which is obtained by applying appropriate fermion operators to |Ψ0 . We may combine such pairs with same total momentum ¯hQ, Q = ke − kh , which is a good quantum number for excitations, to form an exciton state |ΨνQ = (10.38) φνQ (k)|Ψvkh cke , vkh ,cke ke −kh =Q

where φνQ (k) is the Fourier transform of the exciton envelope, which describes the relative motion of the electron–hole pair as a solution for (10.36). In this picture, represented in the rhs of Fig. 10.3, optical transitions take place from the ground state (with zero total momentum) to the exciton state with the rate of transition probability

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10 Light–Matter Interaction

W0→νQ =

2π h ¯

eA0 m

!2

" "2 "ΨνQ | eiκ·rl e·pl |Ψ0 " δ(EνQ − E0 − ¯hω). (10.39) l

The Ap -coupling is a sum over single-particle terms and its matrix element eiκ·rl e · pl |Ψ0 = φνQ (k)cke |eiκ·r e · p|vkh (10.40) ΨνQ | l

can be expressed as matrix element between the Bloch states of the electron– hole pair. In dipole approximation, replacing the exponential by 1, transitions are possible only to the exciton states with Q = 0 (direct or vertical transitions). Let us assume, as in Sect. 10.2, that the momentum matrix elements are almost independent of k = ke = kh , to arrive at W0→νQ

2π = h ¯

eA0 m

!2

" " "φν (0) e · P cv "2 δ(Eν − E0 − ¯hω),

(10.41)

where we have dropped the index Q = 0 and replaced k φν (k) = φν (r e − r h = 0). By comparing with the corresponding result from Sect. 10.2 for uncorrelated interband excitations, we can immediately express the real part of the dielectric function in terms of exciton quantities: ε1 (ω) = 1 +

4e2 fν0 2 − ω2 ε0 m ν ω0ν

(10.42)

with the exciton oscillator strength fν0 =

2|φν (0)|2 |c|e · p|v|2 . m(Eν − E0 )

(10.43)

This result already contains, what we need to understand the experimental spectrum of Fig. 10.4, without converting it into the expression for the absorption. The discrete bound states (for ν = n = 1, 2, . . . ) are seen as the sharp lines at energies Eg −En , with intensities scaling according to |φn (0)|2 ∼ 1/n3 , if we adopt the wave functions of the hydrogen model (Problem 10.3). But also in the continuum of band-to-band transitions, the transition probability is modified by the factor |φν (0)|2 , which is responsible for the enhancement of the absorption intensity at the fundamental edge. With the solution of Problem 10.3, it is possible to express the absorption coeﬃcient for ¯hω > Eg as α(ω) = α0 (ω)C(ω),

(10.44)

with the absorption coeﬃcient α0 (ω) of the uncorrelated electron–hole pair of (10.23) and the Coulomb enhancement factor or Sommerfeld correction √ 2π x Rexc √ , with x = . (10.45) C(ω) = ¯hω − Eg 1 − exp(−2π x)

10.3 Excitons

303

A more rigorous treatment of the electron–hole correlation is possible by using the two-particle Green function G [64, 288], for which we may write G = G0 + G0 ΓG0 .

(10.46)

Here, G0 is the two-particle Green function of the uncorrelated electron–hole pair and the kernel Γ satisfies the Bethe–Salpeter equation (BSE) Γ = I + IG0 Γ,

(10.47)

with the irreducible electron–hole interaction to be specified later. After multiplication from the right by G0 , the BSE can formally be solved to give ΓG0 = (1 − IG0 )

−1

IG0

(10.48)

and to express G as

1 . (10.49) 1 − IG0 Besides the poles of G0 , the Green function G of the correlated electron–hole pair contains additional poles for 1 − IG0 = 0, which we may also write as G−1 0 − I = 0. For a free electron–hole pair in simple parabolic bands coupled by the Coulomb interaction, this expression can be written in plane wave representation as ′ ′ ′ ′ ke kh |G−1 (10.50) 0 − I|ke k h k e kh |r e r h = 0 , G = G0

k′e ,k′h

where φk′e k′h (r e , rh ) = k′e k′h |r e rh is the exciton wave function. This leads to ke kh |G−1 0

−

I|k′e k′h

! ¯ 2 ke2 h ¯h2 kh2 = Eg + + − E δke k′e δkh k′h 2me 2mh e2 ′ . δ (10.51) − ′ ε0 V |ke − k′e |2 ke −ke ,kh −k h

After introducing relative and center-of-mass coordinates (see (10.34)), taking the Fourier transform, and separating the center-of-mass motion we recover the exciton eﬀective-mass Hamiltonian of (10.36) with the bare Coulomb interaction. We identify the additional poles of the Green function G as the bound states of the electron–hole pair, the excitons. If instead of the plain wave representation the Bloch representation is used, one obtains for the inverse Green function of the uncorrelated electron–hole pair the expression ′ ′ ′ ′ cke vkh |G−1 0 |c k e v k v = (Ec (k e ) + Ev (k h ) − E) δcc′ δvv′ δke ,k′e δkh ,k′h .

(10.52)

It contains the full expressions of the energy bands but is diagonal in the band indices and in the electron and hole wave vectors.

304

10 Light–Matter Interaction

Let us now turn to the irreducible electron-hole interaction I of the BSE [288]. It is composed of graphs like those represented in Fig. 10.5. The left one is the bare Coulomb interaction as in Fig. 4.10 but now for the electronhole pair. It propagates in the Bloch states cke , vkh and is scattered into the pair c′ k′e , v ′ k′h by the Fourier component of the Coulomb interaction with wave vector q = k′e − ke = k′h − kh . Note, that due to time-inversion the propagation direction of the hole is opposite to that of the electron. The bare Coulomb interaction is modified if intermediate excitation of electron– hole pairs is considered. This is depicted in Fig. 10.6 and means to replace the simple broken Coulomb line of the direct interaction graph by the double broken line which contains the polarization diagrams. It represents the screened Coulomb interaction (see Chap. 4), mentioned already in our intuitive considerations. The exchange diagram on the right of Fig. 10.5 also contributes to the irreducible electron–hole interaction. It represents a scattering of the electron– hole pair by the Fourier component of the bare Coulomb potential, with wave vector Q = ke + kh = k′e + k′h . However, due to the topology of the exchange diagram, a replacement of the bare Coulomb line by the screened one of Fig. 10.6 would lead to diagrams, which could be separated into replicas of the simple one by cutting with a horizontal line (if we forget the band indices for the moment). Such diagrams are called reducible and are excluded. This means that in the simple two-band exciton problem the exchange interaction is not screened [288]. As will be outlined in the following supplement, this argument does not apply if one considers the full band structure instead of the two-band model.

+

<

=

<

Fig. 10.5. Diagrams for Coulomb and exchange interaction of an electron–hole pair

+ ....

Fig. 10.6. Diagrams showing the screening of the electron–hole interaction

10.3 Excitons

305

Supplement: Screened Coulomb and exchange interaction Let us start with the screened Coulomb interaction VCoul = −cke vkh | =−

d3 r

e2 |c′ k′e v ′ k ′h

κ|r − r ′ | ∗ d3 r ′ ψck (r)ψvk h (r ′ ) e

e2 ψc′ k ′e (r)ψv∗′ k ′ (r ′ ) , (10.53) h κ|r − r ′ |

where κ = 4πε0 ε and ε = ε∞ is the electronic part of the dielectric constant. Note, that the hole Bloch functions are complex-conjugate because of time-inversion. The products of the lattice periodic parts of the Bloch functions can be expanded in Fourier series with coeﬃcients 1 Ccc′ (ke , k′e , G) = u∗ck e (r)uc′ k ′e (r)eiG·r d3 r (10.54) Vcell ′ ′ 1 Vvv′ (kh , k ′h , G′ ) = uvk h (r ′ )u∗v′ k ′ (r ′ )e−iG ·r d3 r ′ (10.55) h Vcell to write the Coulomb interaction as Ccc′ (k e , k ′e , G)Vvv′ (kh , k′h , G′ ) VCoul = − G,G ′

×

d3 r

′

′

d3 r ′ e−i(k e −k e −G)·r ei(k h −k h −G

′

)·r ′

e2 . κ|r − r ′ |

(10.56)

The double integration can be performed after replacing the Coulomb potential by its Fourier transform, giving e2 ′ d3 r d3 r ′ · · · = δ ′ ′ . (10.57) ε0 εVc |k e − k ′e − G|2 k e −k e −G,k h −k h +G Let us consider the case with Q = 0, i.e., k e = k h = k. Then, for weakly bound (or Wannier–Mott) excitons, one has |k − k′ | ≪ 2π/a and the Fourier component of the Coulomb potential with G = 0 dominates. The corresponding expansion coeﬃcients Ccc′ (k, k′ , 0) and Vvv′ (k, k ′ , 0) reduce to δcc′ and δvv′ , respectively (see Sect. 9.1), and we arrive at the leading term of the screened Coulomb interaction VCoul = −

e2 δcc′ δvv′ , ε0 εVc |k − k ′ |2

(10.58)

which is diagonal in the band indices. For a given band pair with Ec (k) and Ev (k) in isotropic parabolic approximation, this leads back to the exciton problem of (10.49), but now with the screened Coulomb interaction. In order to treat the exchange interaction in the same way, we have to consider the spin part of the Bloch functions explicitly. According to the diagram rules, spin is conserved at each vertex. This has been tacidly taken into account when evaluating the Coulomb term, for which electron or hole lines meet at each vertex and summation over spin variables gives a factor unity. In the exchange diagram (right hand side of Fig. 10.5), however, at each vertex an electron line meets a hole line. As the hole wave function is a time-inverted electron Bloch function, its spin is opposite to that of the electron and the total spin S of the electron–hole pair has to

306

10 Light–Matter Interaction

be zero. Thus only spin-singlet states, S = 0, experience the exchange interaction. This consideration allows one to write the exchange term after summation over the spin variables as e2 ∗ (r)ψv′ k ′h (r ′ ) ψ ∗ (r)ψc′ k ′e (r ′ ). (10.59) Vexch = 2δS,0 d3 r d3 r ′ ψck e κ|r − r ′ | vk h We may again expand the product of the periodic parts of the Bloch functions (here that of a hole and of an electron wave function) in a Fourier series now with coeﬃcients 1 Wcv (ke , k h , G) = (10.60) u∗ck e (r)u∗vk h (r)eiG·r d3 r Vcell

and proceed as before. Using Q = k e + k h = k′e + k′h , this gives Vexch = 2δS,0

Wcv (k, Q, G)Wc∗′ v′ (k′ , Q, G)

G

e2 . ε0 Vc |Q − G|2

(10.61)

The center-of-mass wave vector Q is usually much smaller than a reciprocal lattice vector G and can be neglected except for G = 0. For the term with G = 0, we use the k · p expansion of the Bloch factors around k = 0, while for the terms with G = 0 the zone center Bloch functions are taken as a good approximations to write the leading contribution as & h2 (P cv · Q)(P v′ c′ · Q) ¯ e2 ) Vexch = 2δS,0 lim 2 Q→0 ε0 Vc Q m2 (Ec − Ev )(Ec′ − Ev ′ ' e2 + Wcv (0, 0, G)Wc∗′ v′ (0, 0, G) . (10.62) ε0 Vc G2 G=0

This result does not depend on k (or k ′ ), thus, after Fourier transformation, this exchange interaction is a contact potential ∼δ(r) in the relative coordinate. Moreover, in contrast to the Coulomb term, it is not diagonal in the band indices. It describes the coupling between the band pair c, v forming the lowest energy gap and the band pairs c′ , v ′ with larger energies. Therefore, the argument with the reducibility of exchange diagrams, which was correct for the simple two-band model, is not correct in the more general case, when the dielectric background represented by the band pairs with higher energy is taken into account. It can be considered by a matrix diagonalization procedure (partitioning) of the exchange interaction, which results in a screening of the exchange interaction by the dielectric background. In addition, the first term of (10.62), to be taken in the limit Q → 0, has the peculiar property, that it depends on the orientation of the exciton wave vector Q with respect to the dipole matrix element P cv and is, therefore, called nonanalytic exchange term. It contributes only to longitudinal excitons with Q parallel to the transition dipole P cv . This splitting between longitudinal and transverse excitons is analogous to that of optical phonons (see Chap. 3). We shall come back to this aspect in the following section. The second term in (10.62) is the analytic exchange term [290].

As outlined in this section, excitons are the electronic excitations with the lowest energy (usually in the optical regime) in semiconductors and insulators. For dipole-excitations with this energy, they represent the quanta of the

10.4 Polaritons

307

polarization field as is the case for optical phonons in the far-infrared regime (see Sect. 3.5). Their internal structure, a superposition of electron–hole pair excitations (see (10.37) and (10.38)) suggests to consider them as bosons with † to the electronic ground state |Ψ0 creoperators BνQ . Application of BνQ † ates an exciton according to BνQ |Ψ0 = |ΨνQ and the Hamiltonian for the electron system can be represented in this energy range as † Hel = EνQ BνQ BνQ , (10.63) ν,Q

with the exciton energy EνQ . Note, however, that the bosonic character of the exciton holds only in the approximation of low excitation densities (Problem 10.4). With increasing excitation densities the internal structure of the exciton, its composition of fermions, becomes relevant and is the origin of high-excitation phenomena like biexcitons (or exciton molecules), polyexcitons, electron–hole droplets, and formation of an electron–hole plasma. Bose–Einstein condensation of excitons is another topic in this field [283].

10.4 Polaritons Electro-magnetic waves propagate through an insulating solid with reduced velocity of light, c′ = c/n1 , according to the dispersion relation ω = c′ |κ|. In the frequency region of a dipole-active oscillator, c′ depends on ω due to anomalous dispersion. In a microscopic formulation this is due to the coupling of the propagating light with the oscillator, which, in the context of this section, will be an exciton, but the same considerations also hold for optical phonons. This coupling results in a new kind of excitation, the polariton or, to be more specific, the exciton polariton. As an introduction to the polariton concept, let us first follow the phenomenological approach. Light propagation in insulating matter can be described using the macroscopic Maxwell equations ∇×E = −

∂B ∂D , ∇×H = . ∂t ∂t

(10.64)

The properties of insulating matter are considered in the dielectric function ε(ω), which is determined by the dipole-active oscillators of the solid, here the excitons. It connects the electric field E with the displacement vector D, which is connected with the dielectric polarization P , according to D = ε0 εE = ε0 E + P .

(10.65)

The dielectric polarization P (ω) = χ(ω)E(ω) depends on the dielectric susceptibility χ(ω), with contributions of the possible dipole excitations in the solid.

308

10 Light–Matter Interaction

The fields B and H can be eliminated by combining the two macroscopic Maxwell equations (10.64). One obtains ∇ × (∇ × E) = −

1 ∂2D . ε0 c2 ∂t2

(10.66)

Assuming monochromatic light with the electric field given by E(r, t) = E(ω)ei(κ·r−ωt) + c.c.

(10.67)

one finds

ω2 εE . (10.68) c2 For transverse waves, E ⊥ κ, the lhs reduces to κ2 and we can eliminate the common factor E. The resulting relation κ × (κ × E) =

ε(ω) =

c2 2 κ ω2

(10.69)

determines, for given ε(ω), the dispersion of the coupled light–matter modes, the polaritons as shown in Fig. 10.4 as plots of ω vs. real (κ1 ), and imaginary part (κ2 ) of the wave vector. The frequency dependence of ε(ω) = 1 + χ(ω)/ε0 results from the oscillators connected with dipole excitations of the solid (here, the excitons). A simple model, assuming a single oscillator with eigenfrequency ω0 and oscillator strength f0 , would give χ(ω) = f0 /(ω02 − ω 2 ). At frequencies ω much smaller than ω0 , the propagation follows the linear relation # for light in vacuum, but with a reduced velocity c′ = c/n< with n< = 1 + f0 /ε0 ω02 . For ω ≫ ω0 , the dielectric susceptibility does not contribute and the propagation takes place with c′ = c/n> , where n> = 1. If besides the oscillator at ω0 , other oscillators contribute to χ(ω) but at frequencies ω ≫ ω0 , they can be considered by using ε(ω) = εb + χ(ω)/ε0 , with a background dielectric constant εb , which replaces 1 in the expressions for the refractive indices n< and n> . For a longitudinal wave with E κ, the lhs of (10.68) vanishes and the longitudinal frequency ωL follows from ε(ωL ) = εb +

f0 =0 ε0 (ω02 − ωL2 )

(10.70)

to be ωL2 = ω02 + f0 /ε0 εb , and we recover the Lyddane–Sachs–Teller relation of Chap. 3 ωL2 f0 =1+ . (10.71) ω02 ε0 εb ω02 It can be used to express the dielectric function in terms of the frequencies ω0 and ωL to arrive at the relation κ2 = ε b

ω 2 ωL2 − ω 2 , c2 ω02 − ω 2

(10.72)

10.4 Polaritons

ω

UPB

309

c/n > c/n <

ωL ω0 LPB

κ1

κ2

Fig. 10.7. Polariton dispersion: solutions of (10.69) vs. real and imaginary part of κ

which determines the polariton dispersion ω(κ). The diﬀerence between the resonance frequency ω0 of the oscillator (which is a transverse excitation) and the longitudinal frequency ωL defines the longitudinal-transverse splitting (LT-splitting) of the excitonic resonance, mentioned already in the previous Section. It follows from the Lyddane–Sachs–Teller relation (10.71) by using ω0 + ωL ≈ 2ω0 , and is given by ∆LT = ωL − ω0 =

f0 . 2ε0 εb ω0

(10.73)

For ω < ω0 and ω > ωL we have κ2 > 0 and find two real solutions, the lower and the upper polariton branch, LPB and UPB, respectively in Fig. 10.7. For the frequency intervall ω0 < ω < ωL κ2 is negative and leads to a damped solution. The polariton dispersion curves are shown in Fig. 10.7 in a plot of ω vs. the real and imaginary parts of κ = κ1 + iκ2 . It should be noted, that the expression for ∆LT considers the dielectric background by the phenomenological constant εb . Its microscopic origin is the coupling of the exciton with the excitations between band pairs with higher energy. With the arguments given in Sect. 10.3 it is possible to relate the LT-splitting with the oscillator strength of the exciton (Problem 10.5). In Fig. 10.8, we show the polariton dispersion calculated with the exciton parameters of CuCl for a small interval of the wave vector close to the center of the Brillouin zone. (Note, that the Brillouin zone extends out to k ≈ 108 cm−1 .) The experimental data points are obtained from two-photon absorption (TPA) and hyper-Raman scattering (HRS). In the former case, two photons from diﬀerent sources are simultaneously absorbed to excite the exciton–polariton, while in the latter a virtually excited biexciton decays into two polaritons, one of which is detected as a scattered photon outside the solid.

310

10 Light–Matter Interaction 3.22

CuCi

UPB

(eV)

long.

h ωL 3.20

LPB

HRS - data TPA - data theory

3.18

3.16 0

5

10

15

k (105 cm–1)

Fig. 10.8. Calculated exciton–polariton dispersion of CuCl with experimental data from two-photon absorption (TPA-data) and hyper-Raman scattering (HRS-data), after [283] 3.22 UPB

UPB

E (eV)

3.21 LPB

3.20

LPB

CuCI 2K

3.19

(a) 3.18

10 10–5

0 k (x

106

cm–1)

(b) 10–4

10–3

10–2

10–1

vg / c

Fig. 10.9. Exciton–polariton dispersion of CuCl (left) and measured propagation velocity of polariton modes in CuCl together with values calculated from the dispersion curve (right), after [283]

In contrast to simple absorption experiments with one photon, both methods allow to map out the dispersion of the polaritons by exploiting diﬀerent configurations of the small wave vectors of incident and scattered photons. It is also possible to detect the longitudinal polariton branch. Fig. 10.9 shows, in the left hand part, the dispersion of the transverse polariton branches on a larger k scale. On this scale, the dispersion of the lower polariton branch (LPB), caused by the free center-of-mass term in the exciton energy, becomes relevant. In the right hand part of the figure the group velocity vg , calculated from the polariton dispersion, is compared with experimental data obtained from time-of-flight measurements. Away from the exciton resonance, the light pulses propagate through the sample with the velocity of light in matter, the slope of the steep linear parts of the upper

10.4 Polaritons

311

(UPS) and lower polariton branches (LPB). However, close to the resonance the group velocity is reduced by orders of magnitude, because the travelling pulse polarizes the matter by coupling with the exciton. After this phenomenological survey of exciton–polaritons we turn now to the rigorous formulation that makes use of the quantization of the electromagnetic field and treats light–matter interaction as exciton–photon coupling. The quantization of the electro-magnetic field can be performed analogous to the quantization of the displacement field in Chap. 3 (leading to phonons) with the only diﬀerence, that in the Coulomb gauge the vector potential is transverse (Problem 10.6). With the expansion A(r, t) =

λ,κ

¯ h 2ε0 ω(κ)V

!1/2

eλκ aλκ (t) + a†λκ (t) eiκ·r

(10.74)

in terms of Boson operators aλκ for photons, the quanta of the electromagnetic field, and transverse unit vectors eλκ with eλκ · κ = 0, the Hamiltonian of the radiation field takes the expected form ! 1 Hrad = (10.75) hω(κ) a†λκ aλκ + ¯ 2 λ,κ

with photon energies ¯hω(κ). At the end of Sect. 10.3 (in the energy range of excitonic resonances) we have formulated the Hamiltonian for the electron system in terms of exciton operators BνQ . We want to extend this concept also to the light–matter interaction. The Ap-coupling of Hel−rad is linear in the vector potential, which can be replaced by its quantized form (10.74), and it contains the electron momentum operator pl . According to (4.76) it can be expressed in terms of Fermion operators giving HAp =

nn′ kk′ λ,κ

e m

¯ h 2ε0 ω(κ)V

!1/2

× aλκ + a†λ−κ c†n′ k′ cnk ,

n′ k′ |eiκ·r p · eλκ |nk (10.76)

which has the same form as the electron–phonon coupling (8.8). However, our concern is not the coupling to an individual electron–hole pair but to an exciton, and we should express the electron part in terms of exciton operators. If we simplify to an exciton formed with Bloch functions from the band pair c, v and with the wave function φ(r) for the relative motion, then the coupling to one of the photons can be written (Pcv = c|e · p|v) HAp =

e m κ

¯ h 2ε0 ω(κ)V

!1/2

† , (10.77) |φ(0)||Pcv | aκ + a†−κ Bκ + B−κ

312

10 Light–Matter Interaction

which is linear in the Boson operators of both photons and excitons and we may write the polariton Hamiltonian as Hpol =

1 Eκ Bκ† Bκ + h ¯ ω(κ) a†κ aκ + 2 κ † † + a†−κ Bκ + a†−κ B−κ + Cκ aκ Bκ + aκ B−κ

(10.78)

The coupling reminds of the situation found for anti-ferromagnetic magnons in Sect. 6.4 and can be exactly eliminated by looking for solutions with polariton operators † ακ = A1 aκ + A2 Bκ + A3 a†−κ + A4 B−κ .

(10.79)

For the detailed calculation we refer to the literature [10, 14] and to the original paper by Hopfield [291]. The eigenfrequencies of the polaritons are those found with the phenomenological approach and show the anti-crossing behavior (Fig. 10.7), which is typical for hybridization (see Sect. 5.4). The lower polariton branch away from the exciton resonance is photon-like and follows the linear dispersion, but turns over into the exciton dispersion and becomes exciton-like. The upper polariton branch starts at the longitudinal exciton energy but approaches for higher energy the linear dispersion of photons. The existence of polaritons, as coupled exciton–photon modes is clearly demonstrated by the quoted experiments. Thus, photons propagate through the solid as polaritons, which convert to photons when passing back into the vacuum. This means, that absorption is not the conversion of a photon into an electronic excitation like the exciton but due to polariton scattering (e.g., with phonons or impurities), by which energy is dissipated into other excitations of the solid.

10.5 Light-Scattering Light-scattering, or more precisely inelastic light scattering, has, since its discovery 1928 by Raman,3 gained much importance as a spectroscopical method, especially with the availability of lasers as intense light sources. A comprehensive overview of the method, its applications, and the theoretical concepts is given in a series of books with the topic Light Scattering [118]. A schematic view, typical for all scattering experiments, is shown in Fig. 10.10. The incident monochromatic light (vector potential A0 (r, t)), with photon energy h ¯ ω0 and photon momentum h ¯ κ0 , is scattered at a sample (here the solid), and the cross section for the scattered light (vector potential A1 (r, t)), with photons of energy h ¯ ω1 and momentum ¯hκ1 , is measured. It depends on the energy and momentum transfer 3

Chandrasekhara Venkata Raman 1888–1970, Nobel prize in Physics 1930.

10.5 Light-Scattering

313

Fig. 10.10. Schematic view of the light-scattering experiment

hω = h ¯ ¯ ω0 − ¯ hω 1 , ¯ hq = h ¯ κ0 − ¯hκ1 ,

(10.80)

which, under the assumption of weak light–matter interaction, i.e., validity of the Born approximation, provides information about an elementary excitation with dispersion ω(q) in the sample. The light–matter interaction can be described by the Hamiltonian Hel−rad of (10.13), with the vector potential A(r, t) = A0 (r, t) + A1 (r, t). The cross section depends on both the incident and scattered radiation fields, therefore, we expect contributions linear in the product A0 A1 not only from the first term of Hel−rad (Ap -coupling) in second order but also from the second term, which already contains this product, in first order perturbation theory (A2 -coupling). Fermi’s Golden Rule applied to the A2 -coupling N

HA0 A1 =

e2 A0 A1 e0 · e1 ei(ω0 −ω1 )t ei(κ0 −κ1 )·rl + h.c. m

(10.81)

l=1

allows to write down immediately the transition probability !2 " "2 2π e2 A0 A1 " " eiq·rl |i" δ(Ef − Ei ± ¯hω) (10.82) |e0 · e1 |2 "f | Wi→f = h ¯ m l

and the diﬀerential scattering cross section !2 d2 σ h ¯ ω1 = Wi→f dΩdω 2π ω0 A0 A1 i,f !2 4 e ω1 2 2 |e · e | |f |N−q |i| δ(Ef − Ei ± ¯hω). (10.83) = 0 1 ω0 m2 i,f

As can be seen from this expression, thescattering due to the A2 -coupling is caused by number fluctuations N−q = l exp(iq · rl ) (see Sect. 4.5) and can be related with the dynamic structure factor S(q, ω) = |f |N−q |i|2 δ(ωf − ωi ± ω) i,f

=

1 2π

+∞

−∞

e±iωt Nq (t)N−q (0)dt,

(10.84)

314

10 Light–Matter Interaction

Ef ,h(k+q)

>

>

>

>

Ei ,hk

-hω ,hκ 0 0

-hω ,hκ 1 1

Fig. 10.11. Feynman diagram for the light-scattering process with the A2 -coupling

S

AS ____ d 2σ dωdΩ

-hω 0 |i

anti-Stokes

Stokes

∆ω= 0

∆ω

-hω 1

-hω 1 |f

-hω 0

|f

|i

Fig. 10.12. Schematic picture of the scattering cross section and of the diagram of the involved energy levels. The peak of the cross section at ∆ω = 0 is due to the unscattered light, while the satellites are due to stokes (S) and anti-stokes (AS) scattering

where Ei,f = h ¯ ωi,f , or expressed by the energy–loss function. Thus, lightscattering, employing the A2 -coupling is capable to investigate collective modes of the solids like charge or spin density excitations. We note in passing that for inelastic electron scattering the same expression holds if e2 |e0 · e1 |/m is replaced by the Coulomb interaction vq (see (4.131). The scattering process can also be visualized in a Feynman diagram (Fig. 10.11). The vertex of the diagram represents the number fluctuation. A schematic view of the scattering cross section and the energy levels involved in the scattering process is shown in Fig. 10.12. The central peak at ∆ω = ω − ωf + ωi = 0 is caused by the unscattered laser light, the peaks at positive (negative) ∆ω, called Stokes (anti-Stokes) line, result from creation (annihilation) of an elementary excitation in the solid. The Ap -coupling e HAp = e0 A0 ei(κ0 ·r−ω0 t) + e1 A1 ei(κ1 ·r−ω1 t) + h.c. (10.85) m l

contributes in second order perturbation theory by terms linear in both vector potentials of incident and scattered light. The cross section essentially has

10.5 Light-Scattering

315

the same structure as the one for the A2 -coupling but e0 · e1 is replaced by ↔ e1 ·δ χ f i ·e0 under the sum over initial and final states. Here, the components ↔ of the Raman tensor δ χ f i are given by µ ν M1,f z M0,zi

µν

f |δχ |i ∼

z

Ez − Ei − ¯ hω

+ (0 ↔ 1).

In dipole approximation, the matrix elements µ ν M1,f e−iκ1 ·rl plµ |z and M0,zi = z| eiκ0 ·rl plν |i z = f | l

(10.86)

(10.87)

l

reduce to the momentum matrix elements f |plµ |z and z|plν |i, respectively, if the corresponding transitions are dipole allowed. In the foregoing, the states |i, |z, |f are to be understood as exact eigenstates of the solid, which are not known, but can be described approximately with the concepts developed in the earlier chapters. Let us assume the Hamiltonian of the solid in the form H = Hel + Hph + Hel−ph for electrons, phonons, and electron–phonon interaction as in Chap. 8. The eigenstates |γ of Hel +Hph are products of electron and phonon states in occupation number representation, while the eigenstates of H can be considered as quasi-particle states with energies Eα + iΓα , with finite lifetime due to the interaction Hel−ph . The exact intermediate state |z can be expressed by the perturbation expansion |z = |γ +

|γ ′ γ ′ |Hel−ph |γ γ′

Eγ − Eγ ′

+ ...

(10.88)

and one obtains contributions to f |δχµν |i of first order in Hel−ph µν

f |δχ |i ∼

γ,γ ′

Mfµγ Mγγ ′ Mγν′ i

hω1 − Eγ − iΓγ ) (¯hω0 − Eγ ′ − iΓγ ′ )(¯

+ (0 ↔ 1) ,

(10.89)

where Mγγ ′ = γ|Hel−ph |γ ′ . A graphical picture of such a contribution is shown in Fig. 10.13 together with the energy level scheme. The incoming photon creates via Ap -coupling an electron-hole pair or exciton (a), which emits (or absorbs) a phonon (b) and recombines via Ap -coupling to the scattered photon (c). Higher order processes with Hel−ph are possible. The process described here in detail is used to investigate phonons in bulk material, quantum wells, or at surfaces. If optical phonons are involved, it is called Raman scattering, while Brillouin scattering is due to emission or absorption of acoustic phonons. Instead of the electron–phonon interaction as the origin of the inelastic light-scattering, one could have also considered the coupling between the electron system and other collective excitations like plasmons, plasmon–phonon modes, or magnons to map out dispersion. The energy denominators in the scattering cross section can be exploited to enhance the scattering eﬃciency by using the possible resonances of incident or scattered light with the electronic excitation (resonant light-scattering).

316

10 Light–Matter Interaction

E (a)

>

-h ω ,hκ 0 0

(b)

(b) > (c)

>

-hω ,hκ 1 1

(c)

(a)

exciton Q

Fig. 10.13. Feynman diagram and level diagram for the light-scattering process with Ap and electron–phonon coupling

10.6 Coherent Interband Dynamics The light–matter interaction has been treated in the previous sections of this chapter either as a time-dependent perturbation of the electron system on a long time scale by applying Fermi’s Golden Rule or by creating the instanteneous formation of polaritons as coupled modes of electronic excitation and electro-magnetic field. In this section, we focus on the dynamics of the interband transitions by looking at the light–matter system as a system of driven oscillators or two-level systems on the very short time scale after excitation, for which the polarization is in phase with the radiation field. Experimentally, the investigation of this coherent regime is possible with spectroscopy using light pulses in the ps and fs range. The approach of this and the following section, which, besides light–matter interaction also includes many-body aspects, provides a very general view, because previous results of this chapter can be recovered as special cases[95, 284, 285]. Let us start with the Hamiltonian H = H0 + HAp + HCoul .

(10.90)

We adopt a semiconductor in the form of a two-band model with valence and conduction band (n = c, v) separated by a direct gap at k = 0 and formulate the single-particle part of the system Hamiltonian in second quantization: H0 = En (k)c†nk cnk . (10.91) n=c,v k

The light–matter interaction is reduced to the Ap -coupling written as dcv (k) · E(κ, t)c†ck cvk−κ + h.c. , HAp = − (10.92) k,κ

with the dipole matrix element dcv (k) = eck|r|vk for the transition between the valence and conduction band states at k and the (real) electric field vector E(κ, t) for a harmonic wave with wave vector κ (see Problem 10.1). While the interaction between all electrons in the filled valence band is considered in the energies En (k) of the Bloch states (in the vein of DFT discussed in

10.6 Coherent Interband Dynamics

317

Chap. 5), we have to explicitly consider the Coulomb interaction between the carriers created by the primary excitation HCoul =

1 & † vq cck+q c†ck′ −q cck′ cck + c†vk+q c†vk′ −q cvk′ cvk 2 k,k′ q=0

' + 2c†ck+q c†vk′ −q cvk′ cck .

(10.93)

Further simplifications can be made with the dipole approximation, according to which the spatial variation of the electric field can be ignored with κ = 0, and by considering electron–hole pair transitions only close to the fundamental gap at k = 0, which all have the same dipole matrix element dcv . For an isotropic semiconductor and linearly polarized light we may simplify dcv ·E(t) to the scalar form dcv E(t) and write the light–matter coupling in the form HAp = − (10.94) E(t) dcv c†ck cvk + d∗cv c†vk cck . k

Let us first study the kinetics of optical interband transitions without the Coulomb interaction. This leaves us with a single-particle problem, which can be solved by looking at the equations of motion for the operators c†ck cck , c†vk cvk , c†ck cvk , and c†vk cck . In the Schroedinger picture these are time-dependent operators, their thermal expectation values ρnn′ (k, t) = c†nk (t)cn′ k (t) are elements of the time-dependent density matrix ! ρcc (k, t) ρcv (k, t) (10.95) ρk (t) = ρvc (k, t) ρvv (k, t) of a two-level system at k. We switch to the interaction picture with ¯ = e h¯i H0 t Oe− h¯i H0 t O

(10.96)

and write down the equation of motion for the density matrix d i ¯ ρ¯k (t) = − H ¯k (t) , Ap (t), ρ dt h ¯

(10.97)

which for its elements have the following form (here ǫnk = En (k)/¯h): d ρ¯cv (k, t) = dt d ρ¯cc (k, t) = dt d ρ¯vv (k, t) = dt

i dcv E(t)ei(ǫck −ǫvk )t (¯ ρvv (k, t) − ρ¯cc (k, t)) ¯ h i E(t) dcv ei(ǫck −ǫvk )t ρ¯vc (k, t) − c.c. h ¯ i E(t) d∗cv e−i(ǫck −ǫvk )t ρ¯cv (k, t) − c.c. . h ¯

(10.98) (10.99) (10.100)

318

10 Light–Matter Interaction

The diagonal elements are identical in the Schroedinger and interaction pictures, ρnn (k, t) = ρ¯nn (k, t), and their time derivatives are identical up to a sign change, dρvv (k, t)/dt = −dρcc (k, t)/dt, as can be seen from (10.99) and (10.100). These equations describe the interband kinetics of free carriers at k. It will be shown later, that excitations at diﬀerent k are coupled due to HCoul . Let us briefly assume quasi-thermal equilibrium, i.e., initially created electrons in the conduction band and holes in the valence band are thermalized within their respective bands. Then, the diagonal elements of the density matrix are Fermi–Dirac distribution functions fck and fvk , respectively. For this case, the calculation of the optical polarization P (t) = Tr (ρ(t)d), where d is the electric dipole operator, yields the optical susceptibility χ(ω) and the expression for the dielectric function ε(ω) derived in Sect. 10.2 (Problem 10.7). The more interesting case is the time scale, on which the interband excitation follows the same time dependence as the driving electric field, the coherent regime. Let us look with E0 iωt E(t) = e + e−iωt , (10.101) 2 at the equation of motion (10.98) for the oﬀ-diagonal element of the density matrix ρ¯cv (k, t) = ei(ǫck −ǫvk )t ρcv (k, t) , (10.102) which takes the form ! idcv d + i(ǫck − ǫvk ) ρcv (k, t) = E(t) (ρvv (k, t) − ρcc (k, t)) . dt h ¯

(10.103)

The condition of almost resonant excitation with the detuning νk = ǫck −ǫvk − ω is considered in the so-called rotating wave approximation by dropping the term with exp(iωt) in E(t), which leads to ! d iωR + iνk ρcv (k, t)eiωt = − (ρcc (k, t) − ρvv (k, t)) dt 2 d d iωR ρcv (k, t)eiωt − c.c. , (10.104) ρcc (k, t) = − ρvv (k, t) = − dt dt 2 h. This set of equations is known as with the Rabi4 frequency ωR = dcv E0 /¯ coherent optical Bloch equations with reference to the Bloch equations used in magnetic resonance spectroscopy. In order to show this relation, the complex elements of the density matrix can be used to define three real components of a vector, the Bloch vector

4

U1 (k, t) = ρcv (k, t)eiωt + c.c. U2 (k, t) = i ρcv (k, t)eiωt − c.c. U3 (k, t) = ρcc (k, t) − ρvv (k, t) , Isaac Isidor Rabi 1898–1988, Nobel prize in Physics 1944.

(10.105)

10.6 Coherent Interband Dynamics

319

whose time-dependence is ruled by d U1 (k, t) = −νk U2 (k, t) dt d U2 (k, t) = +νk U1 (k, t) + ωR U3 (k, t) dt d U3 (k, t) = −ωR U2 (k, t) , dt

(10.106)

which can also be written also in the compact form as d U (k, t) = Ω × U (k, t), dt

(10.107)

with Ω = −ωR e1 + νk e3 . An equation of this form has been derived already in Sect. 6.3 for the dynamics of spins; its mechanical analogue is the equation of motion for an angular momentum under the action of a torque. Under resonant excitation, νk = 0, the Bloch equation describes the rotation of the Bloch vector around the −e1 -axis with the Rabi frequency ωR . Let us start at t = 0 with a population inversion described by U (t = 0) = (0, 0, −U3 ). After a half period, one has U (t = π/ωR ) = (0, 0, U3 ), i.e., the initial population inversion has changed its sign. This is called Rabi flopping. At the intermediate time, π/2ωR U3 is completely converted into the component U2 of the Bloch vector, which is connected with the dielectric polarization P (t). During the rotation caused by the coupling to the radiation field, the Bloch vector changes periodically between the components U3 and U2 , or between population inversion and interband polarization, respectively. The resonant condition cannot be fulfilled at the same time for interband transitions at diﬀerent k, and the detuning νk = 0 becomes important. Let us consider a group of electron–hole pairs at diﬀerent k, but initially with same Bloch vector U (1) = (0, 0, U3 ). A short light pulse of duration t1 = π/2ΩR (so-called π/2 pulse) turns these Bloch vectors into the 2-direction, U (2) = (0, U2 , 0), and they start to precess around the 3-direction according to their individual detunings, i.e., the Bloch vectors run out of phase and spread in the U1 − U2 plane. This spreading can be reversed by applying, after some time, T , a π pulse (duration t2 = π/ΩR ), which kicks the Bloch vectors into the directions U (3) = (U1 , −U2 , 0) such that after 2T their further precession around the 3-direction brings them all back to the same phase with U (4) = (0, −U2 , 0). This Bloch vector, which again is the same for all electron–hole pairs, can be detected as emitted light pulse, the photon echo. The photon echo can be observed only, if the phase coherence of each Bloch vector (or electron–hole excitation) with the exciting light is not destroyed. The coherent optical Bloch equations exhibit some principle features in the dynamics of interband excitations, which, in reality, will be changed due to the finite lifetime of the single-particle states and destroy phase coherence. This can be considered by adding phenomenological damping terms to

320

10 Light–Matter Interaction

the Bloch equations U1 (k, t) d U1 (k, t) = − − νk U2 (k, t) dt T2 U2 (k, t) d U2 (k, t) = − + νk U1 (k, t) + ωR U3 (k, t) dt T2 U3 (k, t) + 1 d U3 (k, t) = − − ωR U2 (k, t) . dt T1

(10.108)

The two time constants can be distinguished by their geometrical relation to the components of the Bloch vector: the longitudinal relaxation time T1 (for U3 ) is related with the diagonal elements of the density matrix, while the transverse relaxation time T2 (for the components U1 and U2 ) refers to its oﬀdiagonal elements. Their physical content discloses by looking at the meaning of the components of the Bloch vector: U3 describes the population inversion of the two-level system, which decays with T1 , the population lifetime, while U1,2 , describing the phase-coherent polarization of the two-level system, decay with the phase relaxation, dephasing or decoherence time T2 . Dephasing takes place not only by scattering processes with phonons, lattice defects, and other electronic excitations, but also by electron–hole recombination. If, besides the latter, which is the only process contributing to T1 , no other processes contribute to T2 , then T2 = 2T1 , because the number of electron–hole pairs is proportional to P 2 ∼ exp(−2t/T2). Including all other processes leads to the relation T2 ≤ 2T1 .

10.7 Semiconductor Bloch Equations In order to include the particle interaction HCoul it is convenient to introduce the electron–hole picture by replacing the Fermion operators according † . With these modifications, making use of to cck → αk and cvk → β−k En (k) = En (−k) and replacing under the sum k by −k, the terms of the system Hamiltonian can be written as Ec (k)α†k αk + Ev (k)βk βk† (10.109) H0 = k

HCoul

HAp

1 = 2

k,k′ ;q=0

vq α†k+q α†k′ −q αk′ αk + βk−q βk′ +q βk† ′ βk†

+ 2α†k+q βk′ +q βk† ′ αk † dcv α†k β−k = −E(t) + h.c. . k

(10.110) (10.111)

10.7 Semiconductor Bloch Equations

321

Bringing the hole operator terms into normal order one finds with † vq βk−q βk′ +q βk† ′ βk† = vq βk+q βk† ′ −q βk′ βk k,k′ ;q=0

k,k′ ,q =0

+2

q=0 k

βk† βk −

vq (10.112)

k;q=0

and single-particle energies Ee (k) = Ec (k) for electrons and Eh (k) = −Ev (k) + q=0 vq for holes H0 + HCoul =

1 Ee (k)α†k αk + Eh (k)βk† βk − vq 2 q=0 k 1 † βk† ′ −q βk′ βk vq α†k+q α†k′ −q αk′ αk + βk+q + 2 ′ k,k ;q =0 (10.113) − 2α†k+q βk† ′ −q βk′ αk .

In the first part of this section, we have investigated the dynamics of the two-level system with an electron in the conduction band and a hole in the valence band by taking into account only the single-particle terms in the equation-of-motion of the density matrix. Now, we also consider the contribu′ tion of the particle interactions HCoul represented by the four-operator terms in (10.113). This is done in three steps. The first step is the calculation of the † ′ and the operators α†k αk , βk† βk , and α†k β−k (Probcommutators between HCoul lem 10.8). In the second step, we replace the operator terms by their thermal expectation values. The two-operator terms α†k αk = fek and βk† βk = fhk have the obvious meaning of electron and hole populations at wave vector † k, while α†k β−k = Pk∗ is the macroscopic polarization due to the applied external field. With the results of Problem 10.8, the equations of motion of these quantities take the following form [284]: Ee (k) + Eh (k) ∗ ∂ ∗ P (t) = i Pk (t) + id∗cv E ∗ (t) fek (t) + fhk (t) − 1 ∂t k h ¯ † † +i − α†k α†k′ +q αk′ β−k−q vq α†k−q α†k′ +q αk′ β−k k′ ;q =0

† † , βk† ′ +q βk′ − α†k+q βk† ′ −q βk′ β−k + α†k β−k−q

(10.114)

∂ fek (t) = −2Im dcv E(t)Pk∗ (t) ∂t vq α†k α†k′ −q αk−q αk′ − α†k+q α†k′ −q αk αk′ +i k′ ;q =0

+ α†k α†k−q βk† ′ −q βk′ − α†k+q αk βk† ′ −q βk′ ,

(10.115)

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10 Light–Matter Interaction

∂ fhk (t) = −2Im dcv E(t)Pk∗ (t) ∂t † † † βk† ′ −q β−k βk′ βk′ −q β−k−q βk′ − β−k+q vq β−k +i k′ ;q=0

† † β−k . β−k+q − α†k′ +q αk′ β−k−q + α†k′ +q α†k′ β−k

(10.116)

In the last step we factorize the expectation values of the four-operator terms into expectation values of two-operator terms, following the concept of the Hartree–Fock approximation (see Sect. 4.4). Only those products are considered whose factors are diagonal in the wave vector indices and thus, represent macroscopic expectation values. Let us take as an example: † † α†k−q α†k′ −q αk′ β−k = −α†k−q αk′ α†k′ −q β−k

† + α†k′ −q αk′ α†k−q β−k

(10.117)

The two terms can be factorized into products of macroscopic expectation values appearing in the equations of motion to give † † α†k−q α†k′ −q αk′ β−k ≃ −δk−q,k′ α†k−q αk−q α†k β−k fek−q

† , + δq,0 α†k′ αk′ α†k β−k

Pk∗

(10.118)

where the last term does not contribute under the sum over q because q = 0 is excluded. In the same way, one finds (10.119) α†k α†k′ −q αk−q αk′ ≃ δk,k′ − δq,0 fek (t)fek′ (t) ,

where the second term does not contribute under the sum over q = 0. The neglected terms represent collision terms, while the remaining ones give the following set of equations

∂ Pk (t) = − i ǫek + ǫhk Pk (t) ∂t 0 fek (t) + fhk (t) − 1 / dcv E(t) + −i vq Pk−q (t) (10.120) h ¯ q =0 &/ 0 ' ∂ 2 fek (t) = − Im dcv E(t) + vq Pk−q (t) Pk∗ (t) (10.121) ∂t h ¯ q =0

∂ = fhk (t) . ∂t

(10.122)

Here ¯hǫik = En (k) + Σi (k), i = e, h are the renormalized single-particle energies with the self-energy Σi (k) = − v|k−q| fiq . (10.123) q

Problems

323

We introduce the generalized Rabi frequency ωR,k (t) =

1 dcv E(t) + v|k−q| Pq (t) , h ¯

(10.124)

q =k

which combines the applied field with the dipole field of the generated electron–hole pairs, and arrive at the semiconductor Bloch equations: ∂ Pk (t) = −i (ǫek + ǫhk ) Pk (t) − i (fek (t) + fhk (t) − 1) ωR,k (t) (10.125) ∂t ∂ ∂ fek (t) = −2Im (ωR,k (t)Pk∗ (t)) = fhk (t). (10.126) ∂t ∂t For vq → 0, i.e., when switching oﬀ the Coulomb interaction, these equations reduce to the optical Bloch equations. In spite of their simple form, these equations describe the rather complex dynamics of electron–hole pair excitations in that they include many-body eﬀects within the Hartree–Fock approximaton. This results in the renormalization of the single-particle energies and of the Rabi frequency, but (with the population inversion factor) takes into account also the filling of the phase space by the occupation of the single-particle states due to the light–matter interaction. The occupied states (electrons in the conduction band and holes in the valence band) reduce the available phase space for further excitations because of the Pauli principle. This phase space filling is known as the Pauli blocking. The semiconductor Bloch equations and their extensions, including collision terms, are prerequisite in modelling nonlinear susceptibilities to describe data from time-resolved spectroscopy or in designing semiconductor lasers. Damping terms with longitudinal and transverse relaxation times T1,2 can be added as in the optical Bloch equations and part of their microscopic origin can be calculated by considering the collision terms.

Problems 10.1 Show with the help of p/m = −i/¯ h[r, H0 ] that, in matrix representation with eigenstates of H0 , the Ap -coupling term can be expressed (for resonant excitation) as dcv · E with the dipole matrix element dcv = −ec0|r|v0 between band edge Bloch functions. 10.2 Make use of the Wannier representation (see Problem 6.2) to show that the momentum matrix element between Bloch functions is diagonal in the wave vector. 10.3 Make use of the solution of the hydrogen problem to express |φν (0)|2 in terms of a confluent hypergeometric function for positive and negative energies. Find the dependence of this expression for the bound states

324

10 Light–Matter Interaction

in dependence on the quantum number n and calculate the absorption coeﬃcient for the electron–hole continuum. † 10.4 Formulate the exciton operators Bν,Q , Bν,Q in terms of products of fermion operators for electrons in the conduction and valence band. Calculate the commutator between the exciton operators and define the condition under which excitons can be considered as bosons. 10.5 Start with the phenomenological result of (10.42) close to an excitonic resonance at Eν0 = h ¯ ων0 and express the LT-splitting of this exciton in terms of its oscillator strength! Consider the first term of the exchange interaction in (10.62) to calculate the LT-splitting as a perturbation correction to the (transverse) exciton energy Eν0 and verify by comparison with the phenomenological result the expression for the exciton oscillator strength given in (10.43). 10.6 Make use of the expansion A(r, t) =

λ,κ

¯ h 2ε0 ω(κ)V

!1/2

eλκ aλκ (t) + a†λκ (t) eiκ·r

(10.127)

of the vector potential in terms of Boson operators aλκ (t), where eλκ , λ = 1, 2 are transverse unit vectors with eλκ · κ = 0, to show that the Hamiltonian of the radiation field Hrad takes the form ! 1 . (10.128) Hrad = hω(κ) a†λκ aλκ + ¯ 2 λ,κ

10.7 Make use of the equation of motion for the oﬀ-diagonal element of the density matrix under quasi-equilibrium condition to calculate the dielectric polarization P (t) = Tr(ρcv (k, t)dcv ) and the dielectric susceptibility χ(ω). Recover the result for the imaginary part of the dielectric function ε2 (ω) obtained in Sect. 10.2. ′ with the operators α†k αk , βk† βk and 10.8 Calculate the commutators of HCoul † α†k β−k to verify (10.114 . . . 116).

A Appendices

A.1 Elements of Group Theory Geometrical operations (translation, rotation, inversion), which leave a geometrical object (here, the crystal lattice) invariant, are symmetry operations. Mathematically, they form a group, the symmetry group of the crystal: for the translations it is the translation group, for the rotations, inversion, and their combinations it is the point group. The number g of elements in the group is its order. The symmetry of a system implies the invariance of the system Hamiltonian H (be it for phonons, electrons, or magnons) under unitary operations corresponding to the geometrical operations of the symmetry group. These unitary operations form a group which is isomorphic to the symmetry group. When applied to a set of eigenfunctions of H, this set is transformed into another set of eigenfunctions, which can be represented as a linear combination of the former ones. The eigenfunctions of a degenerate eigenvalue transform among each other and form an invariant subspace in the Hilbert space of H. In a chosen basis, these unitary operations can be formulated as matrices which define another group isomorphic to the symmetry group. For a proper choice of the basis, all matrices of the matrix representation have block-diagonal form with the dimension of the block matrices indicating the degeneracy of the invariant subspaces. These subspaces, spanned by the set of degenerate eigenfunctions, can be classified by characteristic properties of the corresponding block matrices using the character tables of the symmetry group and the concept of irreducible representations [47–50]. The symmetry classification of eigenstates is the simplest for the translation group and finds its expression in Bloch’s theorem. A translation operator TR applied to a Bloch function ψk (r) yields TR ψk (r) = ψk (r + R) = e−ik·R ψk (r)

(A.1)

i.e., it multiplies the Bloch function by a phase factor depending on the wave vector k and the translation R. Owing to the fact that translations

326

A Appendices

commute with each other (the translation group is Abelian), there are only one-dimensional representations, namely the phase factors. The spherical symmetry of the Coulomb potential leads to the angular momentum classification of the eigenstates of an atom. For the hydrogen problem, we have the (2l + 1)-fold states with angular momentum l. They transform under rotations with the corresponding (2l + 1) × (2l + 1) matrices Dl (α, β, γ), which form a (2l + 1)-dimensional irreducible representation of the full rotation group. The group elements depend continuously on the parameters α, β, and γ which define the group element by the three Euler angles. In contrast with the full rotation group (which is infinite and continuous) the point groups of crystal lattices are finite and discrete. For example, the symmetry group of a cube, Oh , is the same as that of the sc, bcc, and fcc lattices. It consists of 48 elements: the identity (E), three axes with fourfold rotations (C4 , C42 ), four axes with threefold rotations (C3 ), six axes with twofold rotations (C2 ), and all these operations combined with the inversion (J). In general, the elements of the point group do not commute (the group in non-Abelian). However, the point group falls into disjunct classes of conjugated elements, where group elements A and B are called conjugated to each other if the relation A = XBX −1 holds for all X of the group. For the cubic point group there are 10 classes: E, 3C42 , 6C4 , 8C3 , 6C2 , J, 3JC42 , 6JC4 , 8JC3 , 6JC2

(A.2)

where the numbers in front of the symbols for the symmetry operations give the number of group elements belonging to the class. Consider now the block matrices that transform the degenerate invariant subspaces. They are d dimensional irreducible representations of the symmetry group. Diﬀerent irreducible representations with the same dimension d, D(X), D′ (X), are equivalent if there is a d dimensional matrix M with M = 0 and D(X) = M D′ (X)M −1 for all elements X of the group. Note, that with respect to this operation with M , the coeﬃcients of the characteristic polynomial of D(X), especially the trace of D(X) or the character, do not change. Thus, inequivalent irreducible representations can be distinguished by looking at their characters. Similarly for the operation of conjugation: all matrices of an irreducible representation belonging to a class have the same character. This leads to the character table listing the characters of the inequivalent irreducible representations for the diﬀerent classes of conjugated elements. These irreducible representations play the same role in classifying the eigenstates of H with respect to the point group as the crystal momentum k does for the translation group and the angular momentum l for the rotation group. Their meaning is that of quantum numbers due to the underlying symmetry. Already knowing the classes, it remains now to specify the number of the inequivalent irreducible representations and their dimensions. According to the theorems of the theory of finite groups, the number of classes equals the number of irreducible representations, i.e., the character table has the same number of rows and columns. Moreover, the sum over the

A.1 Elements of Group Theory

327

squared dimensions of the irreducible representations (which is the sum of the characters for the class containing the identity E, because it is represented by d-dimensional unit matrices) must be equal to the order g of the group. As it turns out, the order of the cubic symmetry group can be decomposed only in one way into 10 squared integers: 48 = 2(12 + 12 + 22 + 32 + 32 ),

(A.3)

i.e., the group has four one-dimensional, two two-dimensional, and four threedimensional irreducible representations. As the inversion J is an element of the group, the eigenstates can be classified as having even or odd parity. This is considered in the notation of the irreducible representations by a ± or g, u (for gerade or ungerade). Another remarkable property of the character table is that the rows and columns understood as vectors are orthogonal to each other when properly weighted with the number of elements in a class. These are the famous orthogonality relations of the characters. Diﬀerent notations are in use for the irreducible representations of the cubic point group Oh (see the character table). The notation with the symbols A, E, and T for one-, two-, and three-dimensional representations, respectively, is applied to characterize localized (e.g., impurity) states of the given point symmetry, while the notation with the symbol Γ refers to the Bloch states at the center of the Brillouin zone with k = (0, 0, 0). (Note, that this wave vector does not change under the symmetry operations and therefore, the group of this wave vector is the point group of the crystal.) Of the two different notations with the symbol Γ the one with double indices is the older one and indicates the removal of the level degeneracy for finite k. These symbols are found e.g., in some figures of Chaps. 5 and 9. Character table of the point group Oh E 3C42 8C3 6C4 6C2 J 3JC42 8JC3 6JC4 6JC2 A1g A2g Eg T1g T2g A1u A2u Eu T1u T2u

Γ+ 1 Γ+ 2 Γ+ 3 Γ+ 4 Γ+ 5 Γ− 1 Γ− 2 Γ− 3 Γ− 4 Γ− 5

Γ+ 1 Γ+ 2 Γ+ 12 Γ+ 15 Γ+ 25 Γ− 1 Γ− 2 Γ− 12 Γ− 15 Γ− 25

1 1 2 3 3 1 1 2 3 3

1 1 2 −1 −1 1 1 2 −1 −1

1 1 −1 0 0 1 1 −1 0 0

1 −1 0 1 −1 1 −1 0 1 −1

1 −1 0 −1 1 1 −1 0 −1 1

1 1 2 3 3 −1 −1 −2 −3 −3

1 1 2 −1 −1 −1 −1 −2 1 1

1 1 −1 0 0 −1 −1 1 0 0

1 −1 0 1 −1 −1 1 0 −1 1

1 −1 0 −1 1 −1 1 0 1 −1

It is quite instructive to specify objects (wave functions, operators) that transform according to these irreducible representations. Using the fact, that

328

A Appendices

the crystal point groups are subgroups of the full rotation group, this can be done by formulating the spherical harmonics in Cartesian coordinates to find the so-called cubic harmonics, which for l = 0, 1, 2, 3 are given by l = 0 → 1 (Γ+ 1) l = 1 → x, y, z (Γ− 4) 1 + l = 2 → z 2 − (x2 + y 2 ), x2 − y 2 (Γ+ 3 ) ; yz, zx, xy (Γ5 ) 2 − 2 2 2 2 2 2 l = 3 → xyz (Γ− 2 ) ; z(x − y ), x(y − z ), y(z − x ) (Γ5 ). Thus, the fivefold degeneracy of the l = 2 spherical harmonics splits under the reduced symmetry of the cubic point group into a twofold (Γ+ 3 ) and a threefold level (Γ+ ), which is the crystal field splitting discussed in Sect. 5.5 5 and Problem 5.8. Of the seven spherical harmonics with l = 3 only four appear in this list, while the remaining three are cubic harmonics which transform as Γ− 4 , i.e., the cubic crystal field causes a mixing of angular momentum states with l = 1 and 3. The point group Td of the zinc blende lattice is a subgroup of Oh and contains the classes E, 3C42 , 8C3 , 6JC4 , and 6JC2 . Consequently, the number of irreducible representations is reduced to five. Considering only the corresponding columns in the character table of Oh , we find identical rows for pairs of representations which merge into one irreducible representation of Td . Thus, the character table of Td is obtained from that of Oh . Note that Td does not contain the inversion J and eigenstates cannot be classified by parity. Character table of the point group Td E 3C42 8C3 6JC4 6JC2 A1 A2 E T1 T2

Γ1 Γ2 Γ3 Γ4 Γ5

Γ1 Γ2 Γ12 Γ15 Γ25

1 1 2 3 3

1 1 2 −1 −1

1 1 −1 0 0

1 −1 0 1 −1

1 −1 0 −1 1

The crystal momentum k is a good quantum number. Thus, for finite k the eigenstates have to be classified by the irreducible representations of the group of the wave vector, consisting of all elements of the crystal point group which do not change k. For k (1, 0, 0) or along the ∆-axis of the Brillouin zone the cubic point group Oh reduces to C4v with eight elements in five classes, and Td to C2v with four elements each in one class. Decomposing 8 = 1+1+1+1+22 gives the dimensions of the five irreducible representations of C4v , while (obviously) C2v has four one-dimensional representations. The notation of the irreducible representations reminds of the ∆-axis. The group of the wave vector along (1,0,0) keeps x unchanged while y and z change. Thus, x transforms as ∆1 , while for C4v y and z transform into each other like ∆5 and

A.2 Fourier Series and Fourier Transforms

329

the threefold states of symmetry Γ− 4 split into ∆1 + ∆5 (which is considered the notation Γ− 15 ). This explains the splitting of the phonon dispersion curves (Chap. 3) and of the energy bands (Chap. 5) away from the Γ point. For C2v , y and z transform according to diﬀerent irreducible representations ∆3 , ∆4 but due to time invariance these states are degenerate. Similar considerations hold for the other directions in k space.

C4v

E C42 2C4 2JC42 2JC2

∆1 ∆2 ∆3 ∆4 ∆5

1 1 1 1 2

1 1 1 1 −2

1 1 −1 −1 0

1 −1 1 −1 0

1 −1 −1 1 0

C2v

E C42 JC2 JC2′

∆1 ∆2 ∆3 ∆4

1 1 1 1

1 −1 1 −1

1 1 −1 −1

1 −1 −1 1

A.2 Fourier Series and Fourier Transforms Consider a function f (x) defined in the interval −L/2 ≤ x ≤ +L/2 or a periodic function f (x + L) = f (x). It can be represented by the Fourier series f (x) =

+∞

Fkn eikn x ,

with kn =

n=−∞

2nπ , L

n

integer.

(A.4)

The Fourier coeﬃcients are given by Fkn

1 = L

+L/2

f (x)e−ikn x dx.

(A.5)

−L/2

If the length L is taken as the linear extension of a solid and f (x) as a wave function describing some state of the solid, the periodicity of f (x) reflects the reasonable assumption that the physical properties connected with this state repeat with the period L or, what is equivalent, that they do not depend on L. This is the concept of periodic boundary conditions. The Fourier series expansion makes use of the fact that complex exponentials are normalized and orthogonal, i.e., 1 L

+L/2

e−ikm x eikn x dx = δmn ,

(A.6)

−L/2

and that they form a complete set on the interval of length L: 1 ikn (x−x′ ) e = δ(x − x′ ). L n

(A.7)

330

A Appendices

In the limit L → ∞ the discrete kn become the k space variable and (A.5) the Fourier transform +∞ 1 F (k) = f (x)e−ikx dx. (A.8) 2π −∞ In generalizing to the three-dimensional case we may write f (r) = Fk eik·r ,

(A.9)

k

where the components of the wave vector k can take the values ki = 2πni /Li , i = 1, 2, 3 with integer ni , i.e., the sum over k is to be understood as the triple sum over all integer values of ni . The orthogonality and normalization of the complex exponentials is expressed as ′ 1 (A.10) ei(k−k )·r d3 r = δk,k′ V V and the completeness as 1 ik(r−r′ ) e = δ(r − r ′ ), V

(A.11)

k

where V = L1 L2 L3 is the periodicity or crystal volume. The Fourier coeﬃcients take the form 1 Fk = f (r)e−ik·r d3 r. (A.12) V V In this book, discreteness of the k resulting from the finite volume V is frequently used. Nevertheless, we shall denote Fk as the Fourier transform of f (r). On the other hand, the discrete sum over k can be evaluated as an integral with the replacement V d3 k, (A.13) ⇒ (2π)3 k

where (2π)3 /V is the volume for each discrete k. Let us take f (r) = 1/|r| as an example by showing that 1 1 4π ik·r = e . r V k2

(A.14)

k

This is done in the following steps. First, we carry out the summation over k on the rhs as an integral in spherical polar coordinates ∞ π 1 4π ik·r V 4π 2π dkeikr cos θ . (A.15) dθ sin θ e = dφ V k2 (2π)3 V 0 0 0 k

A.3 Fermi and Bose Integrals

331

The integration over φ gives a factor 2π and the integration over θ can be performed with the substitution cos θ = z +1 1 ∞ 1 ∞ 1 ikr ikrz e − e−ikr . ··· = (A.16) dze = dk dk π 0 π 0 ikr −1

With the substitution x = kr and by writing the complex exponentials as the sin-function, we find 2 1 ∞ sin x ··· = . (A.17) dx πr 0 x The last integral gives π/2 and we arrive at the lhs of (A.14).

A.3 Fermi and Bose Integrals Frequently, physical quantities of fermion systems are expressed in terms of integrals of the form ∞ Iα (μ, T ) = E α f (E, μ, T )dE, (A.18) 0

with the Fermi–Dirac distribution function f (E, μ, T ) =

1 e(E−µ)/kB T

+1

.

(A.19)

Examples are the particle density n and the ground state energy E0 (T ) of free electrons in Sect. 4.1 with α = 1/2 and α = 3/2, respectively. By substituting x = E/kB T and η = μ/kB T the integral Iα (μ, T ) can be written Iα (μ, T ) = (kB T )α+1 Γ(α + 1)Fα (η),

(A.20)

with the Gamma function or Euler integral ∞ Γ(α) = tα−1 e−t dT , Reα > 0

(A.21)

0

and the Fermi integral of index α [244, 245, 292] ∞ 1 xα dx. Fα (η) = x−η Γ(α + 1) 0 e +1

(A.22)

In a three-dimensional system, one has for the particle density (with √ Γ(3/2) = π/2) n=

0

∞

D(E)f (E, μ, T )dE = 2

mkB T 2π¯ h2

!3/2

F1/2

μ kB T

!

,

(A.23)

332

A Appendices

because the density of states√ D(E) ∼ E 1/2 , while the ground state energy can be written (with Γ(5/2) = 3 π/4) ∞ ED(E)f (E, μ, T )dE E0 (T ) = 0

m 2π¯ h2

=3

!3/2

5/2

(kB T )

F3/2

μ kB T

!

.

Sometimes also integrals of the form ∞ d f (E, μ, T ). φ(E) J(μ, T ) = dE 0

(A.24)

(A.25)

appear [245], e.g., in the transport relaxation times for diﬀerent scattering processes in Sect. 8.3. For the derivative of the Fermi–Dirac distribution function we may write with x = (E − μ)/kB T , −1 df 1 df 1 = =− (ex + 1)−1 e−x + 1 , dE kB T dx kB T

(A.26)

which has a pronounced maximum at x = 0. Expand the function φ(E) (which is assumed to be smooth) around x = 0 or E = μ φ(E) = φ(μ + xkB T ) =

∞ (kB T )n (n) φ (μ)xn , n n=0

(A.27)

with φ(n) (μ) being the nth derivative, to obtain (with dE = kB T dx) ∞ ∞ df (kB T )n (n) φ (μ) (A.28) xn dx. J(μ, T ) = n dx −µ/kB T n=0 For μ ≫ kB T , the lower limit of the integral tends to −∞ and one defines ∞ df (A.29) xn−1 dx. Jn = dx −∞ Because df /dx is an even function, all integrals Jn with even n vanish. For odd n we obtain ∞ J1 = f ′ (x)dx = −1 (A.30) −∞

J3 =

∞

−∞

= −4

2 ′

x f (x)dx = −2

0

∞ (−1)k π2 =− 2 (k + 1) 3 k=0

∞

x2 e−x

2 dx

(1 + e−x )

(A.31)

A.4 Sommerfeld Expansion

333

(for the last see [293]). Thus, one obtains J(μ, T ) = −φ(μ) −

π2 (kB T )2 φ′′ (μ) + . . . . 6

For μ = 0 the Fermi integral is given in closed form by [293] ∞ ν−1 x dx = 1 − 21−ν Γ(ν)ζ(ν), Reν > 0 x+1 e 0

and the corresponding Bose integral by ∞ ν−1 x dx = Γ(ν)ζ(ν), x−1 e 0

Reν > 1

(A.32)

(A.33)

(A.34)

with the Γ-function and Riemann’s zeta-function ζ(z) =

∞ 1 , nz n=1

Rez > 1.

(A.35)

A.4 Sommerfeld Expansion In connection with electronic properties, integrals of the form +∞ g(E)f (E, μ, T )dE I(T ) =

(A.36)

−∞

have to be evaluated frequently, where f (E, μ, T ) is the Fermi–Dirac distribution function. Special cases are: the number of electrons in occupied states (with g(E) = D(E), the density of states) and the total energy of these electrons (with g(E) = ED(E)). In these cases, the evaluation of the integral is possible if g(E) fulfills the following conditions: (1) it should not be singular for E ≃ EF , (2) it should vanish for E → −∞, and (3) it should diverge for E → +∞ not stronger than some power in E. These conditions allow to write the integral (A.36) in the form of a series µ ∞ g(E)dE + αn (kB T )2n g (2n−1) (μ), (A.37) I(T ) = −∞

n=1

where g (2n−1) (μ) is the (2n − 1)th derivative of g(E) taken at the chemical potential μ. The coeﬃcients αn are determined by Riemann’s zeta function αn = 2 1 −

1 22n−1

!

ζ(2n),

ζ(n) =

∞ 1 . n p p=1

(A.38)

This expansion, originally derived in [109], is known as the Sommerfeld expansion.

334

A Appendices

The validity of (A.36) can be demonstrated as follows: Define a function p(E) whose derivative is g(E) or p(E) =

E

g(x)dx

←→

−∞

g(E) =

dp(E) dE

and write the integral (A.36) as +∞ dp(E) f (E, μ, T )dE I(T ) = dE −∞ +∞ df (E, μ, T ) dE. = p(E)f (E, μ, T ) |+∞ p(E) − −∞ dE −∞

(A.39)

(A.40)

Making use of the Taylor expansion of p(E) about μ ∞ (E − μ)n (n) p(E) = p(μ) + p (μ) n! n=1

(A.41)

and of the fact that the derivative of the Fermi–Dirac distribution function (with β = 1/kB T ) df (E, μ, T ) βeβ(E−µ) −β = − 2 = 2 dE 4 cosh (β(E − μ)/2) 1 + eβ(E−µ)

(A.42)

is an even function in E − μ, this can be written I(T ) =

µ

g(E)dE + β

−∞

∞ 1 (2n−1) g (μ)I2n (T ), 2n! n=1

(A.43)

where I2n (T ) =

+∞

−∞

(E − μ)2n

= β −(2n−1)

+∞

= = =

2 dE,

(exp (β(E − μ)) + 1)

x2n

−∞

=

exp (β(E − μ))

(ex

ex dx + 1)2

2 x2n−1 dx , −2β exp (λx) + 1 0 λ=1 1 ∞ 2n−1 2 u d −2n λ du −2β −(2n−1) dλ eu + 1 0 λ=1 ∞ 2n−1 u du 4nβ −(2n−1) eu + 1 0 ! 1 2 1 − 2n−1 (2n)!ζ(2n)β −(2n−1) . 2 −(2n−1)

1

d dλ

x = β(E − μ)

∞

Thus, we arrive at the expression (A.37).

λx = u

(A.44)

A.5 Calculation of the Exchange Energy

335

Usually g(E) is a power function and each derivative means a division by μ. As a consequence, subsequent terms in the series decrease by a factor (kB T /μ)2 and for kB T ≪ μ fast convergence is achieved. The coeﬃcients of the leading terms take the values π2 6

α1 = ζ(2) =

and α2 =

7π 4 . 360

(A.45)

A.5 Calculation of the Exchange Energy The sums over q and p can be performed as integrals over the Fermi sphere after replacing |p − q| with the cosine relation Eexch = −

= −

q,p,q=p |q|,|p|≤kF

e2 ε0 V |q − p|2

e2 Vc2 ε0 V (2π)6

d3 qd3 p

|p|,|q|≤kF

p2

+

q2

1 , (A.46) − 2pq cos ϑ

where ϑ is the angle between q and p. The square root of the integrand is known as the generating function of Legendre’s polynomials, i.e., ∞ 1 xL PL (cos ϑ), = (1 + x2 − 2x cos ϑ)1/2 L=0

x < 1.

(A.47)

In the double integral we have to distinguish p < q (x = p/q) and p > q (x = q/p), which gives two identical contributions, and we can write Eexch

1 e2 3 2πq 2 dq 2 d p 2V = − 6 ε0 (2π) p q

q p

!L+L′ (A.48)

0

The integral over ϑ is performed by making use of the orthogonality of the Legendre polynomials to obtain 2 p q !2L+2 e2 V 3 d p dq. (A.49) Eexch = − πε0 (2π)4 |p|
The integral over q yields p/(2L + 3) and allows one to perform the integral over p in spherical polar coordinates: Eexch = −

k4 e2 V 2 4π F . 4 πε0 (2π) (2L + 1)(2L + 3) 4 L

(A.50)

336

A Appendices

Summing over L with the help of ∞

L=0

∞ 1 1 = (2L + 1)(2L + 3) 2

L=0

1 1 − 2L + 1 2L + 3

!

=

1 , 2

(A.51)

we eventually find

Eexch = −

q,p,q=p |q|,|p|≤kF

e2 3 = − N e2 kF . 2 ε0 V |q − p| 16π 2 ε0

(A.52)

A.6 Operators in Fock Representation Using field operators (see Problem 4.8) a one-particle operator can be written Aˆ1 = Ψ† (r)Aˆ1 (r)Ψ(r)d3 r. (A.53) The expansion

Ψ(r) =

cαi ψαi (r)

(A.54)

i

with fermion operators cαi and a complete set of single-particle wave functions ψαi (r) leads immediately to (4.76) Aˆ1 = c†αi cαj ψα∗ i (r)Aˆ1 (r)ψαj (r)d3 r . (A.55) i,j ˆ1 |ψα ψαi |A j

If ψαj (r) is an eigenfunction of Aˆ1 (r) with eigenvalue Aαj , this reduces to (A.56) Aαj c†αj cαj . Aˆ1 = j

The two-particle operator written in fermion field operators ˆ A2 = Ψ† (r1 )Ψ† (r 2 )Aˆ2 (r 1 , r2 )Ψ(r2 )Ψ(r1 )d3 r 1 d3 r 2

(A.57)

takes, with the expansion (A.54), the form Aˆ2 = c†αi c†αj cαk cαl i,j,k,l

×

ψα∗ i (r 1 )ψα∗ j (r 2 )Aˆ2 (r 1 , r2 )ψαk (r 2 )ψαl (r1 )d3 r 1 d3 r 2 . (1)

(2)

(1)

(A.58) (2)

By convention, the matrix element is written as ψαi ψαj |Aˆ2 |ψαl ψαk and one obtains (4.77) Aˆ2 = ψα(1) ψα(2) |Aˆ2 |ψα(1) ψα(2) c†αi c†αj cαk cαl . (A.59) i j l k i,j,k,l

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Solutions

Solutions for Chap. 1 1.1:

Set of lattice vectors Rn = di=1 ni ai , ni integer, ai linear independent (d dimension of the system)

Point lattice:

Reciprocal lattice:

Wigner–Seitz cell:

Gm = dj=1 mj bj , mj integer, bj linear independent, and ai · bj = 2πδij Contains all points which are closer to a given Rn Than to any other Rn′ = Rn

(First) Brillouin zone: Wigner–Seitz cell of the reciprocal lattice d = 2, square lattice: a1 = a(1, 0) , a2 = a(0, 1) → b1 = 2π/a(1, 0) , b2 = 2π/a(0, 1)

d = 3, simple cubic (sc), body centered cubic (bcc), face centered cubic (fcc) a2 = a(0, 1, 0) , a3 = a(0, 0, 1) sc: a1 = a(1, 0, 0) , → b1 = 2π/a(1, 0, 0) , b2 = 2π/a(0, 1, 0) , b3 = 2π/a(0, 0, 1) bcc: a1 = a/2(1, 1, −1) , a2 = a/2(1, −1, 1) , a3 = a/2(−1, 1, 1) → b1 = 2π/a(1, 1, 0) , b2 = 2π/a(1, 0, 1) , b3 = 2π/a(0, 1, 1) fcc: a1 = a/2(0, 1, 1) , a2 = a/2(1, 0, 1) , a3 = a/2(1, 1, 0) → b1 = 2π/a(−1, 1, 1) , b2 = 2π/a(1, −1, 1) , b3 = 2π/a(1, 1, −1)

1.2: Create Fibonacci sequence by replacing LS → L and S → L [21]: LS LSL LSLLS LSLLSLSL LSLLSLSLLSLLS LSLLSLSLLSLLSLSLLSLSL

not periodic

350

Solutions

Replacing LS → L′ and L → S ′ in the last line gives the configuration of the second but last line (self-similarity or fractality). For the Fourier transform see [35]. 1.3: Given two vectors a1 , a2 , with |a1 | = a1 , |a2 | = a2 , and a1 · a2 = a1 a2 cos α, spanning a plane. The following five cases can be distinguished: a1 = a2 , α = π/2 α = π/3

square triangular or hexagonal

α = π/2, π/3

a1 = a2 , α = π/2

rectangular

α = π/2

1.4: TR n is the translation operator. It acts on a function according to TR n φ(r) = φ(r + Rn ) and commutes with the system Hamiltonian, [TR n , H] = 0. Therefore, there exist simultaneous eigenfunctions of H and TR n with the property TR n φk (r) = eik·R n φk (r) i.e. the wave eigenfunctions in diﬀerent Wigner–Seitz cells diﬀer only by a phase factor with wave vector k from the first Brillouin zone. 1.5: Count nearest neighbors (n.n.) and spheres per cube: = 4π a 3 π sc 6 n.n., 1 sphere → a3 = = 0.52, 3 2 6 √ !3 > √ 4π 3a 3π 3 bcc 8 n.n., 2 spheres → 2 = 0.68, a = 3 4 8 !3 > 4π a π √ fcc 12 n.n., 4 spheres → 4 a3 = √ = 0.74, 3 2 2 3 2 √ √ !3 > 3a 3π 4π a3 = = 0.34. diamond 4 n.n., 8 spheres → 8 3 8 16 1.6: A mass density n(r) = δ(r − r i ) and using δ(r − r i ) = q exp (iq · (r − r i ))/V gives for the scattering amplitude iq·r 1 i(k−k ′ +q)·r iq·r e i e = e i δq,k−k ′ F (k, k′ ) = F (k − k′ ) = V V i,q i,q and with r i → R n + τ for a crystalline solid iq·τ iq·R n e , e F (q) = τ

n

where the last sum vanishes except for q = G and n eiq·R n = N δq ,G . Thus the scattering amplitude, which equals the static structure factor (up to a factor N ) has peaks for the reciprocal lattice vectors.

Solutions

351

Consider the reciprocal lattice vector Ghkl = hb1 + kb2 + lb3 and the lattice plane with Miller indices (hkl) spanned by a1 /h′ − a2 /k′ and a3 /l′ − a2 /k′ with (hkl) = p(h′ k′ l′ ), p = integer. The normal to the lattice plane is given by a a2 a3 a2 1 1 1 1 − ′ × − ′ = − ′ ′ a1 × a 2 − ′ ′ a 2 × a 3 − ′ ′ a3 × a 1 . ′ ′ h k l k hk kl hl

Multiplication of this vector with −2πh′ k′ l′ /a1 · (a2 × a3 ) gives Ghkl /p, thus Ghkl is normal to the lattice planes (hkl). The distance of the considered plane from the origin of the vectors a1 , a2 , a3 is d′hkl =

a1 a1 · Ghkl 2π h 2π = = p. h′ a1 Ghkl Ghkl h′ Ghkl

Thus, 2π/Ghkl is the distance between neighboring lattice planes. Solutions for Chap. 2 ? ρˆ] = 0 reads 2.1: The matrix representation of the commutator [H, ? ρˆ − ρˆH|n

? m|H = Hmn′ ρn′ n − ρmn′ Hn′ n = 0. n′

? one has Hmn′ = Em δm,n′ and For eigenstates of H Em − En ρmn = 0,

which means for Em = En or m = n for nondegenerate states that ρmn = 0. The same result is found by starting from the statistical operator of the canonical (or grand canonical) ensemble. ˆ 0 . With ˆ = Ψ0 |A|Ψ ρA 2.2: Denote the ground state by |Ψ0 and show that Trˆ T =0 ? ρˆ = exp (−β H)/Z write −βE ? m ˆ = 1 ˆ m = 1 ˆ m . Tr ρˆA Ψm |e−β H |Ψn Ψn |A|Ψ e Ψm |A|Ψ Z m,n Z m −β(E −E ) m 0 e we can write With Z = e−βE0 ˆ = Tr ρˆA

m

−β(Em −E0 ) ˆ m

Ψm |A|Ψ m e −β(E −E ) m 0 me

−→ T →0

ˆ 0

Ψ0 |A|Ψ

because with Em − E0 > 0 for all m = 0 all exponential factors vanish for T → 0 except the one for the ground state |Ψ0 . ? = n 2.3: The thermal expectation value of the number operator is N α ˆ α where ? The grand canonical n ˆ α is the operator counting the particles in an eigenstate of H. partition function can be written ZG =

∞

N=0 {nα }N

eβ

α (µ−Eα )nα

=

* α

eβ(µ−Eα )nα ,

nα

where {nα }N denotes those sets of particle numbers nα whose sum is N . For fermions: nα can be 0 or 1 and

352

Solutions ZG =

* α

1 + eβ(µ−Eα ) ,

while for bosons nα can be any nonnegative integer and * −1 * . 1 − eβ(µ−Eα ) 1 + eβ(µ−Eα ) + . . . = ZG = α

α

Now make use of

1 1 ∂ZG 1 ∂ ln ZG = β ∂μ β ZG ∂μ 1 ? = to obtain for fermions N ˆ nα = exp (−β(μ − Eα )) + 1 α α 1 ? = . and for bosons N ˆ nα = exp (−β(μ − Eα )) − 1 α α ? = N

The function

fα± =

1 exp (−β(μ − Eα )) ± 1

is the Fermi–Dirac distribution (upper sign) and the Bose–Einstein distribution (lower sign). For large kB T ≫ μ we have exp βμ ≃ 1 and fα± → exp (−βEα ).

2.4: Ohm’s law can be written jλ = σλµ Eµ , thus the observable to be measured is (a component of) the electrical current density Aˆ → ˆ jλ = e

N

vˆl,λ

l=1

with the velocity operator vˆl,λ = pˆl,λ /m. It can be written also as e ˆ (ˆ pl,λ δ(r − r l ) + δ(r − r l )ˆ pl,λ ) . jλ (r) = 2m l

The operator of the kinetic energy in the presence of an electro-magnetic field (here represented by the vector potential A(r, t)) is for the lth electron 2 e 1 p2 ˆ l − eA(r l , t) = l − p pl · A(r l , t) + A(r l , t) · pl + O(A2 ). 2m 2m 2m

Neglecting the last term on the rhs, we identify the perturbation as e or Vext (t) = − pl · A(r l , t) + A(r l , t) · pl 2m l e pl δ(r − r l ) + δ(r − r l )pl ·A(r, t) = − d3 r 2m l ˆ j (r) l

? as another component of the electric current density. The and the observable B electric field component is given by Eµ = −

∂Aµ = −iωAµ ∂t

and we can write

Solutions Vext (t) = −

353

i jµ (r)Eµ (r)eiωt d3 r ˆ ω µ

which in the long-wave length limit, when the dependence of the vector potential on r can be neglected, gives i d3 rjµ (r)Eµ eiωt Vext (t) = − ω µ and we obtain the electric conductivity ∞ i σλµ (ω) = dτ eiωt θ(τ )[jλ (τ ), jµ (0)] 0 hω −∞ ¯ as a correlation function for the components of the current density. 2.5: With ∆ρ(t) = ∆ρ1 (t) + ∆ρ2 (t) + . . . , where the index refers to diﬀerent orders of Vext , we can write the equations: 1.

[H0 , ∆ρ1 ] + [Vext , ρ0 ] = i¯ hρ˙ 1

first order in Vext .

2.

[H0 , ∆ρ2 ] + [Vext , ρ1 ] = i¯ hρ˙ 2

second order Vext .

The solution of (1) ∆ρ1 (t) =

1 i¯ h

t

′

′

dt′ e−iH0 (t−t )/¯h [Vext (t′ ), ρ0 ]eiH0 (t−t )/¯h

−∞

is to be used in (2), which can be solved in the same way as (1), to yield ∆ρ2 (t) =

1 i¯ h

t −∞

dt′

t′

′

dt′′ e−iH0 (t−t )/¯h eiH0 (t−t

′′

)/¯ h

−∞

where the double commutator / 0 ′′ [. . . , [. . . , . . . ]] = Vext (t′ ), e−iH0 (t−t )/¯h [Vext (t′′ ), ρ0 ]

indicates the structure of the second-order response function 0 / ? ), [B(τ ? ′ ), A(0)] ˆ ∆A2 t = . . . B(τ . It is a two-time correlation function.

2.6: Evaluate the principal value integral +∞ χ0 1 Reχ(ω) = P δ(ω0 − ω ′ ) − δ(ω0 + ω ′ ) dω ′ , ′−ω π ω −∞

which yields

Reχ(ω) =

χ0 π

1 1 + ω0 − ω ω0 + ω

!

=

′

[. . . , [. . . , . . . ]]eiH0 (t−t )/¯h

1 2χ0 ω0 . π ω02 − ω 2

354

Solutions

Solutions for Chap. 3 3.1: The potential energy of the linear chain (Fig. 3.16) is ⎧ ⎪ ⎪ ⎪ ! ! !22 1 1 f ⎨ n′ n′ n′ + 1 −u −u u + u U= 1 1 2 2⎪ ⎪ ⎪ n′ ⎩ (1)

(2)

⎫ ⎪ ⎪ !22 ⎪ ⎬ n′ , 2 ⎪ ⎪ ⎪ ⎭

here (1) is the contribution due to the relative displacements of M1 and M2 in the unit cell n and (2) that due to the relative displacement between M1 in unit cell n + 1 and M2 in unit cell n. Distinguish the force constants ! ! nn nn = 2f =φ n = m, τ = τ ′ : φ 2 2 1 1 for the restoring force acting on M1 (M2 ) if the neighbor atoms are kept fixed, ! nn ′ n = m, τ = 1, τ = 2 : φ = −f and 1 2 ! m+1 m ′ n = m + 1, τ = 1, τ = 2 : φ = −f. 1 2 Due to actio = reactio we have ! ! nn nn φ =φ and with (3.13) 1 2 2 1 ! ! ! n n−1 nn nn = 2f − f − f = 0 +φ +φ for τ = 1 : φ 1 2 1 2 1 1 ! ! ! n n+1 nn nn = 2f − f − f = 0. +φ +φ for τ = 2 : φ 2 1 2 1 2 2

Translation invariance allows to shift the cell index. Considering the equilibrium positions xn1 = na for M1 and nn2 = na + r0 for M2 along the chain, we find the elements of the dynamical matrix ! 2f 2f 1 nn D11 (q) = φ = , D22 (q) = 1 1 M1 M1 M2 and D12 (q) = √

1 φ M1 M2 n

nm 1 2

!

eiq(xn1 −xn2 ) .

For n = m, xn1 − xn2 = −r0 and for n = m + 1, xm+1,1 − xm,2 = a − r0 , the force constant equals −f and with a = 2r0 , we have D12 (q) = √

−2f cos (qr0 ) = D21 (q). M1 M2

Solutions

355

The eigensolutions of $ $ $ Dαβ (q) − ω 2 (q)δα,β = $ $

2f −2f √ cos qr0 − ω2 M1 M1 M2 2f −2f √ − ω2 cos qr 0 M2 M1 M2

are (with M = M1 + M2 ) 2 (q) = ω±

$ $ $ $=0 $

/ )1/2 0 ( f . M ± M12 + M22 + 2M1 M2 (1 − 2 sin2 qr0 M1 M2

There are two solutions for each q. For q = 0, the squared frequencies are 2 ω+ (0) =

2f M , M1 M2

and

2 ω− (0) = 0

and for q ≃ 0 with 1 − 2 sin2 qr0 = cos qa ≃ 1 − q 2 a2 /2 (for q ≪ π/a) ω 2 (q) ≃

& / ( )1/2 0 μ 2 2 '0 f / f 1± 1− M ± M 2 − M1 M2 q 2 a2 ≃ q a M1 M2 μ 2M

with the reduced mass μ = M1 M2 /M and 2 (q ≪ ω+

2f π )≃ a μ

independent of q

and

% f 2 2 f π π q a → ω− (q ≪ ) ≃ aq. ≪ )≃ a 2M a 2M For q = π/a and cos qa = −1 the solutions are π )1/2 ( 2 ω± M ± M12 + M22 − 2M1 M2 or a π % 2f π % 2f , ω− . ω+ = = a M2 a M1 A plot of the two branches is shown in Fig. 10.1: the lower branch with the linear dependence around q ≃ 0 for the acoustic phonons and the flat upper branch for the optical phonons are separated by a gap which results from the diﬀerent masses M1 = M2 . For M1 = M2 = M the gap closes and we have a chain with period a and the dispersion % 4f ω(q) = sin qa M extends to π/a (thin dashed line in Fig. 10.1), which is the limit of the Brillouin zone for the chain with lattice constant a. Solving the eigenvector equations, we √find √ for q = 0 : ω− (0) = 0 e− (0) ∼ move with same phase M1 , M2 8 √ √ 2f move with opposite phase e+ (0) ∼ M2 , − M1 ω+ (0) = 8µ 2f for q = πa : ω− ( πa ) = M1 e− ( πa ) ∼ (1, 0) M2 in rest 8 2f e+ ( πa ) ∼ (0, 1) M1 in rest. ω+ ( πa ) = M 2 2 ω− (q)(q

356

Solutions

ω(q)

0

π/2a q

π/a

Fig. A.1. Dispersion for the linear chain with two diﬀerent masses per unit cell (solid lines). The dashed curves show the result if the two masses are equal

y

7

2

6

3

0

1

8

4

5

x

Fig. A.2. Sketch of the two-dimensional quadratic lattice with numbers to address the individual lattice point 3.2: For central forces the adiabatic potential depends on rm,n = |R 0m +um −R0n − un | and the force constants can be written ! 0 0 ∂ 2 U Rki Rkj mn φ = with k = m − n. 2 0 i j ∂rm,n |Rk |2 Nearest neighbors to mass in the center (see Fig. 10.2) are R01 = aex , R02 = aey , R03 = −aex , R 04 = −aey

Solutions with force constants ! 10 =φ φ xx

20 yy

!

3 0 xx

=φ

!

40 yy

=φ

!

k0 xy

= φ1 , φ

!

357

= 00

next nearest neighbors are R 05 = aex − ey , R07 = −aex + aey with force constants φ φ and

! ! ! ! 1 5 0 50 70 70 =φ =φ =φ = φ2 , xx yy xx yy 2 ! ! ! ! 1 70 70 50 50 = − φ2 =φ =φ =φ yx xy yx xy 2

with

φ

80 xx

!

=

1 φ2 . 2

=φ

80 xy

φ

k0 i j

=φ

R 08 = −aex − aey

and

!

!

φ

1 φ2 2

80 yy

=φ

with

=

=φ

60 xy

60 yy

= aex + aey

!

!

!

60 xx

R06

The force constant φ

→

φ

00 i j

!

00 i j

!

follows from

k

=−

8

φ

k=1

k0 i j

!

!

=0

= −2(φ1 + φ2 )δi,j .

The elements of the dynamical matrix are ! 0 1 k 0 φ e−iq·R k Dxx (q) = xx M k

2 =− (φ1 (1 − cos qx a) + φ2 (1 − cos qx a cos qy a)) , M 2 (φ1 (1 − cos qy a) + φ2 (1 − cos qx a cos qy a)) , Dyy (q) = − M 2 Dxy (q) = − φ2 sin qx a sin qy a. M

The secular problem is $ $ $ Dxx (q) − ω 2 $ Dxy (q) $ $ = 0. $ Dxy (q) Dyy (q) − M ω 2 $

For Γ − X : 0 ≤ qx ≤ π/a , qy = 0 , sin qy a = 0 , cos qy a = 1 we have two branches ω12 (qx ) =

2 (φ1 + φ2 )(1 − cos qx a) M

and

with ωqx ∼ qx for qx ≪ π/a but diﬀerent slopes.

ω22 (qx ) =

2 φ2 (1 − cos qx a) M

358

Solutions

√ √ For Γ − M : q = (a, q)/ 2 , 0 ≤ q ≤ 2π/a the secular problem yields ⎛

⎜ qa ⎜ 2 ⎜M ω − 2φ1 1 − cos √ ⎝ 2

!

with the solutions

− 2φ2 2a

⎞2 !⎟ qa ⎟ qa 1 − cos2 √ ⎟ − 4 φ22 sin4 √ = 0 2 ⎠ 2 b2

! qa , M ω 2 = 2a ± 2b : M ω12 = 2φ1 1 − cos √ 2 ! ! qa 2qa M ω22 = 2φ1 1 − cos √ + 2φ2 1 − cos √ . 2 2 Again we find two branches with ω(q) ∼ q for q ≪ π/a and diﬀerent slopes. For X − M : q = (π/a, q) , 0 ≤ q ≤ π/a the oﬀ-diagonal terms of the dynamical matrix vanish and one has ω12 =

2 2φ1 + φ2 (1 + cos qa) M

and

ω22 =

2 φ1 + φ2 − (φ1 − φ2 ) cos qa . M

The two branches have always finite frequencies and connect those already obtained for the X and M point. For each of these directions one eigenvector is longitudinal ( q) and one transverse (⊥ q). 3.3: Periodic boundaries account for the fact that the physics of a solid repeats over macroscopic distances, i.e., the Bloch phase factor equals one for a translation RN = N1 a1 + N2 a2 + N3 a3 over a macroscopic length (Ni ≫ 1, i = 1, 2, 3): eik·R N = 1

or

k · RN = 2π × integer.

This implies (take a simple cubic lattice with V = L1 L2 L3 , Li = Ni a as example) that the components of the wave vector take the discrete values (particle in the box) ki = 2πni /Li , i = 1, 2, 3 and 0 ≤ ni ≤ Ni − 1. Thus N = N1 N2 N3 is the number of k in the first Brillouin zone and each k takes a volume (2π)3 /V . This can be exploited when replacing a sum over k by an integral according to V . . . d3 k. ··· = 3 (2π) k Similar considerations hold for systems with reduced dimension. 3.4: For the commutator [as (q), a†s′ (q ′ )] evaluate ˆ s′ (−q ′ ) − iPˆs′ (q ′ )] = ˆ s (q) + iPˆs (−q), ωs′ (q ′ )Q [ωs (q)Q ′ ′ ˆ s (q), Pˆs′ (q )] +iωs′ (q ) [Pˆs (−q), Q ˆ s′ (−q ′ )] = 2¯ hωs (q)δs,s′ δq,q ′ −iωs (q) [Q i¯ hδs,s′ δq,q ′

−i¯ hδs,s′ δq,q′

to find [as (q), a†s′ (q ′ )] = δs,s′ δq,q ′ . The two other commutation relations follow in the same way (note that ωs (q) = ωs (−q)).

Solutions

359

3.5: In the Schr¨ odinger picture, we have hωs′ (q ′ )as (q) [as (q), H0 ] = ¯

or

as (q)H0 = H0 + ¯ hωs′ (q ′ ) as (q).

Thus, for any power function f (H0 ) we can write

as (q)f (H0 ) = f (H0 + ¯ hωs (q))as (q) and obtain as (q, t)a†s′ (q ′ , 0) = eiH0 t/¯h as (q, 0)e−iH0 t/¯h a†s′ (q, 0) = e−iωs (q)t as (q, 0)a†s′ (q, 0). Using the commutation relation and taking the thermal expectation value gives as (q, t)a†s′ (q ′ , 0) = e−iωs (q)t ns (q, T ) + 1 δs,s′ δq,q ′ . Similarly we have

hωs′ (q ′ ) a†s (q) a†s (q)H0 = H0 − ¯

and by following the same steps we find the second relation. For the third relation we have, after extracting the exponential with the time-dependence, the thermal expectation value of a product of two annihilation operators, which vanishes. 3.6: The displacement is a time-dependent operator in the Heisenberg picture. Start by writing (with [ρ0 , H0 ] = 0) 2 2 q · un (t) = Tr ρ0 eiH0 t/¯h q · un (0) e−iH0 t/¯h

and obtain by cyclic permutation under the trace 2 2 , q · un (t) = Tr q · un (0)

which is independent of t. Formulate the lattice displacement with (3.23) and (3.39) un =

s,q

¯ h 2N M ωs (q)

!1/2

0 a†s (−q) + as (q) es (q)eiq·R n

in terms of phonon operators and evaluate

q · un

2

=

† 1 as′ (−q ′ ) + as′ (q ′ ) a†s′′ (−q ′′ ) + as′′ (q ′′ ) NM ′ ′ s ,q s′′ ,q ′′

×

¯2 h ′ 4ωs′ (q )ωs′′ (q ′′ )

!1/2

ei(q

′

+q ′′ )·R 0 n

q · es′ (q ′ ) q · es′′ (q ′′ ) .

After multiplying out the operator terms, the thermal expectation value of this expression follows with the formulas of problem 3.4 as 2 q · un =

2 1 h ¯ q · es′ (q ′ ) 2ns′ (q ′ , T ) + 1 ′ ′ NM ′ ′ 2ωs (q ) s ,q

360

Solutions

For T → 0 we have ns (q, T ) → 0 which leaves only the contribution of the zero-point motion 2 2 h ¯ 1 q · un = q · es′ (q ′ ) . N M ′ ′ 2ωs′ (q ′ ) s ,q

For T > 0, employ the Debye model by writing ωs (q) = vq for all s and independent 2 of the direction of q. Summation over s′ yields a factor 3. Writing q · es′ (q ′ ) = q 2 cos ϑ2 where ϑ is the angle between q and es′ (q ′ ), the sum over q ′ can be carried out in spherical polar coordinates with the cut-oﬀ at qD = ωD /v (with the Debye frequency ωD ) giving ! ωD 2 1 6q 2 1 dω . + q · un = h ¯ ω 3 M ωD exp (¯ hω/kB T ) − 1 2 0

For high temperatures, kB T ≫ ¯ hω, the distribution function after expanding the hω and by neglecting the term 1/2 we find for the Debye– exponential yields kBT /¯ Waller factor ωD 3q 2 3q 2 kB T W = k T dω = 3 B 2 M ωD M ωD 0

which is always positive. ¯ ω, the integral over ω reads after substituting For low temperatures, kB T ≪ h hω/kB T = x ¯ ωD 2 hωD ¯ (kB T )2 xD xdx . . . dω = + 4 h ¯ ex − 1 0 0

For T → 0 the upper limit goes to ∞ and the integral takes the value π 2 /6 (see Appendix A.3). Thus, we may write ! !2 2 1 T π 2 (kB T )2 π2 3¯ h2 q 2 3q 2 hωD ¯ + + = W = 3 M ωD 4 6 h ¯ M kB Θ D 4 6 ΘD with the Debye temperature ΘD .

3.7: (a) The point group of a cubic lattice consists of 48 elements (24 rotations, each can be combined with the inversion). Under these operations, which can be represented by orthogonal 3×3 matrices Sαi , the coordinates x, y, z are interchanged and (under inversion) change their sign. Likewise the components of the elastic tensor transform according to cαβγδ = Sαi Sβj Sγk Sδl cijkl

(double index summation) .

The invariance of the elastic tensor under these transformations leaves only those components diﬀerent from zero, for which pairs of indices are identical and of the nonvanishing components all those are identical which transform into each other by the symmetry operations. Thus, there are only three independent tensor components cxxxx = cyyyy = czzzz = c11 cxyxy = cxzxz = cyxyx = czxzx = c12 cxxyy = cyyzz = czzxx = cxxzz = czzyy = c44 which are written here in Voigt notation.

Solutions

361

(b) Using ui (r, t) = ui exp (i(q · r − ωt)) the wave equation for the elastic displacement field leads to a set of coupled homogeneous linear equations for the components ui , which has solutions if ρω 2 δil − cijkl qj qk = 0 . For Γ − X or q = (q, 0, 0) it reads $ 2 $ $ ρω − c11 q 2 $ 0 0 $ $ 2 2 $ $=0 0 ρω − c12 q 0 $ $ 2 2 $ 0 0 ρω − c12 q $

and has solutions

ωL = ωT =

%

%

c11 q , eL = (1, 0, 0) ρ

longitudinal

c12 q , eT = (0, 1, 0) ρ

transverse

= (0, 0, 1) transverse . √ For Γ − K or q = (q, q, 0)/ 2, we have cijkl qj qk = cixxl + ciyyl + cixyl + ciyxl q 2 /2 and the secular problem $ 2 1 $ $ ρω − 2 (c11 + c12 )q 2 − 21 (c12 + c44 )q 2 $ 0 $ $ $ $ $ − 1 (c12 + c44 )q 2 ρω 2 − 1 (c11 + c12 )q 2 $ = 0. 0 2 2 $ $ $ $ 2 2$ $ 0 0 ρω − c12 q One solution is immediately found to be % c12 ωT1 = q , eT1 = (0, 0, 1) ρ

transverse .

The other two follow from $ $ 2 $ ρω − A −B $ 2 − 2 $ $ $ −B ρω 2 − A $ = (ρω − A)2 B = 0

with A = (c11 + c12 )q 3 /2 and B = (c12 + c44 )q 2 /2 and read % √ 2c12 + c11 + c44 q , eT2 = (1, −1, 0)/ 2 transverse ωT2 = 2ρ % √ c11 − c44 q , eL = (1, 1, 0)/ 2 longitudinal . ωL = 2ρ

These results can be compared with the phonon dispersion (e.g., those given in Sect. 3.6 which are all for lattices with cubic symmetry). The slope of the acoustic branches for given ρ can be taken to determine the elastic constants. 3.8: The cubic anharmonicity ∆(a† + a)3 ) is first written in normal order (a† + a)3 = a†3 + 3a†2 a + 3(a† + a) + 3a† a2 + a3

362

Solutions

and then truncated by replacing a† a → a† a = n(T ) and omitting the terms a†3 and a3 . Thus the Hamiltonian reduces to H=¯ hω0 a† a + ∆(T )(a† + a) with ∆(T ) = 3∆ n(T ) + 1 .

Calculate now the corrections to the oscillator ground state |n with n = 0 due to the anharmonicity, which by making use of 0|a|1 1|a† |0 = 1 reads in Brillouin–Wigner perturbation theory hω0 ¯ ∆2 (T ) . ε = E0 − = 2 ε−¯ hω0 The smaller solution of the quadratic equation in ε E0 =

1 ∆2 (T ) hω0 − ¯ 2 hω0 ¯

expresses a zero-point energy which decreases as ∆(T ) increases due to thermal phonon excitation with the temperature. This is the behavior of a soft mode. Solutions for Chap. 4 4.1: The electrostatic potential of a homogeneous positive charge density +eN (r) with n(r) = N/V is e n(r ′ ) d3 r ′ . φ(r) = ′ V 4πε0 |r − r | Its interaction energy with the homogeneous electron density −en(r) is !2 e2 N 3 3 Hel−ion = −e d r d3 r ′ n(r)φ(r) d r = − . V 4πε0 |r − r ′ | V With (see Appendix) ′ e2 e2 vq eiq·(r−r ) , vq = = ′ 4πε0 |r − r | ε0 V q 2 q

d3 r

d3 r ′

e2 vq = 4πε0 |r − r ′ | q 2

the double integral can be evaluated

′ d3 reiq·r d3 reiq·r = vq V 2 δq,0 q

V δq,0

V δq,0

giving Hel−ion = −N v0 . Similarly the interaction energy of the homogeneous electron and ions systems can be calculated which each give the same result up to a factor −1/2. Thus the sum of all these divergent interaction energies vanish for the jellium model. 4.2: The electron density n determines via the density of states D(E) and the Fermi– Dirac distribution function the chemical potential μ(T ). For a 3D electron system √ 3/2 we may write D(E) = 3n E/2EF define a function G(E) with

E

D(E ′ )dE ′ = G(E) = n

−∞

to express the particle density as

E EF

!3/2

Solutions "∞ n = G(E)f (E, μ, T )"0 −

∞

dEG(E)

−∞

363

∂f (E, μ, T ) . ∂E

The first term vanishes and the integral can be evaluated by using the fact, that the derivative of f (E, μ, T ) is strongly peaked at E = μ (for kB T ≪ μ). Expand G(E) in a power series around E = μ µ ∞ " (E − μ)n dn G(E) " . G(E) = dE ′ D(E ′ ) + " n n! dE E=µ −∞ n=1

The integral gives G(μ). Because ∂f /∂E is an even function only the even powers of the expansion contribute and we find as the two leading terms ∞ ∞ 3n 1 ∂f ∂f − (∗). dE dE(E − μ)2 n = −G(μ) √ 3/2 dE μ dE −∞ 8EF −∞ The first term gives n(μ/EF )3/2 . With ∂f eβ(E−µ) , β = kB T and substituting x = β(E − μ) = −β β(E−µ) dE (e + 1)2 ∞ ∞ 1 ∂f ex dE(E − μ)2 . dx x2 =− 3 x dE β −∞ (e + 1)2 −∞

The value of the integral is π 2 /6. Thus (∗) reduces to the relation !2 2 !3/2 π μ kB T 1≃ 1+ EF μ 8 which can be solved to give μ(T ) ≃ EF

π2 1− 12

kB T μ

!2

.

This result corresponds to the Sommerfeld expansion (see Appendix A.4). 4.3: For 2D, the number of states (per unit area) D(k)dk in a circular ring with radius k and thickness dk is D(k)dk =

2 2π kdk (2π)2

(the factor 2 counts the spins) .

Use the dispersion relation for free electrons Ek = ¯ h2 k2 /2m to substitute k by E % % dk 2m 1 1 2m √ √ dE D(E)dE = D(k) E dE = dE π h2 ¯ h2 2 E ¯ m = const . π¯ h2 For 1D the corresponding number of states per unit length is % 2 dk 1 2m 1 √ dE D(k)dk = dk and D(k) dE = 2π dE π h2 2 E ¯ or

D(E) =

364

Solutions %

1 m √ . 2π 2 ¯ h2 E For a zero-dimensional system the spectrum is discrete (with energies Ei ) and the density of states is given by D(E) = 2 δ(E − Ei ) . or

D(E) =

i

4.4: The condition to fill n electrons into the lowest (spin-degenerate) Landau level follows from (4.43) or in simplified form from EF = ¯ hωc and reads ¯2 h heB ¯ (3π 2 n)2/3 = 2m m

which can be solved to give

B=

h ¯ (3π 2 n)2/3 . 2e

Note, that B is related to the number of elementary flux quanta. Take the value for ¯ h/e = 0.658 10−15 Tm2 to obtain for a metal n = 1023 cm−3 a magnetic field of B ≃ 1.37 105 T and for a doped semiconductor with n = 1014 cm−3 B ≃ 0.137 T. The latter is easily achieved in a laboratory. 4.5: The Zeeman energy for free electrons is ±μB B. It shifts the density of states of up and down spins (Landau quantization is not considered here) against each other 1 1 1 dD(E) D(E ± μB B) ≃ D(E) ± μB B which yields 2 2 2 dE ! !3/2 √ 2m μB B 1 1 . D± (E, B) ≃ E 1 ± 2 2π 2 ¯ E h2 The number of spin up and down electrons N± (E, B) is obtained by integrating the density of states multiplied with the Fermi–Dirac distribution function and multiplying with the volume V . The first term of D± (E, B) gives N (μ)/2 independent of B. The second term is evaluated by employing the Sommerfeld expansion (see Appendix A.4) leading to & µ dD(E) ' π2 d2 D(E) "" 1 dE + (kB T )2 . · · · = ± μB B 2 E=µ 2 dE 6 dE −∞ √ The integral gives D(μ) and with D(E) ∼ E the second term can be rewritten using d2 D(E)/dE 2 = −D(E)/4E 2 to obtain !2 ' & V π 2 kB T 1 N± (E, B) ≃ N (μ) ± μB BD(μ) 1 − . 2 2 24 μ D± (E, B) =

The magnetization follows as M = μB (N+ − N− )/V =

μ2B BD(μ)

&

π2 1− 24

kB T μ

!2 '

For T = 0 with D(μ ≃ EF ) = 3n/2EF this is identical with (4.53).

.

4.6: The HF approximation is better for the electron system with the smaller density parameter rs . According to Table 4.1 the rs -values of doped semiconductors are smaller than those of metals. On the other hand the electron density (per cm−3 ) is

Solutions

365

higher in metals. Note, that rs is given in the length scale (eﬀective Bohr radius) of the material. 4.7: Second-order perturbation yields a contribution E2 = −

|Ψm |Hint |Ψ0 |2 . Em − E0 m

(a) Applying the interaction operator to the Fermi sphere |Ψ0 gives nonvanishing contributions only if the states with p, q are inside and those with p − k, q + k outside of the Fermi sphere. Consider therefore the corresponding excited states |Ψm . These states with Em − E0 = ¯ h2 k · (q − p + k)/m contribute to E2 . b) The electron taken from p to p − k is put back to p in the direct process but to q in the exchange process. For the former evaluate vk c†p−kσ c†q+kσ ′ cqσ ′ cpσ |Ψ0

vk ′ c†p ′ −k ′ ρ c†q ′ +k ′ ρ′ cq ′ ρ′ cp ′ ρ |Ψm Ψm | Ψ0 | p′ q ′ k′ ρρ′

pqk σσ′

For the direct process the intermediate states have to fulfill the conditions p, σ = p′ − k ′ , ρ

q, σ ′ = q ′ + k ′ , ρ′ or

p − k, σ = p′ , ρ

q + k, σ ′ = q ′ , ρ′

σ = ρ , σ ′ = ρ′ , p′ = p − k , q ′ = q + k , k ′ = −k .

It remains to determine

Ψ0 |c†pσ c†qσ ′ cq+kσ ′ cp−kσ c†p−kσ c†q+kσ ′ cqσ ′ cpσ |Ψ0 = . . . , which can easily be rearranged as for k = 0 all fermion operators anti-commute and one obtains . . . = Ψ0 |c†pσ cpσ c†qσ ′ cqσ ′ cq+kσ ′ c†q+kσ ′ cp−kσ c†p−kσ |Ψ0

= npσ nqσ ′ (1 − nq+kσ ′ )(1 − np−kσ ),

where nqσ = θ(kF − q) is the Fermi–Dirac distribution function for T = 0 K. Summing over spin indices (factor 4) gives for the direct process 2 m npσ nqσ ′ (1 − nq+kσ ′ )(1 − np−kσ ) . E2dir = −4 vk 2 h ¯ k · (q − p + k) p,q,k c) For small k, i.e., excitation close to the Fermi surface np+k ≃ np + k · ∇p np |kF = np − k · ep δ(kF − p)

and

np (1 − np+k ) = np {1 − np + k · ep δ(kF − p)} = np k · ep δ(kF − p) ∼ k .

Replace now the denominator for small k by 2kkF and perform the sum over p and q in polar coordinates. Finally the sum over k is to be performed over an 2 expression which contains 1/k4 from vk , 1/k from the denominator, and 3 k from the 2 numerator which together with k from integration in k-space leads to dk/k = ln k.

366

Solutions

4.8: (a) The meaning of c†α (cα ) of creating(annihilating) a fermion in the state α with the probability amplitude ψα (r) implies, that Ψ† (r)(Ψ(r)) creates(annihilates) a fermion at r. (b) Write ∗ ∗ ψα (r)ψα′ (r ′ ) = δ(r − r ′ ) ψα (r)ψα′ (r ′ ) {c†α , cα′ } = {Ψ† (r), Ψ(r ′ )} = ′ α α,α

δα,α′

and similar for {Ψ(r, Ψ(r ′ )} = {Ψ† (r, Ψ† (r ′ )} = 0. (c) For free electrons the density operator is n ˆ (r) = Ψ† (r)Ψ(r) =

e−i(k

′

−k)·r

k ′ ,k

=

q

e−iq·r

1 † ck+q ck V k

→

1 † c ′ ck V k n ˆq =

and with

q = k′ − k

1 † ck+q ck , V k

† n ˆ0 = c c /V = n ˆ is the operator of particle density (its eigenvalue being n) k †k k ˆ q describes density fluctuations. d) The Coulomb interaction while q ck+q ck = n can be rewritten with c†p+kσ c†q−kσ′ cqσ ′ cpσ = −c†p+kσ δσ,σ′ δq−k,p − cpσ c†q−kσ′ cqσ ′

by applying the fermion commutation rules and by using the number operators the Coulomb interaction becomes 1 1 ? ? vk Nk N−k − N . vk c†p+kσ c†q−kσ′ cqσ ′ cpσ = 2 k=0,p,q 2 k=0 σ,σ′

4.9: Replace in the given expression ?q (τ ), N ?−q (0)] exact → [N ?q (τ ), N ?−q (0)] 0 [N

and

?q

Next → Next + N

?q . From Sect. 4.5, take with the induced number fluctuation N ∞ 0 / e2 1 ?q (τ ), N ?−q (0) 0 = vq π0 (q, ω) dτ eiωτ −Γτ N lim 2 Γ→0 ε0 V q i¯ h 0 to write

?q + Next ?q = vq π0 (q, ω) N N

and identify with

or

?q = N

Next vq π0 (q, ω) 1 − vq π0 (q, ω)

?q

N 1 =1+ ⇒ εRPA (q, ω) = 1 − vq π0 (q, ω) . ε(q, ω) Next 4.10: Using relations given in Sect. 4.6, the lhs of the given equation can be written ∞ ∞ 1 1 ?q |Ψm |2 = −vq dωωIm dωωδ(ω − ωm0 ) . |Ψ0 |N ε(q, ω) h m ¯ 0 0

Solutions

367

After showing (by inserting a complete set of eigenstates |Ψm of Hjell ) that ?q |Ψm |2 = Ψ0 |[[Hjell , N ?q ], N ?−q ]|Ψ0

2¯ hωm0 |Ψ0 |N − m

and one has to evaluate the double commutator. This is done with Hjell written in terms of density fluctuations (see (4.107) and Problem 4.8). Realize first ?† = N ?q . ?−k and [N ?k , N ?k ′ ] = 0 that the interaction term commutes with N with N k Evaluate ?q ] = ?q ] = [ ǫk c†k ck , N ǫk [c†k ck , c†q ′ −q cq ′ ] [Hjell , N k,q ′

k

which by applying fermion commutation rules yields † ?q ] = ǫk ck ck+q − c†k−q ck = (ǫk − ǫk+q ) c†k ck+q [Hjell , N k

k

and with ǫk = ¯ h2 k2 /2m

h2 q 2 h2 ¯ ?q ] = − ¯ nq − k · q c†k ck+q [Hjell , N 2m m k

The first term ∼Nq does not contribute to the double commutator which, therefore, reads 0 / h2 ?q ], N ?−q = − ¯ k · q[c†k ck+q , c†k ′ +q ck ′ ] . [Hjell , N m ′ k,k

Evaluating the commutator with the rules for fermion operators leads to /

0 h2 ?q ], N ?−q = − ¯ [Hjell , N k · q c†k ck − c†k+q ck+q m

and with

k

=−

k+q →k

h2 ¯ ¯2 † h ck ck = − q 2 N . m m k

Consider the factors one has ∞ 1 πvq dωωIm =− ε(q, ω) h ¯ 0

−

1 2¯ h

!

−

¯2 2 h q N m

!

1 = − ωp2 . 2

4.11: Evaluate for ω = 0 ε1 (q, 0) = 1 −

e2 fk+q − fk ε0 V q 2 k,σ ǫk+q − ǫk

fk+q − fk ≃ q · ∇k fk = q · ∇k ǫk and Thus one has

ǫk+q − ǫk ≃ q · ∇k ǫk

by writing for

q → 0:

df ≃ −q · ∇k ǫk δ(ǫk − EF ) dǫ

368

Solutions e2 δ(ǫk − EF ) ε0 V q 2 k,σ ∞ 2V e2 dk k2 δ(ǫk − EF ) 4π = 1+ ε0 V q 2 (2π)3 0 !3/2 e2 1 2m 1/2 = 1+ EF and ε0 q 2 2π 2 ¯ h2 2 e2 1 2m kFT 3ne2 2 . with kFT = ε1 (q) ≃ 1 + 2 kF = 1 + 2 2 2 ε0 q 2π ¯ q 2ε0 EF h

ε1 (q) ≃ 1 +

The meaning of kFT becomes evident when looking at the screened Coulomb interaction vq /ε1 (q, 0) whose Fourier transform is of the form exp (−kFT r)/r: 1/kFT is the Thomas–Fermi screening length. 4.12: Use the pair–distribution function (1.11) g(r) = 1 +

1 iq·r e (S(q) − 1) N q

and use

N npσ = V p,σ

and

1 npσ (1 − np+qσ ) to write N p,σ . 1 iq·r V e npσ (1 − nq +q,σ ) − 1 g(r) = 1 + NV q N p,σ

S(q) =

= 1−

V2 npσ np+qσ eiq·r . N 2 q,p,σ

The summation can be carried out by writing ′ ··· = nq ′ eiq ·r . np+q ei(p+q)·r = 2 np e−ip·r np e−ip·r q

p,σ

p

q′

This double sum with the occupation factors was carried out already in Sect. 4.4 for T = 0 K by integrating over the Fermi sphere and yields g(r) = 1 − ρHF /en (see Fig. 4.13). Solutions for Chap. 5 5.1: Taking spin into account, the expectation value of HN with the Slater determinant ΨN is written ΨN |HN |ΨN =

N

dxψα∗ (x)

α=1

+

! p2 + V (r) ψα (x) 2m

N 1 dxdx′ ψα∗ (x)ψβ∗ (x′ )v(r − r ′ )ψα (x)ψβ (x′ ) 2 α,β=1 α=β

−

N 1 dxdx′ ψα∗ (x)ψβ∗ (x′ )v(r − r ′ )ψβ (x)ψα (x′ ). 2 α,β=1 α=β

Solutions

369

Carrying out the summation over spin variables, the first two terms become identical with (5.8), while the third term (which appears because the Slater determinant is an antisymmetrized product of N single-particle wave functions) contributes only if ψα and ψβ are states with the same spin. The variational principle leads for the first two terms to the Hartree equations, which become modified by a contribution from the third term, the exchange term (5.11). √ 5.2: For free electrons with ψk (r) = exp (ik · r)/ V the averaged exchange density reads 1 ik ′ ·(r−r ′ ) −ik·(r−r ′ ) 2 n ¯ HF (r, r ′ ) = − e e N V ′ k,k |k|,|k′ |≤kF

and with the Fourier transform of 1/|r − r ′ | the exchange potential becomes 1 1 i(k ′ −k+q)·(r−r ′ ) 3 ′ 2e2 e d r Vx,Slater (r) = − ε0 N V q2 V V q k,k′ |k|,|k′ |≤k F

2

=−

2e ε0 N V

k,k′ |k|,|k′ |≤kF

δk−k′ ,q

1 3 = − 2 e 2 kF . |k − k′ |2 8π ε0

With the exchange energy ǫx (n) = −3e2 kF /16π 2 ε0 from Sect. 4.4 one finds VxLDA (r) =

2 4 LDA (n(r)) = Vx,Slater (r). ǫ 3 x 3

5.3: The number of discrete k = (k1 , k2 , k3 ) with ki = 2πni /Li , i = 1, 2, 3 with 0 ≤ ni < Ni in a Brillouin zone is N = N1 N2 N3 , which is the number of unit cells in the crystal or periodicity volume. Thus, for each electron in the unit cell with given spin there is one state in the energy band, i.e. each band can accommodate 2N electrons. 5.4: A point at the Brillouin zone boundary is characterized by the relation k ′ = k − G. The condition of degeneracy is k2 = k′2 , thus (k − G)2 = (k′ )2 becomes 2k · G = G2 , which is the condition for Bragg reflection.

5.5: The primitive reciprocal lattice vectors of the square lattice are b1 = (1, 0)2π/a, and b2 = (0, 1)2π/a. Write the free electron energies !2 h2 ¯ h2 2π ¯ 2 E(k) = κ2 (k + G) = 2m 2m a

for the smallest G at the points Γ, M , and X and connect corresponding points by parabolas defined by G. If ν = 1, 2, 3 is the number of electrons per atom then, for one atom per unit cell, ns =√ν/a2 is the areal electron density. The radius of the Fermi circle is given by kF = 2πns and the Fermi energy by !2 h2 2π ¯ h2 2 ¯ ν EF = kF = 2m 2m a 2π or κ2F = ν/2π which is 0.159 for ν = 1.

370

Solutions

5.6: Proceed as in Problem 5.5 and see [121] for the free-electron bands along Γ − L. 5.7: In the almost free-electron picture, the energy bands of Al, Si, and GaAs derive from the free-electron bands of the fcc lattice (see Problem 5.6). Due to the diﬀerent crystal structures (Bravais lattice for Al, diamond structure for Si, zinc blende structure for GaAs) the energy gaps are determined by diﬀerent Fourier components of the pseudo-potential: (τ ) (τ ) vpsp (r − R0n − τ ) → Vpsp (G) = e−ig·τ vpsp (G). Vpsp (r) = n,τ

For Al with τ = 0, the structure factor S(G) = τ exp (−iG · τ ) equals 1 for all G. For diamond and zinc blende with τ = ±τ ′ with τ ′ = (1, 1, 1)a/8 one has ′

′

Vpsp (G) = e−iG·τ v (+) (G) + eiG·τ v (−) (G) = cos (G · τ ′ )vS (G) − i sin (G · τ ′ )vA (G), where vS (G) = v (+) (G) + v (−) (G)

and

vA (G) = v (+) (G) − v (−) (G).

In Si the anti-symmetric potential vA (G) vanishes. Thus, Fourier components at diﬀerent reciprocal lattice vectors determine the energy bands as Al, Si, GaAs. Especially, for G = (2, 0, 0)2π/a we have cos (G · τ ) = 0 but sin (G · τ ) = 1 and the anti-symmetric potential present in GaAs removes the degeneracy of the level X1 in Si (see Fig. 5.11). 5.8: The crystal field splitting is determined by the matrix formed by 3 ∗ d rφν ′ (r)v(r − R0n )φν (r) Kν ′ ν = n

2

2

2

with ν = xy, yz, zx, 3z − r , x − y 2 . The point group operations of the cubic lattice turn the coordinate triple x, y, z into any other permutation including sign changes of x, y, and z, while leaving n v(r − R 0n ) invariant. Thus, the groups of orbitals dxy , dyz , dzx and d3z 3 −r2 , dx2 −y 2 form invariant sets under the cubic point group, which are classified by the irreducible representations Γ25′ and Γ12 , respectively, and the matrix with the elements Kν ′ ν has block-diagonal form. Further inspection shows, that each of the diagonal blocks is itself diagonal for the given basis with identical diagonal matrix elements, thus, the crystal field splitting gives a threefold (Γ25′ ) and a twofold (Γ12 ) state as can be seen at the Γ point of the band structures depicted in Figs. 5.14 and 5.15. There is no diﬀerence between sc, bcc, and fcc crystal structure because they have the same point group. 5.9: The overlap matrix Sν ′ ν (k) is hermitian and can be diagonalized by a unitary transformation U : U SU −1 = S ′ with Sµ′ ′ µ = Sµ′ δµ,µ′ . The diagonal elements Sµ′ of the transformed overlap matrix represent the norms of the new basis states which are always positive. Thus the eigenvalue equation can be rewritten −1 −1 U UC = E U UC . SU HU H′

C′

S′

C′

One can multiply this equation with the inverse square root of the diagonal matrix S ′ to arrive at the eigenvalue equation

Solutions

371

S ′−1/2 H ′ S ′−1/2 S ′1/2 C ′ = ES ′1/2 C ′ ˜ = S ′−1/2 H ′ S ′−1/2 , C˜ = S ′1/2 C ′ one has the standard eigenvalue and with H problem with the secular equation $ $ $ $˜ $Hµ′ µ − Eδµ′ ,µ $ = 0 ˜ = S ′−1/2 U H ′ U −1 S ′−1/2 . with H

5.10: The nearest neighbors in the sc crystal structure are R0n : a(±1, 0, 0), (0, ±1, 0), (0, 0 ± 1) leading to the dispersion Es (k) = Es + 2Jss (a)(cos (kx a) + cos (ky a) + cos (kz a)) which for k = (k, 0, 0) becomes Es (k) = Es + 2Jss (a) cos (ka) and for k = (k, k, k), Es (k) = Es + 6Jss (a) cos (ka) with band widths E(0) − E((π/a, 0, 0)) = 4Jss (a) and E(0) − E((π/a, π/a, π/a)) = 12Jss (a), respectively. For the bcc crystal structure one has R 0n :

a (±1, ±1, ±1), (∓1, ±1, ±1), (±1, ∓1, ±1), (±1, ±1, ∓1) 2

leading to the dispersion ! a a a & Es (k) = Es + 2Jss √ cos (kx + ky + kz ) + cos (−kx + ky + kz ) 2 2 3 a ' a (kx − ky + kz ) + cos (kx + ky − kz ) + cos 2 2 √ which for k = (k, 0, 0) becomes √ Es (k) = Es + 8Jss (a/ 3) cos (ka/2) and for k = (k, k, k), Es (k) = Es + 2Jss (a/ √3)(cos (3ka/2) + 3 cos (ka/2)) with band widths √ E(0) − E((2π/a, 0, 0)) = 16Jss (a/ 3) and E(0) − E((π/a, π/a, π/a)) = 8Jss (a/ 3).

5.11: For a solution see P.R. Wallace, Phys. Rev. 71, 622 (1947) and the article by S.E. Louis in [114].

5.12: Use the Peierls substitution ǫ(k) → H(p − eA)2 with the vector potential A = (0, B(x cos θ + z sin θ, 0)) corresponding to B = B(sin θ, 0, cos θ) to write H=

p2x 1 p2 + (py − eB(x cos θ + z sin θ))2 + z . 2mt 2mt 2ml

The equations of motion for the components of the momentum are (up to terms ∼py which vanish later due to p˙ y = 0) e2 B 2 i p˙ x = − [px , H] = − (x cos θ + z sin θ) cos θ h ¯ mt e2 B 2 i p˙ z = − [pz , H] = − (x cos θ + z sin θ) sin θ. h ¯ mt Take the derivatives of these equation with respect to t and replace x˙ = px /mt , z˙ = pz /ml to obtain

372

Solutions −¨ px = ωt2 cos2 θpx + ωt ωl sin θ cos θpz −¨ pz = ωt2 sin θ cos θpx + ωt ωl sin2 θpz

with ωl,t = eB/ml,t . With px,z ∼ exp (−iωt) this becomes a set of homogeneous linear equations and the eigenfrequencies follow from $ 2 $ $ ωt cos2 θ − ω 2 ωt ωl sin θ cos θ $ $ 2 $ 2 2 $ ωt sin θ cos θ ωt ωl sin θ − ω $ = 0 with the nontrivial solution

ω 2 = ωt2 cos2 θ + ωt ωl sin2 θ = e2 B 2

sin2 θ cos2 θ + 2 mt ml mt

!

.

The expression in the bracket is the squared inverse cyclotron mass for the anisotropic energy surface if the magnetic field includes the angle θ with the z-axis. See [4]. 5.13: For k = (kx , ky , 0) and ∇V (001) the interface spin–orbit (or Rashba)-term reads HSO (k) = α|∇V |(ky σx − kx σy ) and with k± = kx ± iky = k exp (±iϕ) and σ± = (σx ± iσy )/2

HSO (k, ϕ) = iα|∇V |(k+ σ− − k− σ+ ) = iα′ (eiϕ σ− − e−iϕ σ+ ). The subband Hamiltonian with spin–orbit interaction becomes ! iα′ e−iϕ ǫk . H(k, ϕ) = −iα′ e−iϕ ǫk Its eigenvalues ǫ± (k) = ǫk ± α|∇V |k do not depend on ϕ and are two parabolas shifted against each other. Use the eigenvectors " 1 " 1 |k, ± = √ "" iϕ 2 ∓ie to calculate the expectation value of the vector of Pauli spin matrices: k, ±|σ|k, ± = ±(ex sin ϕ − ey cos ϕ). Thus, the spin is always oriented perpendicular to the wave vector k = (kx , ky , 0) and rotates with ϕ. Note, that the states on each parabola form a Kramers pair. See [153]. Solutions for Chap. 6 6.1: Choose the quantum numbers for the Bloch states ¯ σ , α′ = n′ k′ σ ′ , β ′ = n ¯′σ α = nkσ, β = n ¯ k¯ ¯′k ¯′. After carrying out the summation over spin variables the matrix element reduces to e2 3 ∗ ′ ′ (r)ψn∗¯ k¯ σ¯ (r ′ ) ψ ′ ¯ ′ (r)ψn′ k ′ σ¯ (r ′ ). Vαββ α = d r d3 r ′ ψnkσ κ|r − r ′ | n¯ k σ

Solutions

373

Decompose the Bloch function into plane wave and lattice periodic parts, expand the products of periodic parts with the same argument in a Fourier series iG·r Bn¯nk kσ , u∗nkσ (r)un¯ ′ k¯′ σ (r) = ¯ (G)e G

and use the Fourier transform of the Coulomb interaction to perform the integration over the space variables to find ′ Bn¯n′ k k¯ ′ σ (G)Bn¯ n′ kk Vαββ ′ α′ = ¯ ′σ ¯ (G ) G,G ′ ,q

×

e2 ε0 V q 2

¯ ′ −q−G)·r

d3 re−i(k−k

¯

d3 r ′ e−i(k−k

′

+q−G ′ )·r ′

.

For the single-band approximation set n = n ¯=n ¯ ′ = n′ and with only leading terms ′ in the Fourier series, G = G = 0, write ′ Bn¯n′ k k¯ ′ σ (G) → Bkq , Bn¯ n′ kk ¯ ¯ ′σ ¯ (G ) → Bkq

and obtain Vαββ ′ α′ =

q

v(q)V 2 Bkq Bkq ¯ δk ¯ ′ −q . ¯ ′ ,k−q δk,k

6.2: To check normalization and orthogonality write 3 3 ∗ d rφnσ (r − R)φn′ σ ′ (r − R′ ) s

=

−ik·R−ik ′ ·R ′ 1 e δσ,σ′ N ′ k,k

∗ d3 r ψnkσ (r)ψn′ k ′ σ (r) δn,n′ δk,k′

1 −ik·(R−R ′ ) = e δn,n′ = δn,n′ δR,R ′ N k δR,R′

Thus localized functions, represented by Bloch functions, are orthogonal when centered on diﬀerent sites. Insofar the Wannier representation diﬀers from the LCAO. Introduce fermion operators for these localized Wannier states by 1 ik·R e cnkσ cnRσ = √ N k and calculate the anti-commutator ' ' & 1 ik·R−ik ′ ·R ′ & e cnkσ , c†n′ k ′ σ ′ cnRσ , c†n′ R ′ σ ′ = N ′ k,k

δn,n′ δk,k′ δσ,σ′

1 ik·(R−R ′ ) e = δn,n′ δσ,σ′ δR,R ′ . = δn,n′ δσ,σ′ N k

The other anti-commutators yield

374

Solutions &

c†nRσ , c†n′ R ′ σ ′

'

= {cnRσ , cn′ R ′ σ ′ } .

6.3: Use the representation (6.7) of the Bloch functions to write Vαββ ′ α′ =

1 N2 ×

¯

¯ ′ ·R ′

e−ik·R 1 e−ik·R 2 eik

1

eik

′

·R ′2

R 1 ,R 2 ,R ′1 ,R ′2

d3 r

d3 r ′ φ∗ (r − R1 )φ∗ (r ′ − R2 )

e2 φ(r − R′1 )φ(r ′ − R′2 ). κ|r − r ′ |

Use (6.8) with the exponentials and sums over the wave vectors to replace the fermion operators of Bloch states in the interaction term by those of Wannier states to find (6.10). 6.4: The commutator between c†↑ c↓ and c†↓ c↑ [c†↑ c↓ , c†↓ c↑ ] = c†↑ c↓ c†↓ c↑ − c†↓ c↑ c†↑ c↓ is evaluated by using the anti-commutator for fermion operators and writing c†↑ c↓ c†↓ c↑ = c†↑ (1 − c†↓ c↓ )c↑ = c†↑ c↑ − c†↑ c†↓ c↓ c↑

= c†↑ c↑ − c†↓ (1 − c↑ c†↑ )c↓ = c†↑ c↑ − c†↓ c↓ + c†↓ c↓ c↑ c†↑ ,

where the last term cancels in the commutator. Thus we have [c†↑ c↓ , c†↓ c↑ ] = c†↑ c↑ − c†↓ c↓ . The other commutation relations [c†↑ c↓ , c†↑ c↑ − c†↓ c↓ ] = −2c†↑ c↓ ,

and

[c†↓ c↑ , c†↑ c↑ − c†↓ c↓ ] = 2c†↓ c↑

follow in a similar way. These three commutators are the same as those between S ± and S z and correspond to the operator algebra of the Cartesian components of the angular momentum S x , S y , and S z . 6.5: Evaluate the commutator [Siβ Skβ , Sjx ] = Skβ Sjx Siβ − Sjx Siβ Skβ

= −iεxβγ Sjγ Skβ δi,j − Siβ iεxβγ Sjγ δj,k ,

to write the commutator of Sjx with the first term of the Heisenberg hamiltonian + β β x − k=j β,γ Jjk iεxβγ Sjγ Skβ for i = j, k = j [Si Sk , Sj ] = Jik − i=j β,γ Jij iεxβγ Siβ Sjγ for k = j, k = i i,k β i=k

= −i

i=j

Jij (S i × S j )x

where for obtaining the last line use was made of i = k and of the meaning of the Levi–Civita symbol. For the commutator with the second term of the Heisenberg hamiltonian evaluate z x iεzxβ Sjβ δi,j = iSjy . [Si , Sj ] = i

i,β

Solutions

375

Putting together both contributions find dSjx i 1 = Jij S i × S j x + gμB H ext × S j x . Hspin , Sjx = − dt h ¯ h ¯ i 6.6: Evaluate the commutator 0 0 / / αk , α†k ′ = uk b1k − vk b†2k , uk ′ b†1k ′ − vk ′ b2k ′ / 0 0 / = uk uk ′ b1k , b†1k ′ +vk vk ′ b†2k , b2k ′ = u2k − vk2 δk,k ′ , δk,k′

which for calculate

2

u2k − vk

−δk,k′

= 1 is a boson commutation relation. In the same way,

/ 0 [αk , βk ′ ] = uk b1k − vk b†2k , uk ′ b2k ′ − vk ′ b†1k ′ / 0 / 0 = −uk vk ′ b1k , b†1k ′ − vk uk ′ b†2k , b2k ′ = 0.

The remaining commutation relations are obtained in a similar way. 6.7: Using the new boson operators for coupled magnon–phonon modes, the hamiltonian becomes ( † p Hp−m = αk αk ¯ hωk cos2 Θk + ¯ hωkm sin2 Θk − 2ck sin Θk cos Θk k

p 2 + βk† βk h ¯ ωk sin Θk + ¯ hωkm cos2 Θk + 2ck sin Θk cos Θk hωkp − ¯ hωkm ) sin Θk cos Θk + ck cos2 Θk − sin2 Θk + α†k βk (¯ ) hωkp − ¯ . + αk βk† (¯ hωkm ) sin Θk cos Θk + ck cos2 Θk − sin2 Θk

It can be diagonalized with p hωk − ¯ ¯ hωkm sin Θk cos Θk + ck cos2 Θk − sin2 Θk = 0

or

tan 2Θk =

−2ck . hωkp − ¯ ¯ hωkm

For ωkp = ωkm = ωk we have cos2 Θk = sin2 Θk = 1/2 and find for the eigenenergies of the coupled modes hωk − ck hωkα = ¯ ¯

and

h ¯ ωkβ = ¯ hωk + ck

and for the corresponding boson operators 1 αk = √ ak − bk 2

and

1 βk = √ ak + bk . 2

376

Solutions

6.8: With the inverted transformation relations (6.59) write the terms of (6.56): b†1k b1k = α†k αk u2k + βk† βk vk2 + α†k βk† + αk βk uk vk b†2k b2k = βk† βk u2k + αk α†k vk2 + βk† α†k + βk αk uk vk b†1k b†2k = α†k αk + βk βk† uk vk + α†k βk† u2k + βk αk vk2 b1k b2k = αk α†k + βk† βk uk vk + αk βk u2k + βk† α†k vk2 to obtain

Hspin ≃ Ea + 2Ja νS +

α†k βk†

& k

+ αk βk

α†k αk + βk† βk

u2k + vk2 + 2γk uk vk

' 2uk vk + γk (u2k + vk2 ) + 2(γk uk vk + vk2 .

Add and subtract u2k under the sum to find with u2k − vk2 = 1 the expression (6.60). 6.9: Extend (6.53) by the terms due to the anisotropy field HA and the external field Hext z z Hspin = Ja S 1i · S 2j − gμB − HA − Hext S2i HA + Hext S1i n.n.ij

i

and replace the spin operators by boson operators on each sublattice using the Holstein–Primakoﬀ transformation to obtain Hspin = − 2Ja νN S 2 − 4gμB HA N S & † ' + 2Ja νS b1k b1k + b†2k b2k + γk b†1k b†2k + b1k b2k k

+ 2gμB HA

k

† b†1k b1k + b†2k b2k + gμB Hext b1k b1k − b†2k b2k . k

Eliminate the coupling between the sublattices using the Bogoliubov transformation (6.57) and write with the abbreviations C = −2Ja νN S 2 − 4gμB HA N S , A = 2Ja νS , B = 2gμB HA , C = gμB Hext & Aγk 2uk vk α†k αk + βk† βk + 1 + (u2k + vk2 ) α†k βk† + αk βk Hspin ≃ C + k

+ (A + B) (u2k + vk2 ) α†k αk + βk† βk + 2uk vk α†k βk† + αk βk + 2vk2 ' + C α†k αk − βk† βk .

Diagonalize with Aγk u2k + vk2 = −2(A + B)uk vk . Square this relation and use u2k − vk2 = 1 to get a biquadratic equation for uk . Take its solution to find u2k + vk2 = ±

(A + B)2 (A + B)2 − A2 γk2

1/2 1/2 1 A2 γk2 , uk vk = ∓ . 2 (A + B)2 − A2 γk2

Solutions

377

Identify the prefactor of the magnon number operators as the magnon energy )1/2 ( hωk = (A + B)2 − A2 γk2 ¯

which without anisotropy field (B = 0) reduces to the dispersion relation for the antiferromagnetic magnons which for small k is linear. With anisotropy field one finds a finite magnon energy for k = 0 and a quadratic dispersion for small k as can be seen in Fig. 6.8. 6.10: Evaluate for low temperature (T ≪ TC ) the magnetization (per unit volume) † M (T ) = gμB N S − bk bk

or k

1 exp (¯ h ω /k BT ) − 1 k k ∞ k2 dk 4π = gμB 3 (2π) 0 exp (Dk2 /kB T ) − 1

M (0) − M (T ) = gμB

where the quadratic dispersion for small k is used. The integral (it is a Bose integral) can be solved as described in the Appendix (A.3) to yield Bloch’s T 3/2 law: 3 gμB M (0) − M (T ) = ζ( ) 2 M (0)

kB 4πD

!3/2

T 3/2 .

See [166]. 6.11: Calculate the expectation value in (6.106) for T = 0: ( . . . 0 = Ψ0 |M+ (q, τ )|Ψm Ψm |M− (−q, 0)|Ψ0

m

) − Ψ0 |M− (−q, 0)|Ψm Ψm |M+ (q, τ )|Ψ0 .

Replace the operators M± by the fermion operators and extract the exponentials with the time dependence. The energy diﬀerence Em − E0 of exact eigenstates becomes in HF approximation the energy diﬀerence of single-particle energies for particle–hole excitations with spin-flip across the Fermi energy. By manipulations as in Sect. 4.5 one arrives at the spin susceptibility (6.108). To evaluate (6.108), use ǫk+q↑ − ǫk↓ = −∆ +

¯2 h ¯2 h k · q + ǫq , ǫk↑ − ǫk−q↓ = −∆ + k · q − ǫq m m

with ǫq = ¯ h2 q 2 /2m and perform the sum over k with the substitution k · q = kq cos θ , cos θ = x. The first term can be written A↓ = =

|k|≤kF↓

V (2π)2

1 hω + ǫk+q↑ − ǫk↓ ¯

kF ↓ 0

dk k2

+1 −1

1 dx. hω − ∆ + ǫq + ¯ ¯ h2 kqx/m

378

Solutions

The integral over x can be solved according to " " +1 1 a+b dz 1 " 1 + b/a "" dx . = = ln "" b a−b z b 1 − b/a " −1 a + bx

Looking for collective excitations at small q one has b/a ≪ 1 and can use the expansion + + " " !3 , !2 , 2 b 1 "" 1 + b/a "" 2 1 b 1 b ≃ = ln " + 1 + b 1 − b/a " b a 3 a a 3 a

to obtain

2V 1 A↓ = (2π)2 ¯ hω − ∆ + ǫq

−

kF 0

dk k

2

+

k2 1+ 3

¯2q h m

!2

1 hω − ∆ + ǫq ¯

!2 ,

.

After integration and corresponding evaluation of the second term A↑ one arrives at (6.108). Solutions for Chap. 7 † † 7.1: Evaluate the commutator [ck,¯ ¯ σ , ck+qσ ck ′ −qσ ′ ck ′ σ ′ ckσ ] by successively interwith the four operators appearing in the interaction term. Each step changing ck,¯ ¯ σ leads to a change in sign and in addition gives a Kronecker δ for exchange with the creation operators, thus † † † δσ¯ ,σ c†k ′ −qσ ′ ck ′ σ ′ ckσ − δk,k ck,¯ ¯ σ ck+qσ ck ′ −qσ ′ ck ′ σ ′ ckσ = δk,k+q ¯ ¯ ′ −q ck+qσ ck ′ σ ′ ckσ .

Interchange the last two operators in the second term and replace σ by σ ′ and k by k ′ . Replace in the first term q by −q and consider the properties of the interaction matrix element to find the expression for [ck¯ , Hint ] (7.20). The commutator [c†k¯ , Hint ] is evaluated by analogous steps. 7.2: The equation of motion of the full Green function G(kσ; t − t′ ) (7.21) can be written with (7.23) as ! ∂ i¯ h − ǫkσ G(kσ; t − t′ ) = δ(t − t′ ) + dt′′ Σ(kσ; t − t′′ )G(kσ; t′′ − t′ ). ∂t Replace the Green function and the self-energy by their Fourier transform with respect to the time arguments and identify the integrands to obtain (¯ hω − ǫkσ ) G(kσ; ω) = 1 + Σ(kσ; ω)G(kσ; ω). After multiplication with the Green function of the noninteracting system hω − ǫkσ )−1 G0 (kσ; ω) = (¯ one arrives at the Dyson equation G(kσ; ω) = G0 (kσ; ω) + G0 (kσ; ω)Σ(kσ; ω)G(kσ; ω). 7.3: The commutator of ni−σ ciσ with the single-particle part H0 yields three terms due to interchange of creation and annihilation operators:

[ni−σ ciσ , H0 ] =

mjσ ′

=

mjσ ′

tmj c†i−σ ci−σ ciσ c†mσ ′ cjσ ′

Solutions − c†mσ ′ cjσ ′ c†i−σ ci−σ ciσ

379

tmj δim δσσ′ ni−σ cjσ − δim δσ−σ ′ c†i−σ ciσ cj−σ

− δij δσ−σ ′ c†m−σ ci−σ ciσ = tim ni−σ cmσ + c†i−σ cm−σ ciσ − c†m−σ ci−σ ciσ , m

where the last line is obtained by proper choice of the summation indices and rearranging the operators. Similarly the commutator with the interaction term is evaluated 1 (ni−σ ciσ nmσ ′ nm−σ ′ − nmσ ′ nm−σ ′ ni−σ ciσ ) [ni−σ ciσ , H1 ] = U 2 mσ ′ 1 ni−σ ciσ c†mσ ′ cmσ ′ nm−σ ′ − nmσ ′ nm−σ ′ ciσ = U 2 mσ ′ 1 = U ni−σ (δim δσσ′ ciσ ni−σ + δim δσ−σ ′ ni−σ ciσ ) , 2 ′ mσ

n2i−σ

which using = ni−σ becomes (7.55). 7.4: Calculate the derivative of the self-energy (7.70) and obtain the spectral weight 2 !−1 E − ǫ0 − U (1 − n−σ ) ∂Σ = . Z(E) = 1 − 2 ∂E E − ǫ0 − U (1 − n−σ ) + U 2 n−σ (1 − n−σ ) For the lower Hubbard band with E = ǫ0 + 2t cos ka this becomes 2 2t cos ka − U (1 − n−σ ) Zlower = 2 2t cos ka − U (1 − n−σ ) + U 2 n−σ (1 − n−σ )

and reduces for k = π/2a to Zlower = 1 − n−σ . For the upper Hubbard band with E = ǫ0 + U + 2t cos ka the result is 2 2t cos ka + U n−σ

Zupper = 2 2t cos ka + U n−σ + U 2 n−σ (1 − n−σ )

and Zupper = n−σ for k = π/2a. Note that the spectral weights for k = π/2a are those of the atomic limit (see Fig. 7.3). 7.5: This problem is solved in some detail in [169, 206]. It leads to the so-called t − J model, where t is the hopping integral and J ∼ t2 /U is the eﬀective exchange matrix element. 7.6: Describing the delocalized electrons by fermion operators c†kσ , ckσ in a band with dispersion ǫk and the localized electrons with energy ǫd by fermion operators d†σ , dσ , the Hamiltonian is formulated as H = ǫk c†kσ ckσ + ǫd d†σ dσ k,σ

+

k,σ

Vkσ

σ

d†σ ckσ

+ c†kσ dσ + U nd↑ nd↓ .

380

Solutions

Here the third term describes the hybridization between the localized and delocalized electrons. This is the Anderson impurity model. It can be extended by considering instead of a single impurity a periodic configuration of sites i with d electrons. For this case the fermion operators for d electrons become d†iσ , diσ and summation over the sites i is to be considered. See P.W. Anderson: Phys. Rev. 124 41 (1961) and [64] Sect. 6.2, [122, 195]. 7.7: For the solution see [122, 195]. It uses a contour integration in the complex frequency plane taking into account the position of the poles for quasi-particles and holes. 7.8: Start from (4.136) for the real part of the dielectric function, which for T = 0 reads 2ǫk − ǫk+q − ǫk−q ε1 (q, ω) = 1 + vq . (¯ hω − ǫk + ǫk−q )(¯ hω − ǫk+q + ǫk ) |k|≤kF

h2 k2 /2m the numerator simplifies Making use of the free electron dispersion ǫk = ¯ 2 2 to −¯ h q /m while the denominator takes the form (¯ hω − ǫk + ǫk−q )(¯ hω − ǫk+q + ǫk ) =

hω − ¯

¯2 h k·q m

!2

−

¯2 2 h q 2m

!2

and the expression to be evaluated by integration is !2 1 1 2m , 2 2 = 2 2 − 2k · q)2 − q 4 h ¯ h ¯2 h ¯2 2 (q s hω − m k · q − 2m q |k|≤kF ¯ |k|≤kF

hω/¯ h2 . with qs2 = 2m¯ For d = 1 the vectors become scalars and the sum can be written as the integral (using the substitution x = 2kq) kF xF dk dk L L = . I(q, ω) = 2π −kF qs4 − q 4 − 4qs2 kq + 4k2 q 2 2π −xF x2 − 2qs2 x + qs4 − q 4 The integral can be found in integral tables giving " " " (2kF q − 2q 2 )2 − 4qs4 " L 1 ". " ln I(q, ω) = 2π 4q 3 " (2kF q + 2q 2 )2 − 4qs4 "

For ω = 0 (or qs = 0) this integral diverges for q → kF . For d = 3, considering from the beginning the simpler case ω = 0, the corresponding integral is " " kF +1 kF "k − q" L3 L3 d cos ϑ " " I(q) = dk k 2π = dk ln " k + q ", 2 2 2 4 (2π)3 8π 2 q 3 0 −1 4k q cos ϑ − q 0

and yields by integration ! L3 kF − 1 ln I(q) = 8π 2 q 2 q

! kF −1 − q

For q → kF this integral remains finite.

! kF + 1 ln q

! kF +1 . q

Solutions

381

Solutions for Chap. 8 8.1: The operator for linear electron–phonon interaction " ∇l v(r l − Rnτ )"R 0 · unτ Hel−ph = − nτ

l,n,τ

can be written in terms of phonon and electron operators using (3.22) together with (3.39) and by replacing with (4.76) " ∇l v(r l − Rnτ )"R 0 → nk|∇v|n′ k ′ c†nk cn′ k ′ . nτ

l

n,k,n′ ,k ′

The sum over the lattice sites n eﬀects only the gradient of the potential and has the form of a Bloch function iq·R 0 " n ∇v(r − R " = eiq·r Uqτ (r) e nτ ) R 0 n

nτ

where Uqτ (r) is a lattice periodic function. After decomposition of the electron Bloch functions in the matrix element into plane wave and lattice periodic part the product u∗nk (r)un′ k ′ (r)Uqτ (r) can be expanded in a Fourier series. Thus the integration over the crystal volume can be carried out with ′ d3 r e−i(k−k −q−G)·r ∼ δk,k ′ +q+G and yields the relation between the involved wave vectors. 8.2: Using (3.76) one has ∂ui /∂xi = ǫii and can write ∇ · u(r) =

∂ ui (r) = ǫii = Trǫ. ∂x i i i

For u(r) ∼ eiq·r follows ∇ · u = iq · u which diﬀers from zero only for longitudinal phonons. 8.3: For transverse phonons with q = α qα eα , u = β uβ eβ we have in general q·u = α qα uα = −iTrǫ = 0 but e.g. for the special case q = (q, 0, 0), u = (0, u2 , u3 ) the strain tensor components are ǫ11 = ǫ22 = ǫ33 = ǫ23 = 0, but ǫ12 =

i i qu2 , ǫ13 = qu3 , 2 2

giving a nonvanishing contribution to the electron–phonon interaction. 8.4: Start with the classical expression for the interaction energy (8.23) with 1 ∇ · P (r) = − e14 |εijk | (qi qk uj + qi qj uk ) eiq·r double index summation 4 = −2e14 (qy qz ux + qz qx uy + qx qy uz ) eiq·r .

382

Solutions

Using (8.24) for the charge density, expressing the displacement field by phonon operators, and integrating over the crystal volume converts Eint into the operator 9 qx qy esz (q) + c.p. hN ¯ 2ee14 P Hel−ph = − ε0 ε∞ s,q 2M ωs (q) q2 c†k+q ck . × a†s (−q) + as (q) k

8.5: To be calculated is the expression τ (8.71) with the given dependencies of τ . The integrals are solved with the substitution ǫ/kB T = x and we may write ∞ ∞ x3/2 e−x dx ∼ T 5/2 . ǫ3/2 e−ǫ/kB T dǫ = (kB T )5/2 0

0

Similar for τ ∼ T −1 ǫ−1/2 ∞ τ ǫ3/2 e−ǫ/kB T dǫ = T −1 (kB T )2 0

and for τ ∼ T −1 ǫ1/2 ∞ τ ǫ3/2 e−ǫ/kB T dǫ = T −1 (kB T )3 0

∞ 0

∞ 0

xe−x dx ∼ T

xe−x dx ∼ T 2 .

Thus for deformation potential coupling μ(T ) ∼ T −3/2 and for piezoelectric coupling μ(T ) ∼ T −1/2 . This explains the diﬀerent slopes of the corresponding graphs in Fig. 8.5. For more details see [4, 246]. 8.6: For the solution see [4], Chap. 7. The calculation is essentially the same as for the Fr¨ ohlich coupling but with the 1/q dependence of the interaction potential replaced √ by a q dependence. The number of virtually excited phonons being proportional to the square of the eﬀective mass of the electron turns out to be much smaller than 1. Replacing the electron by the much heavier ion would increase this number to a value much larger than 1 indicating that the perturbation calculation is not appropriate. 8.7: We have to evaluate the commutator / 0 [Hel−ph , S] = Vq Vq ′ c†k+q ck a†−q + aq , c†k ′ +q ′ ck ′ αa†−q ′ + βaq ′ k,q k′ ,q ′

and write 0 / [. . . , . . . ] = a†−q + aq , αa†−q ′ + βaq ′ c†k+q ck c†k ′ +q ′ ck ′ / 0 + c†k+q ck , c†k ′ +q ′ ck ′ αa†−q ′ + βaq ′ a†−q + aq .

The electron part of the first term can be rearranged giving

ˆ k′ c†k+q ck c†k ′ +q ′ ck ′ = c†k+q c†k ′ +q ′ ck ′ ck + δk,k ′ −q n where the first term has the structure of an electron–electron interaction (this is the one we are looking for). The phonon part of the first term gives

Solutions /

383

0

a†−q + aq , αa†−q ′ + βaq ′ = (−α + β)δq,q ′ .

The second term, being bilinear in the phonon operators, is neglected as well as the term with the electron number operator to obtain the approximate form of the commutator [Hel−ph , S]. ¯ relative to the ground 8.8: Using (8.120) the ground state expectation value of H state of the normal system is ¯ + |E(k)| E0 = H

k

and can be written with σkz = cos θk , σkx = sin θk as E0 = − With

k

and

Veﬀ sin θk ′ = E(k) tan θk 4 ′ k

1 E(k) cos θk + tan θk sin θk 2

E(k) tan θk sin θk =

2∆2 Veﬀ

and

cos θk =

!

+

k

|E(k)|.

E(k)

+ ∆2 )1/2 and by replacing the sum over k by an energy integral over the shell with thickness hωD at the Fermi energy, one finds ¯ h¯ ωD ∆2 ǫ2 − , dǫ ǫ − 2 E0 = 2D(EF ) 2 1/2 (ǫ + ∆ ) Veﬀ 0 (E 2 (k)

where D(EF ) is the density of states at the Fermi energy. The integration can easily be performed and yields ⎧ ⎫ !2 .1/2 ⎬ ⎨ ∆ 2 hωD ) hωD )2 Veﬀ . 1− 1+ ≃ − (D(EF )¯ E0 = D(EF )(¯ ⎩ ⎭ hωD ¯

The last expression, obtained for 1 ≫ D(EF )Veﬀ , tells us that the superconducting state is stable as long as the eﬀective interaction is positive. Solutions for Chap. 9 9.1: Besides the poles of G0 (E) the full Green function G(E) has additional poles for 1 − G0 (E)U = 0. In site representation U has block-diagonal form with a nonzero block U 1 = U{RI } only in the diagonal for lattice sites around the impurity. With G0 (E) written in the corresponding block form (with G0,{RI } (E) = G10 (E) and corresponding matrices for the other blocks), the matrix multiplication can be performed to give ! ! ! G10 (E) G20 (E) U 0 G10 (E)U 1 0 . = G0 (E)U = 1 3 4 3 G0 (E)U 0 G0 (E) G0 (E) 0 0 The determinant of 1 − G0 (E)U is the product of the determinants of its diagonal blocks or 1 − G0,{RI } (E)U{RI } = 0.

384

Solutions

9.2: The average of the individual resistances is r =

N 1 ri N i=1

and therefore

R =

N i=1

ri =

N i=1

ri = N r .

The variance of R is given by 2 VarR = ri rj − N 2 r 2 ri rj − N 2 r 2 = ri + ij

i

i=j

and with uncorrelated fluctuations of the ri VarR = N r 2 + N (N − 1)r 2 − N 2 r 2 = N r 2 − r 2 .

The relative variance given by

1 Varr VarR = R 2 N r 2

vanishes for N → ∞.

9.3: The transverse response reads χ(0, ω) = e2

i,f,i′ ,f ′

i|ˆ v |f f ′ |ˆ v |i′

i h ¯

∞ 0

dτ [c†i (τ )cf (τ ), c†f ′ ci′ ] 0 .

h) and the The time dependence of the operators gives a factor exp (i(ǫi − ǫf )τ /¯ integral over τ yields i 1 i ∞ dτ e h¯ (¯hω+ǫi −ǫf +iδ)τ = − , h 0 ¯ hω + ǫi − ǫf + iδ ¯ where the small parameter δ is introduced to regularize the integral. To evaluate the remaining thermal expectation value first calculate the commutator [c†i cf , c†f ′ ci′ ] = c†i ci′ δf,f ′ − c†f ′ cf δi,i′ . With the thermal expectation value of c†i ci′ 0 = f (ǫi )δi,i′ the response function takes the form of (9.87) χ(0, ω) = −e2

" "2 "i|ˆ v |f " i,f

f (ǫi ) − f (ǫf ) . hω + ǫi − ǫf + iδ ¯

9.4: Write (for finite ω) the δ function in the form δ(¯ hω + ǫi − ǫf ) = dE δ(E − ǫi )δ(E + ¯ hω − ǫf ) so that with ǫi = E and ǫf = E + ¯ hω the conductivity is expressed by "2 f (E) − f (E + ¯ hω) "" i|ˆ v |f " δ(E − ǫi )δ(E + ¯ hω − ǫf ). σ(0, ω) = πe2 dE ω i,f

The δ functions can be replaced with (2.77) and E + = E + iδ by the imaginary part of the corresponding single-particle Green functions and give for the double sum

Solutions

385

expression under the integral i|ˆ v |f f |ˆ v |i ImGii (E + )ImGf f (E + + ¯ hω) ... = i,f

i,f

=

1 i|ˆ v G(E + + ¯ hω)|f f |ˆ v G(E + )|i , π 2 i,f

vImG(E + )). which for ω → 0 leads to Tr(ˆ vImG(E + )ˆ

9.5: Write β(g) in the form

β(g) =

dg L . dL g

For g → ∞ use g(L) = σLd−2 and σ independent of L to find g dg = σ(d − 2)Ld−3 = (d − 2) dL L which means lim β(g) = d − 2.

g→∞

For g → 0 use g(L) ∼ exp (−L/λ) with dg/dL = −g/λ and L lim β(g) = − . λ

g→0

The sign of β(g) is determined by dg/dL. For d ≤ 2 and assuming a monotonous function it is always negative while for d > 2 there is a sign change. Solutions for Chap. 10 10.1: For Bloch states ck, vk ′ write i i 1 ck|p|vk ′ = − ck|rH0 − H0 r|vk ′ = Ec (k) − Ev (k ′ ) ck|r|vk ′

m h ¯ h ¯

h. Use E(t) = −∂A/∂t which for direct transitions at k = 0 gives epcv = −iEg dcv /¯ and A(t) exp(iωt) with ¯ hω = Ec (k) − Ev (k) to obtain e i e Eg pcv · A = pcv · E = dcv · E ≃ dcv · E m ωm ¯ω h with ¯ hω ≃ Eg for near band gap excitations.

10.2: Express Bloch functions by Wannier functions (see Problem 6.2) to write 1 −ik·R ik ′ ·R ′ ′ ′ nk|p|n k = e e φ∗n (r − R)pφn′ (r − R′ )d3 r . N ′ R,R

Replace R′ = R − R′′ and change integration variable to r ′ = r − R to find 1 −i(k−k ′ )·R ik ′ ·R ′′ φ∗n (r ′ )pφn′ (r ′ − R′′ )d3 r ′ . e e nk|p|n′ k′ = N R R”

The first sum gives N δk,k ′ .

386

Solutions

10.3: The solution of the Schroedinger equation with a spherically symmetric potential is φnlm (r) = Rnl (r)Ylm (θ, ϕ) . The radial solution for the Coulomb problem is given by Rnl (r) = Nn ρl e(−ρ/2)F (l + 1 − η; 2l + 2; ρ) with the hypergeometric function depending on the variables η (for the energy) and ρ (for the radial coordinate). For bound states one has En = −Rexc /n2 and ρ = 2r/naB . Only s-states with l = 0 have a nonvanishing amplitude at r = 0, thus φn (0) = Nn F (1 − n; 2; 0) with Nn = 1/naB and |φn (0)|2 = (πa3B n3 )−1 . For the continuum (E = ¯ hω − Eg > 0), η = iγ, γ = (Rexc /(¯ hω − Eg ))1/2 and N = Γ(1 − iγ) exp(πγ/2) one obtains |φ(0)|2 = πγ

exp(πγ) . sinh(πγ)

The absorption coeﬃcient is the one of (10.23) for the uncorrelated electron-hole pairs multiplied by the enhancement factor C(ω). 10.4: Using (10.37) and (10.38) one can write † |Ψ0 = |ΨνQ = ΦνQ (k)c†ck e cvk h |Ψ0 . BνQ ckh ,cke ke −kh =Q

Thus, in application to the ground state |Ψ0 , the exciton operator is a linear combination of products c†ck e cvk h . Let us assume a two-band model, i.e. the excitons are formed from states with fixed c, v, and consider for simplicity the exciton with † , Bν ′ ,0 ] is determined Q = 0 which implies k e = kh = k. Then, the commutator [Bν,0 by the commutators [c†vk cck , c†ck ′ cvk ′ ] = c†vk cck c†ck ′ cvk ′ − c†ck ′ cvk ′ c†vk cck

= c†vk cvk ′ δk,k ′ − c†ck ′ cck δk,k ′ = δk,k ′ c†vk cvk − c†ck cck = (1 − ne (k) − nh (k)) δk,k ′ .

Without the electron and hole occupation this relation would lead to [Bν,0 , Bν†′ ,0 ] = δν,ν ′ , classifying excitons as bosons. This relation is valid, however, only if we consider a single exciton. The electron and hole occupations remind of the fact, that excitons are composed of fermions. 10.5: The zero of (10.42) close to the exciton resonance at ων0 determines the frequency of the corresponding longitudinal exciton: ε1 (ωL ) ≃ 1 +

4e2 fν0 2 2 − ωL = 0 , ε0 m ων0

which for ωL ≃ ων0 and ∆LT = ¯ h(ωL − ων0 ) allows to express the LT-splitting in terms of the exciton oscillator strength: ∆LT =

h2 fν0 2e2 ¯ ε0 m Eν0

with

Eν0 = ¯ hων0 .

Solutions

387

On the other hand the first term of the exchange interaction Vexch in (10.62) gives a contribution only for longitudinal excitons (with Q P cv ) which can be calculated as a perturbation correction to the transverse exciton energy Eν0 with the exciton 3 envelope function φν0 (r) after taking the Fourier transform of Vexch (note: 1/Vc exp(ik · r)d3 k = δ(r)) and yields (for singlet excitons with S = 0) ∆LT =

¯ 2 |Pcv |2 |φν0 (0)|2 2e2 h . 2 ε0 m2 Eν0

With Eν0 ≃ Eg one finds by comparison between the two obtained expressions for ∆LT 1 |Pcv |2 |φν0 (0)|2 fν0 = m Eν0 in accordance with (10.43). 10.6: Use the expansion of A(r, t) with aλκ (t) = aλκ eiω(κ)t to write the two contributions of Hrad as !2 1/2 h ¯ ∂ eλκ · eλ′ κ′ ω(κ)ω(κ′ ) = A(r, t) ∂t 2ε0 V λ,κ λ′ ,κ′ ′ × aλκ (t) − a†λ−κ (t) aλ′ κ ′ (t) − a†λ′ −κ ′ (t) ei(κ+κ )·r and

2 ¯ h c2 ∇ × A(r, t) = − (κ × eλκ ) · (κ′ × eλ′ κ′ ) 2ε0 V ′ ′ λ,κ λ ,κ ′ × aλκ (t) + a†λ−κ (t) aλ′ κ ′ (t) + a†λ′ −κ ′ (t) ei(κ+κ )·r

and perform the volume integral with ′ 1 ei(κ+κ )·r d3 r = δκ,−κ ′ . V

With eλκ · eλ′ κ = δλ,λ′ and (κ × eλκ ) · (−κ × eλ′ κ )) = −κ2 δλ,λ′ for transverse unit vectors this reduces with ω(κ) = ω(−κ) = cκ to Hrad =

1 1 hω(κ) a†λκ aλκ + ¯ hω(κ) aλκ a†λκ + a†λκ aλκ = ¯ . 2 2 λ,κ

λ,κ

10.7: In quasi-equilibrium one has to find only the solution for the oﬀ-diagonal elements of the density matrix, because the diagonal elements are the equilibrium distribution functions (Fermi–Dirac distribution). Therefore, one finds from i d ρ¯cv (k, t) = dcv E(t)ei(ǫck −ǫvk )t (fvk − fck ) dt h ¯ with E(t) = and integration over t ρ¯cv (k, t) =

1 2π¯ h

dω

1 2π

E(ω)e−iωtdω

dcv E(ω)ei(ǫck −ǫvk −ω)t (fvk − fck ) . ǫck − ǫvk − ω − iγ

388

Solutions

The dielectric polarization is obtained by ¯ P (t) = Tr ρ¯(t)d(t) 1 ρ¯cv (t)d¯vc (t) + ρ¯vc (t)d¯cv (t) = V k |dcv |2 (fvk − fck ) 1 dω = E(ω)e−iωt + c.c. 2π¯ hV k ǫck − ǫvk − ω − iγ and the dielectric susceptibility with χ(ω) = P (ω)/E(ω) is 1 1 |dcv |2 (fvk − fck ) χ(ω) = − V k h(ǫvk − ǫck + ω + iγ) ¯ 1 . − h(ǫck − ǫvk + ω + iγ) ¯ Finally one recovers the result from Sect. 10.2 with ε2 (ω) = Imχ(ω)/ε0 . 10.8: We demonstrate here the calculation only for one term, the others follow in † similar way. The example is [α†k αk , k,k α†k ′ −q αk ′ αk¯ ]. The commutator ¯ ′ ;q=0 vq αk+q ¯ is evaluated by using the fermion commutation rules when changing the order of the operators: [α†k αk , α†k+q α†k ′ −q αk ′ αk¯ ] = α†k αk α†k+q α†k ′ −q αk ′ αk¯ − ↔ ¯ ¯ α†k α†k ′ −q αk ′ αk¯ −α†k α†k+q αk α†k ′ −q αk ′ αk¯ − ↔ = δk,k+q ¯ ¯ (1)

= (1) + δk,k+q α†k α†k ′ −q αk αk ′ αk¯ − ↔ α†k+q α†k αk ′ αk¯ −α†k+q ¯ ¯ ¯ (2)

= (1) + (2) + δk,k ′ α†k+q α†k ′ −q αk¯ αk −α†k+q α†k ′ −q αk ′ α†k αk¯ αk − ↔ ¯ ¯ (3)

α†k ′ −q αk ′ αk¯ α†k αk − ↔ . α†k ′ −q αk ′ αk +α†k+q = (1) + (2) + (3) − δk,k¯ α†k+q ¯ ¯ (4)

In the last line, the last two terms cancel each other and one obtains the four contributions (1)–(4) which under the sum combine to the final result vq α†k+q α†k ′ −q αk ′ αk¯ ] = [α†k αk , ¯ ¯ ′ ;q=0 k,k

=

k ′ ;q=0

vq α†k α†k ′ −q αk−q αk ′ − α†k+q α†k ′ −q αk αk ′ .

The thermal expectation values of these terms are found in (10.115).

Index

absorption, 295 absorption coeﬃcient, 58 acoustic phonon, 48 adiabatic approximation, 21, 37, 231 potential, 37, 44, 232 alkali halide, 7 alloy, 273 Anderson model, 219 anti-ferromagnet, 176, 179 approximation adiabatic, 21, 231 Born, 278 Born, self-consistent, 278 Born–Oppenheimer, 21, 39, 231 coherent potential, 279 continuum, 52 Cooperon, 285 dipole, 295 eﬀective-mass, 147, 153, 266, 268, 299 harmonic, 39 Hartree, 98, 120 Hartree–Fock, 97, 98, 167, 181, 185, 322 local density, 123, 125 local spin density, 163 mean field, 100, 181, 258 molecular field, 179 one-band, 267 random phase, 106, 189, 282 relaxation time, 243 self-consistent field, 106 t-matrix, 278

Thomas–Fermi, 131 tight-binding, 142, 165, 202 virtual crystal, 277 atomic limit, 204 back scattering, 285 band conduction, 139 heavy hole, 152 index, 120, 128 light hole, 152 structure, 120, 127, 128, 136 valence, 139, 146 band gap, 134, 146 engineering, 155 problem, 126 band oﬀset, 153 band structure free electron, 136 of Al, 137 of Cu, 145 of Fe, 162 of Si, 139 of transition metals, 144 photonic, 8 BCS theory, 251 Bethe–Salpeter equation, 284, 303 binding chemical, 4, 61, 75 covalent, 16, 62, 65, 145, 147 heteropolar, 16, 61 homopolar, 16 ionic, 16

390

Index

metallic, 16 van der Waals, 16 Bloch electron, 127, 129, 185 equation, 171 function, 127, 129, 140, 267, 281, 326 representation, 172, 174, 177 state, 265 theorem, 6, 41, 326 vector, 318 Bloch equations optical, 318, 323 semiconductor, 320, 323 Bogoliubov transformation, 177, 215 Boltzmann equation, 243 bond charge, 62, 147 model, 62 Born approximation, 278 series, 271, 274 Born–Oppenheimer approximation, 21, 39, 231 Bose–Einstein condensation, 262 distribution, 47 Bragg condition, 134 Brillouin function, 182 zone, 6, 41, 43, 49, 120 bulk modulus, 71 carbon nanotubes, 146, 214 center coordinate, 86 central-cell correction, 268, 269 potential, 269 character table, 325 chemical potential, 81, 82, 89, 200 cluster, 4, 17 collective excitation, 42 mode, 42, 44 collision term, 243 compliance constant, 53 composite fermion, 228 compressibility, 54 conditions periodic boundary, 41, 77 conductance, 288

conductivity dc, 283, 287 electric, 25, 32, 232, 244, 280 heat, 32 configuration average, 275 equilibrium, 18, 64 of ions, 38 constant compliance, 53 elastic, 52 Madelung, 19 stiﬀness, 52 Cooper pairs, 255 cooperon approximation, 285 correlation, 12 density–density, 9, 32, 69 eﬀects, 195 electron, 195 electronic, 83 energy, 110 hole, 113 spin–spin, 30 Coulomb enhancement, 302 coupling deformation potential, 235 exciton–photon, 177 Fr¨ ohlich, 238 phonon–photon, 177 piezoelectric, 240 plasmon–phonon, 177 spin–orbit, 18 critical exponent, 183 point, 48, 297 temperature, 161, 179, 250, 261 crystal liquid, 4 mixed, 7, 273 molecular, 7 momentum, 5 photonic, 8 quasi, 4 structure, 6 crystal field splitting, 141 crystalline structure, 4 Curie temperature, 184 Curie–Weiss law, 184

Index current density, 282 diamagnetic, 282 electric, 25 heat, 25 paramagnetic, 282 cyclotron frequency, 84 de Haas–van Alphen oscillations, 136, 137 Debye frequency, 49, 255, 260 law, 50 model, 49, 175 temperature, 49, 261 Debye–Waller factor, 50, 69, 74 defect anti-site, 266, 272 Frenkel, 266 point, 266 deformation potential, 236 degeneracy Kramers, 128 spin, 129 density average, 11 fluctuation, 11, 211, 215 matrix, 317 parameter, 79, 99, 111 density functional theory, 64, 120, 123 density of states, 154, 187, 260, 271, 286 combined, 297 electron, 81 phonon, 47 DFT-LDA, 126, 163 diagram exchange, 304 ladder, 284 maximally crossing, 285 diamagnetism Landau–Peierls, 91 diamond, 50 dielectric function, 25, 32, 296 polarization, 25 screening, 31 dielectric constant high frequency, 58 static, 58

dielectric function, 57 electronic part, 102 longitudinal, 103 transverse, 103 dimensionality, 8, 153, 213, 214, 289 dipole approximation, 56, 295 moment, 55 operator, 56, 294 direct exchange, 169 term, 98 dislocation, 273 disorder compositional, 8, 67, 265, 273 diagonal, 273 isotope, 273 oﬀ-diagonal, 273 structural, 5, 8, 265, 273, 287 substitutional, 287 distribution Bose–Einstein, 47 Fermi–Dirac, 82, 283 doping, 266 Dulong–Petit law, 49 dynamical matrix, 41 Dyson equation, 199, 274, 277 edge channel, 228 eﬀect de Haas–van Alphen, 91, 136 quantum Hall, 155 eﬀective g factor, 86 mass, 86, 91, 150 single-particle potential, 120, 154 Einstein relation, 289 elastic constant, 52, 54 modulus, 53 electric conductivity, 25, 232 electron closed shell, 15 core, 75, 131 gas, 76 valence, 15, 75, 131 electron system low-dimensional, 84 two-dimensional, 153, 195, 222

391

392

Index

electron–electron interaction, 75, 96, 110, 120, 157, 165, 195, 198, 290 electron–hole excitation, 189 electron–hole excitation, 299 electron–hole pair, 299 electron–phonon interaction, 231 elementary flux quantum, 86 energy band structure, 120 dissipation, 58, 242 functional, 124 gap, 134, 269 loss spectroscopy, 105 relaxation, 242 energy bands of Al, 137 of Cu, 145, 298 of ferromagnetic Fe, 162 of Ge, 146 of Si, 139 of transition metals, 144 energy gap, 139, 146 energy–loss function, 32, 107, 109, 282, 314 ensemble canonical, 24 grand-canonical, 24 mean value, 22 equilibrium configuration, 18 position, 18, 38 thermodynamic, 20 exchange density, 122 diagram, 304 direct, 169 energy, 163 hole, 102 Rudermann–Kittel, 170 term, 98 excitation collective, 42, 109, 170, 189 electron–hole, 105, 299 elementary, 44 particle–hole, 108 spin-flip, 187

exciton longitudinal, 306 polariton, 307 transverse, 306 Wannier–Mott, 300 Fermi circle, 135 contour, 136 energy, 79, 144, 147, 210 gas, ideal, 78 hole, 102, 113 integral, 82, 331 liquid, 195 sphere, 77, 78, 92, 97, 98, 105, 108, 110, 163 surface, 91, 136, 257 temperature, 79 velocity, 79 wavelength, 79 Fermi surface of Ag, 138 of Al, 138 Fermi–Dirac distribution, 82, 211 fermion annihilation operator, 94 composite, 155, 228 creation operator, 94 heavy, 83, 91, 217 vacuum, 94 ferromagnet, 179 ferromagnetic ground state, 167, 171 insulator, 167, 169 ferromagnetism, 187 strong, 188 weak, 188 field quantization, 45 filling factor, 87, 155, 156 fine structure constant, 239 fluctuation density, 103 number, 103 spin, 180 Fock representation, 95 force constant, 40, 52, 61 four-center integral, 202

Index Fr¨ ohlich coupling constant, 239 polaron, 247 fractional charge, 228 filling, 223 quantum Hall states, 222 free energy, 71, 89, 211 function density–density correlation, 9 dielectric, 30, 296 pair–distribution, 9, 10, 102 Wannier, 165, 271 gap direct, 148 indirect, 148 parameter, 260 Gr¨ uneisen parameter, 72 relation, 72 gradient approximation, 126 grand-canonical ensemble, 24 potential, 89 Green function, 28 exciton, 303 retarded, 33, 196 spectral representation, 34 two-particle, 284, 303 ground state, 4, 78, 97 energy, 79, 99 of the lattice, 38, 45 theorem, 112 group of the wave vector, 7, 328 point, 6, 40, 237, 325 space, 7 translation, 5, 325 harmonic approximation, 39, 44 oscillator, 39 Hartree approximation, 98 equation, 121 self-energy, 200

393

Hartree–Fock approximation, 97, 105, 162, 185, 322 equation, 122 self-energy, 200 heavy fermion, 83, 91, 217 Heisenberg Hamiltonian, 166, 168, 180 model, 169 heterostructure, 153, 214 hierarchy problem, 197 Hohenberg–Kohn theorem, 123 hole, 147 Holstein–Primakoﬀ transformation, 173 Hooke’s law, 52 hopping matrix element, 165, 219 Hubbard band, 207, 279 Hamiltonian, 203 model, 201 Hund’s rule, 171 hybridization, 144, 162, 219 impurity, 265, 286 concentration, 275 deep, 269 interstitial, 266 shallow, 268, 269 substitutional, 266 insulator, 139, 201, 287 anti-ferromagnetic, 176 ferromagnetic, 169 Mott–Hubbard, 207 interaction eﬀective electron–electron, 252 electron–electron, 75, 96, 110, 120, 157, 195, 198 electron–phonon, 231, 233 electron–phonon, deformation potential, 236 electron–phonon, Fr¨ ohlich, 239 electron–phonon, piezoelectric, 240 exchange, 161 fermion–boson, 234 light–matter, 291, 293, 317 van der Waals, 62 interface, 273 ionization energy, 298

394

Index

irreducible diagram, 277 representation, 144, 325 Ising model, 169 isotope, 67 disorder, 273 eﬀect, 251, 261 Jastrow-type wave function, 226 jellium model, 19, 76, 96, 119 k.p theory, 149 Kohn–Sham equation, 123, 125 Kramers degeneracy, 128 Kramers–Kronig relation, 58, 296 Kubo formula, 28 Kubo–Greenwood formula, 284 Lam´e constants, 54 Landau level, 85, 279, 286 quantization, 92 lattice body-centered cubic, 7 Bravais, 6 displacement, 18, 38 distortion, 269 dynamics, 37, 42 dynamics, nonlinear, 44 face-centered cubic, 7 point, 5 reciprocal, 6, 11 thermal expansion, 44 translation, 40 vector, 5 with basis, 6 Laughlin wave function, 226 LCAO method, 269 LDA + U, 208 lifetime, 231, 241 population, 320 quasi-particle, 279 single-particle, 285 light–matter interaction, 291, 293, 317 light-scattering, 312 Lindhard function, 106, 213 linear response, 24

liquid, 9 Bose, 209 classical, 208 Fermi, 209 Luttinger, 213 quantum, 208 localization, 222 length, 265 weak, 285 longitudinal phonon, 56 Lyddane–Sachs–Teller relation, 60, 308 Madelung constant, 19 magnetic length, 85, 224 order, 161 magnetism, 165 itinerant electron, 185 theory of, 168 magnetization, 181, 185 magnon, 161, 170, 173, 189, 234 magnon dispersion, 190 mass eﬀective, 150, 154 operator, 125 matrix element hopping, 165 transfer, 165, 273 matter condensed, 4 soft, 4 mean field, 181 mean free path, 285 mesoscopic physics, 155 regime, 275 system, 17, 247 metal, 16, 50, 61, 110, 201, 287 alkali, 83 noble, 6, 80, 83, 136, 142 normal, 6, 80, 136, 142 transition, 6, 83, 143, 169 metal–insulator transition, 207, 265, 286, 287, 289 mobility, 246 edge, 287

Index model Anderson, 219 bond charge, 62 Hubbard, 219 jellium, 76 rigid ion, 61 shell, 61 Stoner, 185 Tomonaga–Luttinger, 217, 228 modulus bulk, 54 rigidity, 54 Young, 54 molecular field, 181 nanophysics, 8 nonlinear lattice dynamics, 44 nonparabolicity, 150 normal coordinate, 42, 44, 56 occupation number, 46, 95 number operator, 46, 95 operator annihilation, 44 creation, 44 dipole, 294 field, 281 statistical, 23 orbital sp2 , 146 sp3 , 146 anti-bonding, 146 bonding, 146 directed, 146 hybrid, 146 localized, 166 order ferrimagnetic, 170 ferromagnetic, 170 parameter, 183, 261 oscillations de Haas–van Alphen, 91, 138 quantum, 136 Shubnikov–de Haas, 92, 136 oscillator strength, 296, 302 overlap integral, 141

pair Cooper, 255 electron–hole, 139, 148 pair–distribution function, 10, 12, 102, 113 paramagnetism Pauli spin, 90, 162 particle independent, 195 particle–hole continuum, 108, 190 excitation, 108 partition function canonical, 24 grand-canonical, 24, 46 Pauli blocking, 323 principle, 93, 98 periodic boundary condition, 128, 329 potential, 119, 130, 138 phase ferromagnetic, 183 liquid, 4 paramagnetic, 183 solid, 4 transition, 179, 183, 184, 209 phonon, 19, 37, 45 acoustic, 48 confined modes, 67 dispersion curves, 60 energy, 45 focusing, 54 hot, 242 lifetime, 58 longitudinal, 56 operator, 56 optical, 55 surface, 66 transverse, 56 vacuum, 45 virtual, 248 phonon dispersion of α-Fe, 62 of Al, 61 of Cu(100), 67 of GaAs, 64 of GaAs(110), 68 of KI, 62

395

396

Index

of Si, 63 of solid 4 He, 63, 64 phonons acoustic, 235 optical, 238 photo-electron spectroscopy, 298 photoemission, 298 photon echo, 319 photonic band structure, 8 crystal, 8 piezoelectric coupling, 240 eﬀect, 240 tensor, 240 plasma frequency, 109 plasmon, 109 plasmon dispersion, 190 plasmon–phonon modes, 110 point defect, 265 point group, 325 polariton, 307 polarization dielectric, 307 function, 106 polaron, 232, 248 energy, 250 Fr¨ ohlich, 248 mass, 250 polymer, 214 potential eﬀective, 120 adiabatic, 22, 52, 63, 232 chemical, 24, 46, 81, 82, 89, 210 eﬀective, 122 eﬀective single-particle, 18, 120 empty core, 130 exchange–correlation, 154 grand-canonical, 89 Hartree, 121, 154 self-consistent, 270 process normal, 235 Umklapp, 235 propagator, 196 pseudo-potential, 132

quantum liquid, 209 oscillations, 89 well, 153 wire, 214 quantum Hall eﬀect, 92, 155 eﬀect, integer, 156 plateau, 156 quantum limit electric, 154 magnetic, 89, 156 quasi-crystal, 4 quasi-hole, 201 quasi-particle, 101, 129, 201, 205, 209–211, 213, 217, 218, 232 correction, 126 dispersion, 220 distribution, 213 eﬀective mass, 212 energy, 126, 210 lifetime, 279 weight, 213 Rabi frequency, 318, 323 random phase approximation, 106 Rashba eﬀect, 154 reflection coeﬃcient, 59 relation dispersion, 33 Kramers–Kronig, 33, 58 relaxation electron–lattice, 231 relaxation time transport, 231, 281 representation Bloch, 174 occupation number, 44, 46, 93, 241, 257 Wannier, 165, 174 resonance cyclotron, 85 electron spin, 86 paramagnetic, 86 spin–flip, 86 response function, 28, 35, 52, 55 linear, 24, 281

Index nonlinear, 32 transverse, 282 Reststrahlen band, 59, 66 RPA, 106, 189 Rudermann–Kittel exchange, 170 scaling theory, 289 scanning tunneling microscopy, 8 scattering back, 245 backward, 281 cross section, 313 forward, 245, 281 inelastic, 60 neutron, 60 rate, 241, 243 screening, 103 Thomas–Fermi, 109, 131 SDFT-LDA, 163 self-averaging, 275 self-consistent harmonic approximation, 74 solution, 121, 126 self-energy, 199, 205, 206, 209, 235, 241, 276, 284 exchange–correlation, 125 Hartree, 200 Hartree–Fock, 200 single-particle, 199 semiconductor, 7, 16, 80, 110, 136, 139, 146–148, 266 shell model, 61 Shubnikov–de Haas oscillations, 136 Slater determinant, 93, 121 soft mode, 74 solid amorphous, 4, 9, 273 crystalline, 4, 9 Sommerfeld coeﬃcient, 82, 212, 217, 222 correction, 302 expansion, 82, 333 model, 78, 97 sound propagation, 48, 54 velocity, 49 sound propagation longitudinal, 54 transverse, 54

specific heat, 48, 72, 175, 210, 212 of electrons, 82 of phonons, 48 spectral function, 34 spin, 17 degeneracy, 18, 129 density, 163 dynamics, 154 paramagnetism, 162 polarization, 126, 162, 186, 191 susceptibility, 188, 189, 210, 222 spin polarized electrons, 163 spin waves, 161, 170, 173 in anti-ferromagnets, 175 in ferromagnets, 170 spin–orbit coupling, 18 spintronics, 161 splitting longitudinal–transverse, 66 state anti-bonding, 146 bonding, 146 extended, 6, 265, 286 localized, 8, 265, 286 states anti-bonding, 269 bonding, 269 extended, 287 localized, 287 stiﬀness constant, 52 Stoner condition, 187, 191 continuum, 188 model, 185, 187 strain tensor, 52, 53, 240 structure diamond, 7 rocksalt, 7 zinc blende, 7 structure factor, 9, 71 dynamic, 11, 69, 107, 111, 313 static, 9, 10, 111 subband, 152 super-exchange, 169 superconductivity, 232, 250 superconductor high-Tc , 7 superlattice, 67, 136 isotope, 68

397

398

Index

surface, 8, 273 magnon, 161 phonon, 66 susceptibility, 28 dielectric, 28, 55 magnetic, 29, 90 system mesoscopic, 17 t-J model, 379 t-matrix, 274, 281 temperature critical, 179 Curie, 184 N´eel, 179 term direct, 167 exchange, 167 theorem dissipation–fluctuation, 35, 213 ground state, 112 Hohenberg–Kohn, 123 thermal average, 9 energy, 48 expansion, 44, 71 expectation value, 22, 46, 196, 282 lattice expansion, 44 time decoherence, 320 dephasing, 320 phase relaxation, 320 time reversal, 128 transfer matrix element, 165, 202, 273 transformation Bogoliubov, 177, 215 Holstein–Primakoﬀ, 173 Schrieﬀer–Wolﬀ, 252 transition direct, 295, 302 metal–insulator, 207, 265, 287, 289 vertical, 295, 302

translation group, 325 operator, 127 transport ballistic, 247 linear, 243 nonlinear, 243 relaxation time, 231, 243, 285 transverse eﬀective charge, 60 phonon, 56 truncation, 197 vacuum fermion, 94 phonon, 45 van Hove singularity, 48, 297 velocity Fermi, 79 of sound, 49 sound, 49, 54 vertex operator, 284 Voigt notation, 53, 240 Wannier function, 165 representation, 165, 174 Wannier function, 271 Wannier–Mott exciton, 300 warping, 54, 150, 152 Weiss constant, 181 Wigner crystal, 114, 228 Wigner–Seitz cell, 5, 128 work function, 298 XY model, 169 Zeeman term, 84 zero point contribution, 178 motion, 45 zone scheme extended, 133 reduced, 133 repeated, 133